nLab geometry of physics -- supersymmetry



this entry is one section of “geometry of physics – supergeometry and superphysics” which is one chapter of “geometry of physics

previous section: geometry of physics – supergeometry

next section geometry of physics – fundamental super p-branes


In the broad sense of the word a super-symmetry is an action of a supergroup, just as an ordinary symmetry is an action of some group. In fundamental particle physics the term is used more specifically for supergroups that extend some spacetime symmetry group, and for their action on the field content of some field theory.

If, by default, spacetime is locally modeled on Minkowski spacetime (of some dimension) whose isometry group is called the Poincaré group in this dimension, then supersymmetry in the strict sense of the word is super-extension of the Poincaré group by a supergroup whose odd-graded component is a real spin representation (“Majorana representation”) of the spin group in the given dimension. The result is called the super Poincaré group for the given spacetime dimension and choice of real spin representation, the latter also being called the “number of supersymmetries”.

Just as – in the spirit of Klein geometry – one recovers Minkowski spacetime as the coset space of the Poincaré group by the Lorentz group, so the coset superspace of the super Poincaré group by the spin group-cover of the Lorentz group defines super Minkowski spacetime. A superspacetime is then a supermanifold locally modeled on super Minkowski spacetime. This is what is mostly called “superspace” in the physics literature. But other local model spacetimes may be used, such as anti-de Sitter spacetime, leading similarly to super anti-de Sitter spacetimes etc.

Just as ordinary Lie groups are usefully studied via their Lie algebras, so super Lie groups are conveniently studied via their super Lie algebras, such as the super Poincaré Lie algebra. This is hence a super Lie algebra extensions of the Poincaré Lie algebra, again with the odd-graded part identified with the given real spin representation. Such representations have the special property that they admit a SpinSpin-equivariant bilinear pairing of two spinors to a vector, i.e. to an element in the Minkowski spacetime, regarded as a translation operation. This is precisely the structure that gives the odd-odd graded component of the bracket in the super Lie algebra. It is in this sense that in supersymmetry two odd spinorial transformations pair to a spacetime translation. It is noteworthy that the same bilinear spinor pairing also underlies other algebraic phenomena, such as the inner workings of twistors or the positivity relations that enter the spinorial proof of the positive energy theorem.

The Fano place

Since real spin representations have a comparatively rigid classification, there are algebraic constraints on supersymmetry groups in various dimensions. By a remarkable algebraic coincidence, the real spin representations in spacetime dimensions 3,4,5,6,7,10, and 11 are given by simple linear algebra over the real normed division algebras: the real numbers, the complex numbers, the quaternions and the octonions (Kugo-Townsend 82, Sudbery 84, Baez-Huerta 09, Baez-Huerta 10). (These are controled by the Fano plane, shown on the right.) This indicates some deep relation between supersymmetry and fundamental structures in mathematics (stable homotopy theory) where these algebras, and their associated Hopf fibrations, play a pivotal role in the Hopf invariant one theorem and the Adams spectral sequence.

Conversely, it turns out (theorem below) that the super Minkowski spacetimes in these dimensions are characterized as being the iterated maximal invariant central extensions of the superpoint (Huerta-Schreiber 17). This shows that supersymmetry in the special sense of spacetime supersymmetry is mathematically singled out among all supergroups. Given that supergroups themselves are mathematically singled out by Deligne's theorem on tensor categories, this shows that spacetime supersymmetry is not an ad-hoc concept and is of intrinsic interest independently of debated speculations on its realization at the (comparatively “low”) electroweak energy scale in the observable universe.

(In fact superconformal symmetry has an even more rigid classification: it exists only in dimensions 3,4,5, and 6, where it turns out to form the local super-symmetry groups appearing in the AdS-CFT correspondence.)






We start by considering the general concept of super-symmetry extensions of given ordinary symmetries:

There we find that super-extensions of spacetime symmetry are induced by real spin representations. There are several ways to get hold of real spin representations, and each of these gives one version of spacetime supersymmetry.

One way is to consider complex representations, which are easy to come by, and then try to carve out real sub-representations inside them by finding a real structure on the representation. In physics this is called the Majorana spinor construction. This we discuss in

Another way to get real spin representations is to invoke some algebraic magic that allows to construct them right away. This turns out to work in spacetime dimensions 3,4,6 and 10 as well as 4,5,7 and 11 by considering 2×22 \times 2 matrices with coefficients in one of the four real normed division algebras, equivalently one of the four real alternative division algebras. This we discuss in

Using the properties of real spin representations thus established, it is then immediate to construct spacetime supersymmetry super Lie algebras and supergroups. This we consider in

Finally we discuss that instead of pre-describing bosonic spacetime symmetry and then asking for super-extensions of it, one may discover spacetime, spin geometry and supersymmetry all at once by a systematic mathematical process starting from just the superpoint:

Supersymmetry extensions

We start by saying what it means, in generality, to have a supersymmetric extension of an ordinary symmetry. Here we are concerned with symmetry groups that are Lie groups, and we start by considering only the infinitesimal approximation, hence their Lie algebras.

To discuss super-extensions of Lie algebras, recall from geometry of physics – supergeometry the concept of super Lie algebras:


A super Lie algebra is a Lie algebra internal to the symmetric monoidal category sVect=(Vect /2, k,τ super)sVect = (Vect^{\mathbb{Z}/2}, \otimes_k, \tau^{super} ) of super vector spaces. Hence this is

  1. a super vector space 𝔤\mathfrak{g};

  2. a homomorphism

    [,]:𝔤 k𝔤𝔤 [-,-] \;\colon\; \mathfrak{g} \otimes_k \mathfrak{g} \longrightarrow \mathfrak{g}

    of super vector spaces (the super Lie bracket)

such that

  1. the bracket is skew-symmetric in that the following diagram commutes

    𝔤 k𝔤 τ 𝔤,𝔤 super 𝔤 k𝔤 [,] [,] 𝔤 1 𝔤 \array{ \mathfrak{g} \otimes_k \mathfrak{g} & \overset{\tau^{super}_{\mathfrak{g},\mathfrak{g}}}{\longrightarrow} & \mathfrak{g} \otimes_k \mathfrak{g} \\ {}^{\mathllap{[-,-]}}\downarrow && \downarrow^{\mathrlap{[-,-]}} \\ \mathfrak{g} &\underset{-1}{\longrightarrow}& \mathfrak{g} }

    (here τ super\tau^{super} is the braiding natural isomorphism in the category of super vector spaces)

  2. the Jacobi identity holds in that the following diagram commutes

    𝔤 k𝔤 k𝔤 τ 𝔤,𝔤 super kid 𝔤 k𝔤 k𝔤 [,[,]][[,],] [,[,]] 𝔤. \array{ \mathfrak{g} \otimes_k \mathfrak{g} \otimes_k \mathfrak{g} && \overset{\tau^{super}_{\mathfrak{g}, \mathfrak{g}} \otimes_k id }{\longrightarrow} && \mathfrak{g} \otimes_k \mathfrak{g} \otimes_k \mathfrak{g} \\ & {}_{\mathllap{[-,[-,-]]} - [[-,-],-] }\searrow && \swarrow_{\mathrlap{[-,[-,-]]}} \\ && \mathfrak{g} } \,.

Externally this means the following:


A super Lie algebra according to def. is equivalently

  1. a /2\mathbb{Z}/2-graded vector space 𝔤 even𝔤 odd\mathfrak{g}_{even} \oplus \mathfrak{g}_{odd};

  2. equipped with a bilinear map (the super Lie bracket)

    [,]:𝔤 k𝔤𝔤 [-,-] : \mathfrak{g}\otimes_k \mathfrak{g} \to \mathfrak{g}

    which is graded skew-symmetric: for x,y𝔤x,y \in \mathfrak{g} two elements of homogeneous degree σ x\sigma_x, σ y\sigma_y, respectively, then

    [x,y]=(1) σ xσ y[y,x], [x,y] = -(-1)^{\sigma_x \sigma_y} [y,x] \,,
  3. that satisfies the /2\mathbb{Z}/2-graded Jacobi identity in that for any three elements x,y,z𝔤x,y,z \in \mathfrak{g} of homogeneous super-degree σ x,σ y,σ z 2\sigma_x,\sigma_y,\sigma_z\in \mathbb{Z}_2 then

    [x,[y,z]]=[[x,y],z]+(1) σ xσ y[y,[x,z]]. [x, [y, z]] = [[x,y],z] + (-1)^{\sigma_x \cdot \sigma_y} [y, [x,z]] \,.

A homomorphism of super Lie algebras is a homomorphisms of the underlying super vector spaces which preserves the Lie bracket. We write

sLieAlg sLieAlg

for the resulting category of super Lie algebras.

Some obvious but important classes of examples are the following:


Every /2\mathbb{Z}/2-graded vector space VV becomes a super Lie algebra (def. , prop. ) by taking the super Lie bracket to be the zero map

[,]=0. [-,-] = 0 \,.

These may be called the “abelian” super Lie algebras.


Every ordinary Lie algebras becomes a super Lie algebra (def. , prop. ) concentrated in even degrees. This constitutes a fully faithful functor

LieAlgsLieAlg. LieAlg \hookrightarrow sLieAlg \,.

which is a coreflective subcategory inclusion in that it has a left adjoint

LieAlgAAAA()sLieAlg LieAlg \underoverset {\underset{ \overset{ \rightsquigarrow}{(-)} }{\longleftarrow}} {\hookrightarrow} {\phantom{AA}\bot\phantom{AA}} sLieAlg

given on the underlying super vector spaces by restriction to the even graded part

𝔰=𝔰 even. \overset{\rightsquigarrow}{\mathfrak{s}} = \mathfrak{s}_{even} \,.

Using this we may finally say what a super-extension is supposed to be:


Given an ordinary Lie algebra 𝔤\mathfrak{g}, then a super-extension of 𝔤\mathfrak{g} is super Lie algebra 𝔰\mathfrak{s} (def. , prop. ) equipped with a monomorphism of the form

i:𝔤𝔰 i \;\colon\; \mathfrak{g} \hookrightarrow \mathfrak{s}

(where 𝔤\mathfrak{g} is regarded as a super Lie algebra according to example )

such that this is an isomorphism on the even part (example )

i:𝔤𝔰 even. \overset{\rightsquigarrow}{i} \;\colon\; \mathfrak{g} \overset{\simeq}{\longrightarrow} \mathfrak{s}_{even} \,.

We now make explicit structure involved in super-extensions of Lie algebras:


Given an ordinary Lie algebra 𝔤\mathfrak{g}, then a choice of super-extension 𝔤𝔰\mathfrak{g} \hookrightarrow \mathfrak{s} according to def. is equivalently the following data:

  1. a vector space SS;

  2. a Lie action of 𝔤\mathfrak{g} on SS, hence a Lie algebra homomorphism

    ρ ():𝔤𝔤𝔩(S) \rho_{(-)} : \mathfrak{g} \longrightarrow \mathfrak{gl}(S)

    from 𝔤\mathfrak{g} to the endomorphism Lie algebra of SS;

  3. a symmetric bilinear map

    (,):S kS𝔤 (-,-) \;\colon\; S \otimes_k S \longrightarrow \mathfrak{g}

such that

  1. the pairing is 𝔤\mathfrak{g}-equivariant in that for all t𝔤t \in \mathfrak{g} then

    ρ t(,)=(ρ t(),())+(,ρ t()) \rho_{t}(-,-) = (\rho_t(-),(-)) + (-,\rho_t(-))
  2. the pairing satisfies

    ρ (ϕ,ϕ)(ϕ)=0 \rho_{(\phi,\phi)}(\phi) = 0

    for all ϕS\phi \in S.


By definition of super-extension, the underlying super vector space of 𝔰\mathfrak{s} is necessarily of the form

𝔰 even=𝔤𝔰 evenS𝔰 odd \mathfrak{s}_{even} = \underset{\mathfrak{s}_{even}}{\underbrace{\mathfrak{g}}} \oplus \underset{\mathfrak{s}_{odd}}{\underbrace{S}}

for some vector space SS.

Moreover the super Lie bracket on 𝔰\mathfrak{s} restricts to that of 𝔤\mathfrak{g} when restricted to 𝔤 k𝔤\mathfrak{g} \otimes_{k}\mathfrak{g} and otherwise constitutes

  1. a bilinear map

    ρ ()()[,]| 𝔰 even𝔰 odd:𝔤 kSS \rho_{(-)}(-) \coloneqq [-,-]\vert_{\mathfrak{s}_{even}\oplus \mathfrak{s}_{odd}} \;\colon\; \mathfrak{g} \otimes_k S \longrightarrow S
  2. a symmetric bilinear map

    (,)[,]| 𝔰 odd𝔰 odd:S kS𝔤. (-,-) \coloneqq [-,-]\vert_{\mathfrak{s}_{odd} \oplus \mathfrak{s}_{odd}} \;\colon\; S \otimes_k S \longrightarrow \mathfrak{g} \,.

This yields the claimed structure. The claimed properties of these linear maps are now just a restatement of the super-Jacobi identity in terms of this data:

  1. The restriction of the super Jacobi identity of 𝔰\mathfrak{s} to 𝔰 even k𝔰 even k𝔰 even\mathfrak{s}_{even} \otimes_k \mathfrak{s}_{even} \otimes_k \mathfrak{s}_{even} is equivalently the Jacobi identity on 𝔤\mathfrak{g} and hence is no new constraint.

  2. The restriction of the super Jacobi identity of 𝔰\mathfrak{s} to 𝔰 even k𝔰 even k𝔰 odd\mathfrak{s}_{even} \otimes_k \mathfrak{s}_{even} \otimes_k \mathfrak{s}_{odd} says that for t 1,t 2𝔤t_1,t_2 \in \mathfrak{g} and ϕS\phi \in S then

    ρ t 1(ρ t 2(ϕ))=ρ [t 1,t 2](ϕ)+ρ t 1(ρ t 2(ϕ)). \rho_{t_1}( \rho_{t_2}(\phi) ) = \rho_{[t_1,t_2]}(\phi) + \rho_{t_1}( \rho_{t_2}(\phi) ) \,.

    This is equivalent to

    ρ t 1ρ t 2ρ t 2ρ t 1=ρ [t 1,t 2] \rho_{t_1} \circ \rho_{t_2} - \rho_{t_2} \circ \rho_{t_1} = \rho_{[t_1,t_2]}

    which means equivalently that ρ ()\rho_{(-)} is a Lie algebra homomorphism from 𝔤\mathfrak{g} to the endomorphism Lie algebra of SS, hence that it is a Lie algebra representation of 𝔤\mathfrak{g} on SS.

  3. The restriction of the super Jacobi identity of 𝔰\mathfrak{s} to 𝔰 even k𝔰 odd k𝔰 odd\mathfrak{s}_{even} \otimes_k \mathfrak{s}_{odd} \otimes_k \mathfrak{s}_{odd} says that for t𝔤t \in \mathfrak{g} and ϕ,ψS\phi,\psi \in S then

    ρ t(ϕ,ψ)=(ρ t(ϕ),ψ)+(ϕ,ρ t(ψ)). \rho_{t}(\phi,\psi) = (\rho_t(\phi), \psi) + (\phi, \rho_t(\psi)) \,.

    This is exactly the claimed 𝔤\mathfrak{g}-equivariance of the pairing.

  4. The restriction of the super Jacobi identity of 𝔰\mathfrak{s} to 𝔰 odd k𝔰 odd k𝔰 odd\mathfrak{s}_{odd} \otimes_k \mathfrak{s}_{odd} \otimes_k \mathfrak{s}_{odd} implies that for all ψS\psi \in S that

    [ψ,(ψ,ψ)]=[(ψ,ψ),ψ][ψ,(ψ,ψ)] [\psi,(\psi,\psi)] = [ (\psi,\psi), \psi ] - [\psi, (\psi, \psi)]

    and hence in particular that

    [(ψ,ψ),ψ]=ρ (ψ,ψ)(ψ)=0. [(\psi,\psi),\psi] = \rho_{(\psi,\psi)}(\psi) = 0 \,.

    Therefore it only remains to show that this special case is in fact equivalent to the full odd-odd-odd super Jacobi identity. This follows by polarization: First insert ψ=ϕ 1+ϕ 2\psi = \phi_1 + \phi_2 into the above cubic condition to obtain a quadratic condition, then polarize once more in ϕ 2\phi_2.


(trivial super extension)

Given an ordinary Lie algebra 𝔤\mathfrak{g}, then for every choice of vector space VV there is the trivial super extension (def. ) of 𝔤\mathfrak{g}, with underlying vector space

𝔰𝔤S \mathfrak{s} \coloneqq \mathfrak{g} \oplus S

and with both the action and the pairing (via prop. ) trivial:

ρ=0 \rho = 0


(,)=0. (-,-) = 0 \,.

The key example of interest now is going to be this:


(super Poincaré super Lie algebra)

For dd \in \mathbb{N}, a super extension (def. ) of the Poincaré Lie algebra 𝔦𝔰𝔬( d1,1)\mathfrak{iso}(\mathbb{R}^{d-1,1}) (recalled as def. below) which is non-trivial (def. ) is obtained from the following data:

  1. a Lie algebra representation ρ\rho of 𝔰𝔬(d1,1)\mathfrak{so}(d-1,1) on some real vector space SS;

  2. an 𝔰𝔬\mathfrak{so}-equivariant symmetric \mathbb{R}-bilinear pairiing (,):S kS d1,1(-,-) \colon S \otimes_k S \to \mathbb{R}^{d-1,1}

It turns out that data as in example is given for ρ\rho the Lie algebra version of a real spin representation of the spin group Spin(d1,1)Spin(d-1,1) (this is prop. below). These we introduce and discuss now in Real spin representations.

The super-extensions of the Poincaré Lie algebra induced by real spin representations are called super Poincaré Lie algebras (def. ) below. These are the standard supersymmetry algebras in the physics literature.

But beware that there are more (“exotic”) super-extensions of the Poincaré Lie algebra than the “standard supersymmetry” super Poincaré Lie algebra from example : (The following example uses facts which we establish further below, the reader may want to skip this now and come back to it later.)


(an exotic super-extension of the Poincaré Lie algebra)

Let d{3,4,6,10}d \in \{3,4,6,10\} and let SRep (Spin(d1,1))S \in Rep_{\mathbb{R}}(Spin(d-1,1)) be an irreducible real spin representation in that dimension. Let (,):SS d1,1(-,-) \colon S \otimes S \to \mathbb{R}^{d-1,1} be the symmetric spinor pairing as in example , but let the action on SS via the even-odd super-bracket not be the Spin-action of 𝔰𝔬\mathfrak{so}, but the Clifford algebra action

Γ(): d1,1End(S). \Gamma(-) \;\colon \; \mathbb{R}^{d-1,1} \longrightarrow End(S) \,.

of d1,1\mathbb{R}^{d-1,1}. Then the condition

Γ(ψ,ψ)(ψ)=0 \Gamma(\psi,\psi)(\psi) = 0

from prop. does hold: this turns out to be equivalent to the Green-Schwarz superstring cocycle condition in these dimensions, here in its incarnation as the “3-ψ\psi rule” of Schray 96, see Baez-Huerta 09, theorem 10.

Now Γ()\Gamma(-) thus defined is clearly not a Lie algebra action and hence fails one of the other conditions in prop. , but this is readily fixed: take SS +S S \coloneqq S_+ \oplus S_- to be the direct sum of two copies of the Majorana spinor representation and take Γ()\Gamma(-) to map as before, but from S +S_+ to S S_-, acting as zero on S S_-. This forces the commutator of endomorphisms in the image of Γ\Gamma to vanish, and hence makes Γ()\Gamma(-) a Lie algebra action of the abelian Lie algebra d1,1\mathbb{R}^{d-1,1}. Hence we get an “exotic” super-extension of the Poincaré Lie algebra.


By prop. the data in example is sufficient for producing super-extensions (in the sense of def. ) of Poincaré Lie algebras, namely the super Poincaré Lie algebras. It is however not necessary: example is a super-extension in the sense of def. of the Poincaré Lie algebra which is not a super Poincaré Lie algebra in the standard sense of example .

One may add further natural conditions on the super-extension in order to narrow down to the super Poincaré super Lie algebras:

  1. From the assumption alone that SS is a spin representation and using that the SpinSpin-equivariant pairing has to take irreducible representations to irreducible representations, one may in some dimensions already deduce that the pairing has to land in d𝔦𝔰𝔬( d1,1)\mathbb{R}^{d} \hookrightarrow \mathfrak{iso}(\mathbb{R}^{d-1,1}). For d=4d = 4 and SS the irreducible Majorana representation this is done in Varadarajan 04, section 3.2.

  2. One may appeal to the Haag-Łopuszański-Sohnius theorem. This does rule out exotic super-extensions, by imposing the additional condition that P aP aP_a P^a remains a Casimir operator after super-extension, and more conditions. These conditions are well motivated from the expected symmetry-behaviour of S-matrices in field theory.

Below in supersymmetry from the superpoint we discuss a more fundamental statement: The super Poincaré Lie algebras at least in certain dimensions are singled out from a different perspective: they are precisely the result of iterative maximal invariant central extensions of the superpoint.

Real spin representations

By example we want a real spin representation in order to construct a spacetime supersymmetry super Lie algebra. There are different ways to get hold of real spin representations.

One way is to first consider complex representations, which are easy to come by, and then try to carve out real sub-representations inside them by finding a real structure on the representation. In physics this is called the Majorana spinor construction. This we discuss in

Another way to get real spin representations is to invoke some algebraic magic that allows to construct them right away. This turns out to work in spacetime dimensions 3,4,6 and 10 as well as 4,5,7 and 11 by considering 2×22 \times 2 matrices with coefficients in one of the four real normed division algebras, equivalently one of the four real alternative division algebras. This we discuss in

Real spinors as Majorana spinors

We will discuss the following concept, the ingredients of which we explain in the following


For dd \in \mathbb{N}, write Spin(d1,1)Spin(d-1,1) for the spin group (def. ) double cover (prop. ) of the proper orthochronous Lorentz group (def. ), let

ρ:Spin(d1,1)GL (V) \rho \colon Spin(d-1,1) \longrightarrow GL_{\mathbb{C}}(V)

be a unitary linear representation of Spin(d1,1)Spin(d-1,1) on some complex vector space VV.

Then ρ\rho is called

  • a Majorana representation if it admits a real structure JJ (def. );

    an element ψV\psi \in V is then called a real spinor if J(ψ)=ψJ(\psi) = \psi.

  • a symplectic Majorana representation if it admits a quaternionic structure JJ (def. )

    In this case J˜(0 J J 0)\tilde J \coloneqq \left( \array{ 0 & J \\ -J & 0 }\right) is a real structure on VVV \oplus V and in this way also symplectic Majorana spinors are regarded as a real spin representation.

We discuss this now in components (i.e. in terms of choices of linear bases), using standard notation and conventions from the physics literature (e.g. Castellani-D’Auria-Fré), but taking care to exhibit the abstract concept of real representations.

Below we work out the following:


Let V= d1,1V = \mathbb{R}^{d-1,1} be Minkowski spacetime of some dimension dd.

The following table lists the irreducible real spin representations of Spin(V)Spin(V).

ddSpin(d1,1)Spin(d-1,1)minimal real spin representation SSdim Sdim_{\mathbb{R}} S\;\;VV in terms of S *S^\astsupergravity
1 2\mathbb{Z}_2SS real1V(S *) 2V \simeq (S^\ast)^{\otimes}^2
2 >0× 2\mathbb{R}^{\gt 0} \times \mathbb{Z}_2S +,S S^+, S^- real1V(S + *) 2(S *) 2V \simeq ({S^+}^\ast)^{\otimes^2} \oplus ({S^-}^\ast)^{\otimes 2}
3SL(2,)SL(2,\mathbb{R})SS real2VSym 2S *V \simeq Sym^2 S^\ast
4SL(2,)SL(2,\mathbb{C})S SSS_{\mathbb{C}} \simeq S' \oplus S''4V S *S *V_{\mathbb{C}} \simeq {S'}^\ast \oplus {S''}^\astd=4 N=1 supergravity
5Sp(1,1)Sp(1,1)S S 0 WS_{\mathbb{C}} \simeq S_0 \otimes_{\mathbb{C}} W8 2S 0 *V \wedge^2 S_0^\ast \simeq \mathbb{C} \oplus V_{\mathbb{C}}
6SL(2,H)S ±S 0 ± WS^\pm_{\mathbb{C}} \simeq S_0^\pm \otimes_{\mathbb{C}} W8V 2S 0 + *( 2S 0 *) *V_{\mathbb{C}} \simeq \wedge^2 {S_0^+}^\ast \simeq (\wedge^2 {S_0^-}^\ast)^\ast
7S S 0 WS_{\mathbb{C}} \simeq S_0 \otimes_{\mathbb{C}} W16 2S 0 *V 2V \wedge^2 S_0^\ast \simeq V_{\mathbb{C}} \oplus \wedge^2 V_{\mathbb{C}}
8S S S S_{\mathbb{C}} \simeq S^\prime \oplus S^{\prime\prime}16S *S *V 3V {S'}^\ast {S''}^\ast \simeq V_{\mathbb{C}} \oplus \wedge^3 V_{\mathbb{C}}
9SS real16Sym 2S *V 4VSym^2 S^\ast \simeq \mathbb{R} \oplus V \wedge^4 V
10S +,S S^+ , S^- real16Sym 2(S ±) *V ± 5VSym^2(S^\pm)^\ast \simeq V \oplus \wedge_\pm^5 Vtype II supergravity
11SS real32Sym 2S *V 2V 5VSym^2 S^\ast \simeq V \oplus \wedge^2 V \oplus \wedge^5 V11-dimensional supergravity

Here WW is the 2-dimensional complex vector space on which the quaternions naturally act.

(table taken from Freed 99, page 48)


We recall the basics of Minkowski spacetimes d1,1\mathbb{R}^{d-1,1}, their Clifford algebras and spin groups.


For dd \in \mathbb{N}, we write d1,1\mathbb{R}^{d-1,1} for the real vector space d\mathbb{R}^{d} equipped with the quadratic form η\eta of signature

η=diag(+1,,+1,1). \eta = diag(+1, \cdots, +1, - 1) \,.

We write the standard coordinates on d1,1\mathbb{R}^{d-1,1}

(x 0,x 1,x 2,,x d1) (x_0, x_1, x_2, \cdots, x_{d-1})

with x 0x_0 the coordinate along the timelike direction: for v=v ax a d1,1v = v^a x_a \in \mathbb{R}^{d-1,1} any vector, then

η(v,v)=(v 0) 2+i=1d1(v i) 2. \eta(v,v) = -(v^0)^2 + \underoverset{i = 1}{d-1}{\sum} (v^{i})^2 \,.

For dd \in \mathbb{N}, write

O(d1,1)GL( d) O(d-1,1) \hookrightarrow GL(\mathbb{R}^d)

for the subgroup of the general linear group on those linear maps AA which preserve this bilinear form on Minkowski spacetime (def ), in that

η(A(),A())=η(,). \eta(A(-),A(-)) = \eta(-,-) \,.

This is the Lorentz group in dimension dd.

The elements in the Lorentz group in the image of the special orthogonal group SO(d1)O(d1,1)SO(d-1) \hookrightarrow O(d-1,1) are rotations in space. The further elements in the special Lorentz group SO(d1,1)SO(d-1,1), which mathematically are “hyperbolic rotations” in a space-time plane, are called boosts in physics.

One distinguishes the following further subgroups of the Lorentz group O(d1,1)O(d-1,1):

  • the proper Lorentz group

    SO(d1,1)O(d1,1) SO(d-1,1) \hookrightarrow O(d-1,1)

    is the subgroup of elements which have determinant +1 (as elements SO(d1,1)GL(d)SO(d-1,1)\hookrightarrow GL(d) of the general linear group);

  • the proper orthochronous (or restricted) Lorentz group

    SO +(d1,1)SO(d1,1) SO^+(d-1,1) \hookrightarrow SO(d-1,1)

    is the further subgroup of elements AA which preserve the time orientation of vectors vv in that (v 0>0)((Av) 0>0)(v^0 \gt 0) \Rightarrow ((A v)^0 \gt 0).


As a smooth manifold, the Lorentz group O(d1,1)O(d-1,1) (def. ) has four connected components. The connected component of the identity is the proper orthochronous Lorentz group SO +(3,1)SO^+(3,1) (def. ). The other three components are

  1. SO +(d1,1)PSO^+(d-1,1)\cdot P

  2. SO +(d1,1)TSO^+(d-1,1)\cdot T

  3. SO +(d1,1)PTSO^+(d-1,1)\cdot P T,

where, as matrices,

Pdiag(1,1,1,,1) P \coloneqq diag(1,-1,-1, \cdots, -1)

is the operation of point reflection at the origin in space, where

Tdiag(1,1,1,,1) T \coloneqq diag(-1,1,1, \cdots, 1)

is the operation of reflection in time and hence where

PT=TP=diag(1,1,,1) P T = T P = diag(-1,-1, \cdots, -1)

is point reflection in spacetime.

The following concept of the Clifford algebra (def. ) of Minkowski spacetime encodes the structure of the inner product space d1,1\mathbb{R}^{d-1,1} in terms of algebraic operation (“geometric algebra”), such that the action of the Lorentz group becomes represented by a conjugation action (example below). In particular this means that every element of the proper orthochronous Lorentz group may be “split in half” to yield a double cover: the spin group (def. below).


For dd \in \mathbb{N}, we write

Cl( d1,1) Cl(\mathbb{R}^{d-1,1})

for the /2\mathbb{Z}/2-graded associative algebra over \mathbb{R} which is generated from dd generators {Γ 0,Γ 1,Γ 2,,Γ d1}\{\Gamma_0, \Gamma_1, \Gamma_2, \cdots, \Gamma_{d-1}\} in odd degree (“Clifford generators”), subject to the relation

Γ aΓ b+Γ bΓ a=2η ab \Gamma_{a} \Gamma_b + \Gamma_b \Gamma_a = - 2\eta_{a b}

where η\eta is the inner product of Minkowski spacetime as in def. .

These relations say equivalently that

Γ 0 2=+1 Γ i 2=1fori{1,,d1} Γ aΓ b=Γ bΓ aforab. \begin{aligned} & \Gamma_0^2 = +1 \\ & \Gamma_i^2 = -1 \;\; \text{for}\; i \in \{1,\cdots, d-1\} \\ & \Gamma_a \Gamma_b = - \Gamma_b \Gamma_a \;\;\; \text{for}\; a \neq b \end{aligned} \,.

We write

Γ a 1a p1p!permutationsσ(1) |σ|Γ a σ(1)Γ a σ(p) \Gamma_{a_1 \cdots a_p} \;\coloneqq\; \frac{1}{p!} \underset{{permutations \atop \sigma}}{\sum} (-1)^{\vert \sigma\vert } \Gamma_{a_{\sigma(1)}} \cdots \Gamma_{a_{\sigma(p)}}

for the antisymmetrized product of pp Clifford generators. In particular, if all the a ia_i are pairwise distinct, then this is simply the plain product of generators

Γ a 1a n=Γ a 1Γ a nifi,j(a ia j). \Gamma_{a_1 \cdots a_n} = \Gamma_{a_1} \cdots \Gamma_{a_n} \;\;\; \text{if} \; \underset{i,j}{\forall} (a_i \neq a_j) \,.

Finally, write

()¯:Cl( d1,1)Cl( d1,1) \overline{(-)} \;\colon\; Cl(\mathbb{R}^{d-1,1}) \longrightarrow Cl(\mathbb{R}^{d-1,1})

for the algebra anti-automorphism given by

Γ a¯Γ a \overline{\Gamma_a} \coloneqq \Gamma_a
Γ aΓ b¯Γ bΓ a. \overline{\Gamma_a \Gamma_b} \coloneqq \Gamma_b \Gamma_a \,.

By construction, the vector space of linear combinations of the generators in a Clifford algebra Cl( d1,1)Cl(\mathbb{R}^{d-1,1}) (def. ) is canonically identified with Minkowski spacetime d1,1\mathbb{R}^{d-1,1} (def. )

()^: d1,1Cl( d1,1) \widehat{(-)} \;\colon\; \mathbb{R}^{d-1,1} \hookrightarrow Cl(\mathbb{R}^{d-1,1})


x aΓ a, x_a \mapsto \Gamma_a \,,

hence via

v=v ax xv^=v aΓ a, v = v^a x_x \mapsto \hat v = v^a \Gamma_a \,,

such that the defining quadratic form on d1,1\mathbb{R}^{d-1,1} is identified with the anti-commutator in the Clifford algebra

η(v 1,v 2)=12(v^ 1v^ 2+v^ 2v^ 1), \eta(v_1,v_2) = -\tfrac{1}{2}( \hat v_1 \hat v_2 + \hat v_2 \hat v_1) \,,

where on the right we are, in turn, identifying \mathbb{R} with the linear span of the unit in Cl( d1,1)Cl(\mathbb{R}^{d-1,1}).

The key point of the Clifford algebra (def. ) is that it realizes spacetime reflections, rotations and boosts via conjugation actions:


For dd \in \mathbb{N} and d1,1\mathbb{R}^{d-1,1} the Minkowski spacetime of def. , let v d1,1v \in \mathbb{R}^{d-1,1} be any vector, regarded as an element v^Cl( d1,1)\hat v \in Cl(\mathbb{R}^{d-1,1}) via remark .


  1. the conjugation action v^Γ a 1v^Γ a\hat v \mapsto -\Gamma_a^{-1} \hat v \Gamma_a of a single Clifford generator Γ a\Gamma_a on v^\hat v sends vv to its

reflection at the hyperplane x a=0x_a = 0;

  1. the conjugation action

    v^exp(α2Γ ab)v^exp(α2Γ ab) \hat v \mapsto \exp(- \tfrac{\alpha}{2} \Gamma_{a b}) \hat v \exp(\tfrac{\alpha}{2} \Gamma_{a b})

    sends vv to the result of rotating it in the (a,b)(a,b)-plane through an angle α\alpha.


This is immediate by inspection:

For the first statement observe that conjugating the Clifford generator Γ b\Gamma_b with Γ a\Gamma_a yields Γ b\Gamma_b up to a sign, depending on whether a=ba = b or not:

Γ a 1Γ bΓ a={Γ b |ifa=b Γ b |otherwise. - \Gamma_a^{-1} \Gamma_b \Gamma_a = \left\{ \array{ -\Gamma_b & \vert \text{if}\, a = b \\ \Gamma_b & \vert \text{otherwise} } \right. \,.

Therefore for hatv=v bΓ bhat v = v^b \Gamma_b then Γ a 1v^Γ a\Gamma_a^{-1} \hat v \Gamma_a is the result of multiplying the aa-component of vv by 1-1.

For the second statement, observe that

12[Γ ab,Γ c]=Γ aη bcΓ bη ac. -\tfrac{1}{2}[\Gamma_{a b}, \Gamma_c] = \Gamma_a \eta_{b c} - \Gamma_b \eta_{a c} \,.

This is the canonical action of the Lorentzian special orthogonal Lie algebra 𝔰𝔬(d1,1)\mathfrak{so}(d-1,1). Hence

exp(α2Γ ab)v^exp(α2Γ ab)=exp(12[Γ ab,])(v^) \exp(-\tfrac{\alpha}{2} \Gamma_{ab}) \hat v \exp(\tfrac{\alpha}{2} \Gamma_{ab}) = \exp(\tfrac{1}{2}[\Gamma_{a b}, -])(\hat v)

is the rotation action as claimed.


Since the reflections, rotations and boosts in example are given by conjugation actions, there is a crucial ambiguity in the Clifford elements that induce them:

  1. the conjugation action by Γ a\Gamma_a coincides precisely with the conjugation action by Γ a-\Gamma_a;

  2. the conjugation action by exp(α4Γ ab)\exp(\tfrac{\alpha}{4} \Gamma_{a b}) coincides precisely with the conjugation action by exp(α2Γ ab)-\exp(\tfrac{\alpha}{2}\Gamma_{a b}).


For dd \in \mathbb{N}, the spin group Spin(d1,1)Spin(d-1,1) is the group of even graded elements of the Clifford algebra Cl( d1,1)Cl(\mathbb{R}^{d-1,1}) (def. ) which are unitary with respect to the conjugation operation ()¯\overline{(-)} from def. :

Spin(d1,1){ACl( d1,1) even|A¯A=1}. Spin(d-1,1) \;\coloneqq\; \left\{ A \in Cl(\mathbb{R}^{d-1,1})_{even} \;\vert\; \overline{A} A = 1 \right\} \,.

The function

Spin(d1,1)GL( d1,1) Spin(d-1,1) \longrightarrow GL(\mathbb{R}^{d-1,1})

from the spin group (def. ) to the general linear group in dd-dimensions given by sending ASpin(d1,1)Cl( d1,1)A \in Spin(d-1,1) \hookrightarrow Cl(\mathbb{R}^{d-1,1}) to the conjugation action

A¯()A \overline{A}(-) A

(via the identification of Minkowski spacetime as the subspace of the Clifford algebra containing the linear combinations of the generators, according to remark )


  1. a group homomorphism onto the proper orthochronous Lorentz group (def. ):

    Spin(d1,1)SO +(d1,1) Spin(d-1,1) \longrightarrow SO^+(d-1,1)
  2. exhibiting a /2\mathbb{Z}/2-central extension.


That the function is a group homomorphism into the general linear group, hence that it acts by linear transformations on the generators follows by using that it clearly lands in automorphisms of the Clifford algebra.

That the function lands in the Lorentz group O(d1,1)GL(d)O(d-1,1) \hookrightarrow GL(d) follows from remark :

η(A¯v 1A,A¯v 2A) =12((A¯v^ 1A)(A¯v^ 2A)+(A¯v^ 2A)(A¯v^ 1A)) =12(A¯(v^ 1v^ 2+v^ 2v^ 1)A) =A¯A12(v^ 1v^ 2+v^ 2v^ 1) =η(v 1,v 2). \begin{aligned} \eta(\overline{A}v_1A , \overline{A} v_2 A) &= \tfrac{1}{2} \left( \left(\overline{A} \hat v_1 A\right) \left(\overline{A}\hat v_2 A\right) + \left(\overline{A} \hat v_2 A\right) \left(\overline{A} \hat v_1 A\right) \right) \\ & = \tfrac{1}{2} \left( \overline{A}(\hat v_1 \hat v_2 + \hat v_2 \hat v_1) A \right) \\ & = \overline{A} A \tfrac{1}{2}\left( \hat v_1 \hat v_2 + \hat v_2 \hat v_1\right) \\ & = \eta(v_1, v_2) \end{aligned} \,.

That it moreover lands in the proper Lorentz group SO(d1,1)SO(d-1,1) follows from observing (example ) that every reflection is given by the conjugation action by a linear combination of generators, which are excluded from the group Spin(d1,1)Spin(d-1,1) (as that is defined to be in the even subalgebra).

To see that the homomorphism is surjective, use that all elements of SO(d1,1)SO(d-1,1) are products of rotations in hyperplanes. If a hyperplane is spanned by the bivector (ω ab)(\omega^{a b}), then such a rotation is given, via example by the conjugation action by

exp(α2ω abΓ ab) \exp(\tfrac{\alpha}{2} \omega^{a b}\Gamma_{a b})

for some α\alpha, hence is in the image.

That the kernel is /2\mathbb{Z}/2 is clear from the fact that the only even Clifford elements which commute with all vectors are the multiples aCl( d1,1)a \in \mathbb{R} \hookrightarrow Cl(\mathbb{R}^{d-1,1}) of the identity. For these a¯=a\overline{a} = a and hence the condition a¯a=1\overline{a} a = 1 is equivalent to a 2=1a^2 = 1. It is clear that these two elements {+1,1}\{+1,-1\} are in the center of Spin(d1,1)Spin(d-1,1). This kernel reflects the ambiguity from remark .

Real structure on Unitary representations

We are interested in spin representations on real vector spaces. It turns out to be useful to obtain these from unitary representations on complex vector spaces by equipping these with real structure. In any case this is the approach used in much of the (physics) literature (with the real structure usually not made explicit, but phrased in terms of (symplectic) Majorana conditions).

Hence for reference, we here recollect the basics of the concept of unitary representations equipped with real structure.

All vector spaces in the following are taken to be finite dimensional vector spaces.


Let VV be a complex vector space. A real structure or quaternionic structure on VV is a real-linear map

ϕ:VV \phi \;\colon\; V \longrightarrow V

such that

  1. ϕ\phi is conjugate linear, in that ϕ(λv)=λ¯ϕ(v)\phi(\lambda v) = \overline{\lambda} \phi(v) for all λ\lambda \in \mathbb{C}, vVv \in V;

  2. ϕ 2={+id for real structure id for quaternionic structure\phi^2 = \left\{ \array{ +id & \text{for real structure} \\ -id & \text{for quaternionic structure} } \right.


A real structure ϕ\phi, def. , on a complex vector space VV corresponds to a choice of complex linear isomorphism

V V + V \simeq \mathbb{C} \otimes_{\mathbb{R}} V_+

of VV with the complexification of a real vector space V +V_+, namely the eigenspace of ϕ\phi for eigenvalue +1, while V iV +V_- \coloneqq i V_+ is the eigenspace of eigenvalue -1.

A quaternionic structure, def. , on VV gives it the structure of a left module over the quaternions (def. ) extending the underlying structure of a module over the complex numbers. Namely let

  1. Ii():VVI \coloneqq i(-) \colon V \to V be the operation of multiplying with ii \in \mathbb{C}

  2. Jϕ:VVJ \coloneqq \phi \colon V \to V be the given endomorphisms,

  3. KIJK \coloneqq I \circ J their composite,

then the conjugate complex linearity of ϕ\phi implies that

JI=IJ J \circ I = - I \circ J

and hence with J 2=1J^2 = -1 and I 2=1I^2 = -1 this means that II, JJ and KK act like the imaginary quaternions.


Let GG be a Lie group, let VV be a complex vector space and let

ρ:GGL (V) \rho \;\colon\; G \longrightarrow GL_{\mathbb{C}}(V)

be a complex linear representation of GG on VV, hence a group homomorphism form GG to the general linear group of VV over \mathbb{C}.

Then a real structure or quaternionic structure on (V,ρ)(V,\rho) is a real or complex structure, respectively, ϕ\phi on VV (def. ) such that ϕ\phi is GG-invariant under ρ\rho, i.e. such that for all gGg \in G then

ϕρ(g)=ρ(g)ϕ. \phi \circ \rho(g) = \rho(g) \circ \phi \,.

We will be interested in complex finite dimensional vector spaces equipped with hermitian forms, i.e. finite-dimensional complex Hilbert spaces:


A hermitian form (or symmetric complex sesquilinear form) ,\langle -,-\rangle on a complex vector space VV is a real bilinear form

,:V×V \langle -,- \rangle \;\colon\; V \times V \longrightarrow \mathbb{C}

such that for all v 1,v 2Vv_1, v_2 \in V and λ\lambda \in \mathbb{C} then

  1. (sesquilinearity) v 1,λv 2=λv 1,v 2\langle v_1, \lambda v_2 \rangle = \lambda \langle v_1, v_2 \rangle ,

  2. (conjugate symmetry) v 1,v 2 *=v 2,v 1\langle v_1, v_2\rangle^\ast = \langle v_2, v_1\rangle .

  3. (non-degeneracy) if v 1,=0\langle v_1,-\rangle = 0 then v 1=0v_1 = 0.

A complex linear function f:VVf \colon V \to V is unitary with respect to this hermitian form if it preserves it, in that

f(),f()=,. \langle f(-), f(-)\rangle = \langle -,-\rangle \,.


U(V)GL (V) U(V) \hookrightarrow GL_{\mathbb{C}}(V)

for the subgroup of unitary operators inside the general linear group.

A complex linear representation ρ:GGL (V)\rho \colon G \longrightarrow GL_{\mathbb{C}}(V) of a Lie group on VV is called a unitary representation if it factors through this subgroup

ρ:GU(V)GL (V). \rho \;\colon\; G \longrightarrow U(V) \hookrightarrow GL_{\mathbb{C}}(V) \,.

The following proposition uses assumptions stronger than what we have in the application to Majorana spinors (compact Lie group, positive definite hermitian form) but it nevertheless helps to see the pattern.


Let VV be a complex finite dimensional vector space, ,\langle -,-\rangle some positive definite hermitian form on VV, def. , let GG be a compact Lie group, and ρ:GU(V)\rho \colon G \to U(V) a unitary representation of GG on VV. Then ρ\rho carries a real structure or quaternionc structure ϕ\phi on ρ\rho (def. ) precisely if it carries a symmetric or anti-symmetric, respectively, non-degenerate complex-bilinear map

(,):V V. (-,-) \;\colon\; V \otimes_{\mathbb{C}} V \longrightarrow \mathbb{C} \,.


Given a real/quaternionic structure ϕ\phi, then the corresponding symmetric/anti-symmetric complex bilinear form is

(,)ϕ(),. (-,-) \coloneqq \langle \phi(-), -\rangle \,.

Conversely, given (,)(-,-), first define ϕ˜\tilde \phi by

(,)=ϕ˜(),, (-,-) = \langle \tilde\phi(-),-\rangle \,,

and then ϕ1|ϕ|ϕ\phi \coloneqq \frac{1}{\vert \phi\vert} \phi is the corresponding real/quaternionic structure.

If ϕ˜=ϕ\tilde\phi = \phi then (,)(-,-) is called compatible with ,\langle-,- \rangle.

(e.g. Meinrenken 13, p. 81)

Dirac and Weyl representations

Hence the task is now first to understand representations of the spin group on complex vector spaces (such as to then equip these with real structure). The basic such are called the Dirac representations.

One advantage of this approach of constructing real representations inside complex representations is the following:


For d2d \in 2\mathbb{N} an even natural number, then the complexification Cl( d1,1) Cl(\mathbb{R}^{d-1,1}) \otimes_{\mathbb{R}} \mathbb{C} of the Clifford algebra Cl( d1,1)Cl(\mathbb{R}^{d-1,1}) (def. ) is a central simple algebra, and hence by the Artin-Wedderburn theorem is isomorphic simply to a matrix algebra over the complex numbers.

Clearly, this drastically simplifies certain considerations about Clifford algebra, for instance it helps with analyzing Fierz identities.

This abstract isomorphism

Cl( 2ν1,1) Mat 2 ν×2 ν() Cl(\mathbb{R}^{2\nu-1,1}) \otimes_{\mathbb{R}} \mathbb{C} \;\simeq\; Mat_{2^\nu \times 2^\nu}(\mathbb{C})

is realized by the construction of the Dirac representation, below in prop. .

In the following we use standard notation for operations on matrices with entries in the complex numbers (and of course these matrices may in particular be complex row/column vectors, which may in particular be single complex numbers):

We will be discussing three different pairing operations on complex column vectors ψ 1,ψ 2 ν\psi_1, \psi_2 \in \mathbb{C}^\nu:

  • ψ 1 ψ 2\psi_1^\dagger \psi_2 – the standard hermitian form on ν\mathbb{C}^\nu, this will play a purely auxiliary role;

  • ψ 1,ψ 2ψ¯ 1ψ 2ψ 1 Γ 0ψ 2\langle \psi_1,\psi_2\rangle \coloneqq \overline{\psi}_1 \psi_2 \coloneqq \psi_1^\dagger \Gamma_0 \psi_2 – the Dirac pairing, this is the hermitian form with respect to which the spin representation below is a unitary representation;

  • (ψ 1,ψ 2)ψ 1 TCψ 2(\psi_1,\psi_2) \coloneqq \psi_1^T C \psi_2 – the Majorana pairing (for CC the charge conjugation matrix, prop. below), this turns out to coincide with the Dirac pairing above if ψ 1\psi_1 is a Majorana spinor.

The following is a standard convention for the complex representation of the Clifford algebra for d1,1\mathbb{R}^{d-1,1} (Castellani-D’Auria-Fré, (II.7.1)):


(Dirac representation)


d{2ν,2ν+1}forν,d4. d \in \{ 2\nu, 2 \nu + 1 \} \;\;\;\; \text{for}\, \nu \in \mathbb{N}\,,\; d\geq 4 \,.

Then there is a choice of complex linear representation of the Clifford algebra Cl( d1,1)Cl(\mathbb{R}^{d-1,1}) (def. ) on the complex vector space

V (2 ν) V \coloneqq \mathbb{C}^{(2^{\nu})}

such that

  1. Γ 0\Gamma_{0} is hermitian: Γ 0 =Γ 0\Gamma_0^\dagger = \Gamma_0;

  2. Γ spatial\Gamma_{spatial} is anti-hermitian: (Γ spatial) =Γ spatial(\Gamma_{spatial})^\dagger = - \Gamma_{spatial}.

Moreover, the pairing

,() Γ 0():V×V \langle -,-\rangle \coloneqq (-)^\dagger \Gamma_0 (-) \;\colon\; V \times V \longrightarrow \mathbb{C}

is a hermitian form (def. ) with respect to which the resulting representation of the spin group (def. ) is unitary:

Γ 0 1exp(ω abΓ ab) Γ 0 =exp(ω abΓ ab) =exp(ω abΓ ab) 1. \begin{aligned} \Gamma_0^{-1} \exp(\omega^{a b} \Gamma_{a b})^{\dagger} \Gamma_0 & = \exp(-\omega^{a b} \Gamma_{a b }) \\ & = \exp(\omega^{a b} \Gamma_{a b})^{-1} \end{aligned} \,.

These representations are called the Dirac representations, their elements are called Dirac spinors.


In the case d=4d = 4 consider the Pauli matrices {σ a} a=0 3\{\sigma_{a}\}_{a = 0}^3, defined by

σ ax a(x 0+x 1 x 2+ix 3 x 2ix 3 x 0x 1). \sigma_a x^a \coloneqq \left( \array{ x^0 + x^1 & x^2 + i x^3 \\ x^2 - i x^3 & x^0 - x^1 } \right) \,.

Then a Clifford representation as claimed is given by setting

Γ 0(0 id id 0) \Gamma_0 \coloneqq \left( \array{ 0 & id \\ id & 0 } \right)
Γ a(0 σ a σ a 0). \Gamma_a \coloneqq \left( \array{ 0 & \sigma_a \\ -\sigma_a & 0 } \right) \,.

From d=4d = 4 we proceed to higher dimension by induction, applying the following two steps:

odd dimensions

Suppose a Clifford representation {γ a}\{\gamma_a\} as claimed has been constructed in even dimension d=2νd = 2 \nu.

Then a Clifford representation in dimension d=2ν+1d = 2 \nu + 1 is given by taking

Γ a{γ a |ad2 ϵγ 0γ 1γ d2 |a=d1 \Gamma_a \coloneqq \left\{ \array{ \gamma_a & \vert \; a \leq d - 2 \\ \epsilon \gamma_0 \gamma_1 \cdots \gamma_{d-2} & \vert\; a = d-1 } \right.


ϵ={1 |νodd i |νeven. \epsilon = \left\{ \array{ 1 & \vert \; \nu \, \text{odd} \\ i & \vert \; \nu \, \text{even} } \right. \,.

even dimensions

Suppose a Clifford representation {γ a}\{\gamma_a\} as claimed has been constructed in even dimension d=2νd = 2 \nu.

Then a corresponding representation in dimension d+2d+2 is given by setting

Γ a<d(0 γ a γ a 0),Γ d=(0 id id 0),Γ d+1=(iid 0 0 iid). \Gamma_{a \lt d} \coloneqq \left( \array{ 0 & \gamma_a \\ \gamma_a & 0 } \right) \;\;\,, \;\;\; \Gamma_{d} = \left( \array{ 0 & id \\ -id & 0 } \right) \;\;\,, \;\;\; \Gamma_{d+1} = \left( \array{ i \mathrm{id} & 0 \\ 0 & -i \mathrm{id} } \right) \,.

Finally regarding the statement that this gives a unitary representation:

That ,() Γ 0()\langle -,-\rangle \coloneqq (-)^\dagger \Gamma_0 (-) is a hermitian form follows since Γ 0\Gamma_0 obtained by the above construction is a hermitian matrix.

Let a,b{1,,d1}a,b \in \{1, \cdots, d-1\} be spacelike and distinct indices. Then by the above we have

Γ 0 1(Γ aΓ b) Γ 0 =Γ 0 1Γ 0(Γ b Γ a ) =(Γ b)(Γ a) =Γ bΓ a =Γ aΓ b \begin{aligned} \Gamma_0^{-1} (\Gamma_a \Gamma_b)^\dagger \Gamma_0 & = \Gamma_0^{-1} \Gamma_0 (\Gamma_b^\dagger \Gamma_a^\dagger) \\ & = (-\Gamma_b) (-\Gamma_a) \\ & = \Gamma_b \Gamma_a \\ & = - \Gamma_a \Gamma_b \end{aligned}


Γ 0 1(Γ 0Γ a) =Γ 0 1Γ 0Γ a Γ 0 =(Γ a)(Γ 0) =Γ aΓ 0 =Γ 0Γ a. \begin{aligned} \Gamma_0^{-1} (\Gamma_0 \Gamma_a)^\dagger & = - \Gamma_0^{-1} \Gamma_0 \Gamma_a^\dagger \Gamma_0^\dagger \\ & = - (- \Gamma_a) (\Gamma_0) \\ & = \Gamma_a \Gamma_0 \\ & = - \Gamma_0 \Gamma_a \end{aligned} \,.

This means that the exponent of exp(ω abΓ aΓ b)\exp(\omega^{a b} \Gamma_a \Gamma_b) is an anti-hermitian matrix, hence that exponential is a unitary operator.


(Weyl representation)

Since by prop. the Dirac representations in dimensions d=2νd = 2\nu and d+1=2ν+1d+1 = 2\nu+1 have the same underlying complex vector space, the element

Γ dΓ 0Γ 1Γ d1 \Gamma_{d} \propto \Gamma_0 \Gamma_1 \cdots \Gamma_{d-1}

acts Spin(d1,1)Spin(d-1,1)-invariantly on the representation space of the Dirac Spin(d1,1)Spin(d-1,1)-representation for even dd.

Moreover, since Γ 0Γ 1Γ d1\Gamma_0 \Gamma_1 \cdots \Gamma_{d-1} squares to ±1\pm 1, there is a choice of complex prefactor cc such that

Γ d+1cΓ 0Γ 1Γ d1 \Gamma_{d+1} \coloneqq c \Gamma_0 \Gamma_1 \cdots \Gamma_{d-1}

squares to +1. This is called the chirality operator.

(The notation Γ d+1\Gamma_{d+1} for this operator originates from times when only d=4d = 4 was considered. Clearly this notation has its pitfalls when various dd are considered, but nevertheless it is still commonly used this way, see e.g. Castellani-D’Auria-Fré, section (II.7.11) and top of p. 523).

Therefore this representation decomposes as a direct sum

V=V +V V = V_+ \oplus V_-

of the eigenspaces V ±V_{\pm} of the chirality operator, respectively. These V ±V_{\pm} are called the two Weyl representations of Spin(d1,1)Spin(d-1,1). An element of these is called a chiral spinor (“left handed”, “right handed”, respectively).


For a Clifford algebra representation on (2 ν)\mathbb{C}^{(2^\nu)} as in prop. , we write

()¯() Γ 0:Mat ν×1()Mat(1×ν)() \overline{(-)} \coloneqq (-)^\dagger \Gamma_0 \;\colon\; Mat_{\nu \times 1}(\mathbb{C}) \longrightarrow Mat(1 \times \nu)(\mathbb{C})

for the map from complex column vectors to complex row vectors which is hermitian congugation () =(() *) T(-)^\dagger = ((-)^\ast)^T followed by matrix multiplication with Γ 0\Gamma_0 from the right.

This operation is called Dirac conjugation.

In terms of this the hermitian form from prop. (Dirac pairing) reads

,=()¯(). \langle -,-\rangle = \overline{(-)}(-) \,.

The operator adjoint A¯\overline{A} of a 2 ν×2 ν2^\nu \times 2^\nu-matrix AA with respect to the Dirac pairing of def. , characterized by

A(),()=,A¯and,A=A¯, \langle A (-), (-)\rangle = \langle - , \overline{A} -\rangle \;\;\;\text{and} \;\;\; \langle -, A -\rangle = \langle \overline{A} - , -\rangle

is given by

A¯=Γ 0 1A Γ 0. \overline{A} = \Gamma_0^{-1} A^\dagger \Gamma_0 \,.

All the representations of the Clifford generators from prop. are Dirac self-conjugate in that

Γ¯ a=Γ a \overline{\Gamma}_a = \Gamma_a

saying that this Dirac representation respects the canonical antihomomorphism from def. .


For the first claim consider

Aψ 1,ψ 2 =ψ 1 A Γ 0ψ 2 =ψ 1 Γ 0(Γ 0 1A Γ 0)ψ 2 =ψ 1,(Γ 0 1AΓ 0)ψ 2. \begin{aligned} \langle A \psi_1, \psi_2\rangle & = \psi_1^\dagger A^\dagger \Gamma_0 \psi_2 \\ & = \psi_1^\dagger \Gamma_0 (\Gamma_0^{-1} A^\dagger \Gamma_0) \psi_2 \\ & = \langle \psi_1, (\Gamma_0^{-1} A \Gamma_0)\psi_2\rangle \end{aligned} \,.


ψ 1,Aψ 2 =ψ 1 Γ 0Aψ 2 =ψ 1 Γ 0AΓ 0 1Γ 0ψ 2 =((Γ 0 1) A (Γ 0) ψ 1) Γ 0ψ 2 =(Γ 0 1A Γ 0ψ 1) Γ 0ψ 2 =A¯ψ 1,ψ 2, \begin{aligned} \langle \psi_1, A \psi_2\rangle & = \psi_1^\dagger \Gamma_0 A \psi_2 \\ & = \psi_1^\dagger \Gamma_0 A \Gamma_0^{-1} \Gamma_0 \psi_2 \\ & = ( (\Gamma_0^{-1})^\dagger A^\dagger (\Gamma_0)^\dagger \psi_1 )^\dagger \Gamma_0 \psi_2 \\ & = ( \Gamma_0^{-1} A^\dagger \Gamma_0 \psi_1 )^\dagger \Gamma_0 \psi_2 \\ &= \langle \overline{A} \psi_1, \psi_2\rangle \end{aligned} \,,

where we used that Γ 0 1=Γ 0\Gamma_0^{-1} = \Gamma_0 (by def. ) and Γ 0 =Γ 0\Gamma_0^\dagger = \Gamma_0 (by prop. ).

Now for the second claim, use def. and prop. to find

Γ¯ 0 =Γ 0 1Γ 0 Γ 0 =Γ 0 1Γ 0Γ 0 =Γ 0 \begin{aligned} \overline{\Gamma}_0 & = \Gamma_0^{-1}\Gamma_0^\dagger \Gamma_0 \\ & = \Gamma_0^{-1} \Gamma_0 \Gamma_0 \\ & = \Gamma_0 \end{aligned}


Γ¯ spatial =Γ 0 1Γ spatial Γ 0 =Γ 0 1Γ spatialΓ 0 =+Γ 0 1Γ 0Γ spatial =Γ spatial. \begin{aligned} \overline{\Gamma}_{spatial} & = \Gamma_0^{-1} \Gamma_{spatial}^\dagger\Gamma_0 \\ &= - \Gamma_0^{-1} \Gamma_{spatial} \Gamma_0 \\ & = + \Gamma_0^{-1} \Gamma_0 \Gamma_{spatial} \\ &= \Gamma_{spatial} \end{aligned} \,.

Majorana spinors and Real structure

We now define Majorana spinors in the traditional way, and then demonstrate that these are real spin representations in the sense of def. .

The key technical ingredient for the definition is the following similarity transformations relating the Dirac Clifford representation to its transpose:


Given the Clifford algebra representation of the form of prop. , consider the equation

C (±)Γ a=±Γ a TC (±) C_{(\pm)} \Gamma_a = \pm \Gamma_a^T C_{(\pm)}

for C (±)Mat ν×n()C_{(\pm)} \in Mat_{\nu \times n}(\mathbb{C}).

In even dimensions d=2νd = 2 \nu then both these equations have a solution, wheras in odd dimensions d=2ν+1d = 2 \nu + 1 only one of them does (alternatingly, starting with C (+)C_{(+)} in dimension 5). Either C (±)C_{(\pm)} is called the charge conjugation matrix.

Moreover, all C (±)C_{(\pm)} may be chosen to be real matrices

(C (±)) *=C (±) (C_{(\pm)})^\ast = C_{(\pm)}

and in addition they satisfy the following relations:

4C (+) T=C (+)C_{(+)}^T = -C_{(+)}; C (+) 2=1C_{(+)}^2 = -1C () T=C (+)C_{(-)}^T = -C_{(+)}; C () 2=1C_{(-)}^2 = -1
5C (+) T=C (+)C_{(+)}^T = -C_{(+)}; C (+) 2=1C_{(+)}^2 = -1
6C (+) T=C (+)C_{(+)}^T = -C_{(+)}; C (+) 2=1C_{(+)}^2 = -1C () T=C ()C_{(-)}^T = C_{(-)}; C () 2=1C_{(-)}^2 = 1
7C () T=C ()C_{(-)}^T = C_{(-)}; C () 2=1C_{(-)}^2 = 1
8C (+) T=C (+)C_{(+)}^T = C_{(+)}; C (+) 2=1C_{(+)}^2 = 1C () T=C ()C_{(-)}^T = C_{(-)}; C () 2=1C_{(-)}^2 = 1
9C (+) T=C (+)C_{(+)}^T = C_{(+)}; C (+) 2=1C_{(+)}^2 = 1
10C (+) T=C (+)C_{(+)}^T = C_{(+)}; C (+) 2=1C_{(+)}^2 = 1C () T=C ()C_{(-)}^T = -C_{(-)}; C () 2=1C_{(-)}^2 = -1
11C () T=C ()C_{(-)}^T = -C_{(-)}; C () 2=1C_{(-)}^2 = -1

(This is for instance in Castellani-D’Auria-Fré, section (II.7.2), table (II.7.1), but beware that there C ()C_{(-)} in d=10,11d = 10, 11 is claimed to be symmetric, while instead it is anti-symmetric as shown above, see van Proeyen 99, table 1, Laenen, table E.3).


Prop. implies that for all C (±)C_{(\pm)} listed there, then

C 1=C T. C^{-1} = C^T \,.

This implies in all cases that

Γ aC (±) T=±C (±) TΓ a T. \Gamma_a C_{(\pm)}^T = \pm C_{(\pm)}^T \Gamma_a^T \,.

For d{4,8,9,10,11}d \in \{4,8,9,10,11\}, let V= νV = \mathbb{C}^\nu as above. Write {Γ a}\{\Gamma_a\} for a Dirac representation according to prop. , and write

C{C () ford=4 C (+) ford=8 C (+) ford=9 C (+)orC () ford=10 C () ford=11 C \coloneqq \left\{ \array{ C_{(-)} & \text{for}\; d = 4 \\ C_{(+)} & \text{for}\; d = 8 \\ C_{(+)} & \text{for}\; d = 9 \\ C_{(+)} or C_{(-)} & \text{for}\; d = 10 \\ C_{(-)} & \text{for}\; d = 11 } \right.

for the choice of charge conjugation matrix from prop. as shown. Then the function

J:VV J \colon V \longrightarrow V

given by

ψCΓ 0 Tψ * \psi \mapsto C \Gamma_0^T \psi^\ast

is a real structure (def. ) for the corresponding complex spin representation on ν\mathbb{C}^\nu.


The conjugate linearity of JJ is clear, since () *(-)^\ast is conjugate linear and matrix multiplication is complex linear.

To see that JJ squares to +1 in the given dimensions: Applying it twice yields,

J 2ψ =CΓ 0 T(CΓ 0 Tψ *) * =CΓ 0 TCΓ 0 ψ =CΓ 0 TC=±CΓ 0Γ 0ψ =±C (±) 2Γ 0 2ψ =±C (±) 2ψ, \begin{aligned} J^2 \psi &= C \Gamma_0^T (C \Gamma_0^T\psi^\ast)^\ast \\ & = C \Gamma_0^T C \Gamma_0^\dagger \psi \\ &= C \underset{= \pm C \Gamma_0}{\underbrace{\Gamma_0^T C}} \Gamma_0 \psi \\ & = \pm C_{(\pm)}^2 \Gamma_0^2 \psi \\ & = \pm C_{(\pm)}^2 \psi \end{aligned} \,,

where we used Γ 0 =Γ 0\Gamma_0^\dagger = \Gamma_0 from prop. , C *=*C^\ast = \ast from prop. and then the defining equation of the charge conjugation matrix Γ a TC (±)=±C (±)Γ a\Gamma_a^T C_{(\pm)} = \pm C_{(\pm)} \Gamma_a (def. ), finally the defining relation Γ 0 2=+1\Gamma_0^2 = +1.

Hence this holds whenever there exists a choice C (±)C_{(\pm)} for the charge conjugation matrix with C (±) 2=±1C_{(\pm)}^2 = \pm 1. Comparison with the table from prop. shows that this is the case in d=4,8,9,10,11d = 4,8,9,10,11.

Finally to see that JJ is spin-invariant (in Castellani-D’Auria-Fré this is essentially (II.2.29)), it is sufficient to show for distinct indices a,ba,b, that

J(Γ aΓ bψ)=Γ aΓ bJ(ψ). J(\Gamma_a \Gamma_b \psi) = \Gamma_a \Gamma_b J(\psi) \,.

First let a,ba,b both be spatial. Then

J(Γ aΓ bψ) =CΓ 0 TΓ a *Γ b *ψ * =CΓ 0 T(Γ a T)(Γ b T)ψ * =CΓ 0 TΓ a TΓ b Tψ * =CΓ a TΓ b TΓ 0 Tψ * =Γ aΓ bCΓ 0 Tψ * =Γ aΓ bJ(ψ). \begin{aligned} J(\Gamma_a \Gamma_b \psi) & = C \Gamma_0^T \Gamma_a^\ast \Gamma_b^\ast \psi^\ast \\ & = C \Gamma_0^T (-\Gamma_a^T)(-\Gamma_b^T) \psi^\ast \\ & = C \Gamma_0^T \Gamma_a^T \Gamma_b^T \psi^\ast \\ & = C \Gamma_a^T \Gamma_b^T \Gamma_0^T \psi^\ast \\ & = \Gamma_a \Gamma_b C \Gamma_0^T \psi^\ast \\ & = \Gamma_a \Gamma_b J(\psi) \end{aligned} \,.

Here we first used that Γ spatial =Γ spatial\Gamma_{spatial}^\dagger = -\Gamma_{spatial} (prop. ), hence that Γ spatial *=Γ spatial T\Gamma_{spatial}^\ast = - \Gamma_{spatial}^T and then that Γ 0\Gamma_0 anti-commutes with the spatial Clifford matrices, hence that Γ 0 T\Gamma_0^T anti-commutes the the transposeso fthe spatial Clifford matrices. Then we used the defining equation for the charge conjugation matrix, which says that passing it through a Gamma-matrix yields a transpose, up to a global sign. That global sign cancels since we pass through two Gamma matrices.

Finally, that the same conclusion holds for Γ spatialΓ spatial\Gamma_{spatial} \Gamma_{spatial} replaced by Γ 0Γ spatial\Gamma_0 \Gamma_{spatial}: The above reasoning applies with two extra signs picked up: one from the fact that Γ 0\Gamma_0 commutes with itself, one from the fact that it is hermitian, by prop. . These two signs cancel:

J(Γ 0Γ aψ) =CΓ 0 TΓ 0 *Γ a *ψ * =CΓ 0 T(+Γ 0 T)(Γ a T)ψ * =CΓ 0 TΓ 0 TΓ a Tψ * =+CΓ 0 TΓ a TΓ 0 Tψ * =Γ 0Γ aΓ 0 Tψ * =Γ 0Γ aJ(ψ). \begin{aligned} J(\Gamma_0 \Gamma_a \psi) & = C \Gamma_0^T \Gamma_0^\ast \Gamma_a^\ast \psi^\ast \\ & = C \Gamma_0^T (+\Gamma_0^T)(-\Gamma_a^T) \psi^\ast \\ & = - C \Gamma_0^T \Gamma_0^T \Gamma_a^T \psi^\ast \\ & = + C \Gamma_0^T \Gamma_a^T \Gamma_0^T \psi^\ast \\ & = \Gamma_0 \Gamma_a \Gamma_0^T \psi^\ast \\ &= \Gamma_0 \Gamma_a J(\psi) \end{aligned} \,.

Prop. implies that given a Dirac representation (prop. ) VV, then the real subspace SVS \hookrightarrow V of real elements, i.e. elements ψ\psi with Jψ=ψJ \psi = \psi according to prop. is a sub-representation. This is called the Majorana representation inside the Dirac representation (if it exists).


If C=C (±)C = C_{(\pm)} is the charge conjugation matrix according to prop. , then the real structure JJ from prop. commutes or anti-commutes with the action of single Clifford generators, according to the subscript of C=C (±)C = C_{(\pm)}:

J(Γ a())=±Γ aJ(). J(\Gamma_a(-)) = \pm \Gamma_a J(-) \,.

This is same kind of computation as in the proof prop. . First let aa be a spatial index, then we get

J(Γ aψ) =CΓ 0 TΓ a *ψ * =CΓ 0 T(Γ a T)ψ * =+CΓ a TΓ 0 Tψ * =ϵC TΓ a TΓ 0 T =±ϵΓ aC TΓ 0 Tψ * =±ϵ 2Γ aCΓ 0 Tψ * =±Γ aJ(ψ), \begin{aligned} J(\Gamma_a \psi) & = C \Gamma_0^T \Gamma_a^\ast \psi^\ast \\ & = C \Gamma_0^T (-\Gamma_a^T) \psi^\ast \\ & = + C \Gamma_a^T \Gamma_0^T \psi^\ast \\ & = \epsilon C^T \Gamma_a^T \Gamma_0^T \\ & = \pm \epsilon \Gamma_a C^T \Gamma_0^T \psi^\ast \\ & = \pm \epsilon^2 \Gamma_a C \Gamma_0^T \psi^\ast \\ & = \pm \Gamma_a J(\psi) \end{aligned} \,,

where, by comparison with the table in prop. , ϵ\epsilon is the sign in C T=ϵCC^T = \epsilon C, which cancels out, and the remaining ±\pm is the sign in C=C (±)C = C_{(\pm)}, due to remark .

For the timelike index we similarly get:

J(Γ 0ψ) =CΓ 0 TΓ 0 *ψ * =+CΓ 0 TΓ 0 Tψ * =ϵC TΓ 0 TΓ 0 T =±ϵΓ 0C TΓ 0 Tψ * =±Γ 0CΓ 0 Tψ * =±Γ 0J(ψ). \begin{aligned} J(\Gamma_0 \psi) & = C \Gamma_0^T \Gamma_0^\ast \psi^\ast \\ & = + C \Gamma_0^T \Gamma_0^T \psi^\ast \\ & = \epsilon C^T \Gamma_0^T \Gamma_0^T \\ & = \pm \epsilon \Gamma_0 C^T \Gamma_0^T \psi^\ast \\ & = \pm \Gamma_0 C \Gamma_0^T \psi^\ast \\ & = \pm \Gamma_0 J(\psi) \end{aligned} \,.

We record some immediate consequences:


The complex bilinear form

(,)J(),() (-,-) \coloneqq \langle J(-),(-)\rangle

induced via the real structure JJ of prop. from the hermitian form ,\langle -,-\rangle of prop. is that represented by the charge conjugation matrix of prop.

(,)=() TC(). (-,-) = (-)^T C (-) \,.

By direct unwinding of the various definitions and results from above:

J(ψ 1),ψ 2 =CΓ 0 Tψ 1 *,ψ 2 =(CΓ 0 Tψ 1 *) Γ 0ψ 2 =ψ 1 TC Γ 0 *Γ 0ψ 2 =ψ 1 TCψ 2. \begin{aligned} \langle J(\psi_1),\psi_2 \rangle &= \langle C \Gamma_0^T\psi_1^\ast, \psi_2\rangle \\ & = (C \Gamma_0^T \psi_1^\ast)^\dagger \Gamma_0 \psi_2 \\ & = \psi_1^T C^\dagger \Gamma_0^\ast \Gamma_0 \psi_2 \\ & = \psi_1^T C \psi_2 \end{aligned} \,.

For a Clifford algebra representation on ν\mathbb{C}^\nu as in prop. , then the map

() TC:Mat 2 ν×1()Mat 1×2 ν() (-)^T C \;\colon\; Mat_{2^\nu \times 1}(\mathbb{C}) \longrightarrow Mat_{1 \times 2^\nu}(\mathbb{C})

(from complex column vectors to complex row vectors) which is given by transposition followed by matrix multiplication from the right by the charge conjugation matrix according to prop. is called the Majorana conjugation.


In dimensions d=4,8,9,10,11d = 4,8,9,10,11 a spinor ψ (2 ν)\psi \in \mathbb{C}^{(2^\nu)} is a real spinor according to def. with respect to the real structure from prop. , precisely if

ψ=CΓ 0 Tψ * \psi = C \Gamma_0^T \psi^\ast

(as e.g. in Castellani-D’Auria-Fré, (II.7.22)),

This is equivalent to the condition that the Majorana conjugate (def. ) coincides with the Dirac conjugate (def. ) on ψ\psi:

ψ TC=ψ Γ 0 \psi^T C = \psi^\dagger \Gamma_0

and such ψ\psi are called Majorana spinors.

This condition is also equivalent to the condition that

(ψ,)=ψ,, (\psi,-) = \langle \psi,-\rangle \,,

where on the left we have the complex bilinear form of prop. and on the right the hermitian form from prop. .


The first statement is immediate. The second follows by applying the transpose to the first equation, and using that C 1=C TC^{-1} = C^T (from prop. ). Finally the last statement follows from this by prop. .

Of course we may combine the condition Majorana and Weyl conditions on spinors:


In the even dimensions among those dimensions dd for which the Majorana projection operator (real structure) JJ exists (prop. ) also the chirality projection operator Γ d\Gamma_{d} exists (def. ). Then we may ask that a Dirac spinor ψ\psi is both Majorana, J(ψ)=ψJ(\psi) = \psi, as well as Weyl, Γ dψ=±iψ\Gamma_d \psi = \pm i \psi. If this is the case, it is called a Majorana-Weyl spinor, and the sub-representation these form is a called a Majorana-Weyl representation.


In Lorentzian signature for 4d114 \leq d \leq 11, then Majorana-Weyl spinors (def. ) exist precisely only in d=10d = 10.


According to prop. Majorana spinors in the given range exist for d{4,8,9,10,11}d \in \{4,8,9,10,11\}. Hence the even dimensions among these are d{4,8,10}d \in \{4,8,10\}.

Majorana-Weyl spinors clearly exist precisely if the two relevant projection operators in these dimensions commute with each other, i.e. if

[J,ϵΓ 0Γ d1]=0 [J, \epsilon \Gamma_0 \cdots \Gamma_{d-1}] = 0


ϵ={1 |νodd i |νeven. \epsilon = \left\{ \array{ 1 & \vert \; \nu \, \text{odd} \\ i & \vert \; \nu \, \text{even} } \right. \,.

with d=2νd = 2\nu (from the proof of prop. ).

By prop. all the Γ a\Gamma_a commute or all anti-commute with JJ. Since the product Γ 0Γ d1\Gamma_0 \cdots \Gamma_{d-1} contains an even number of these, it commutes with JJ. It follows that JJ commutes with Γ d\Gamma_d precisely if it commutes with ϵ\epsilon. Now since JJ is conjugate-linear, this is the case precisely if ϵ=1\epsilon = 1, hence precisely if d=2νd = 2\nu with ν\nu odd.

This is the case for d=10=25d = 10 = 2 \cdot 5, but not for d=8=24d = 8 = 2 \cdot 4 neither for d=4=22d = 4 = 2 \cdot 2.

Pseudo-Majorana spinors and Symplectic structure

In d=5d = 5, for example, the reality/Majorana condition

ψ=CΓ 0 Tψ * \psi = C \Gamma_0^T \psi^\ast

from prop. has no solution. But if we consider the direct sum of two copies of the complex spinor representation space, with elements denoted ψ 1\psi_1 and ψ 2\psi_2, then the following condition does have a solution

CΓ 0 Tψ 1 *=ψ 2CΓ 0 Tψ 2 *=+ψ 1 C \Gamma_0^T \psi_1^\ast = -\psi_2 \;\;\;\; C \Gamma_0^T \psi_2^\ast = +\psi_1

(e.g Castellani-D’Auria-Fré, II.8.41). Comparison with prop. and def. shows that this exhibits a quaternionic structure on the original complex spinor space, and hence a real structure on its direct sum double.

The spinor bilinear pairing to antisymmetric pp-tensors

We now discuss, in the component expressions established above, the complex bilinear pairing operations that take a pair of Majorana spinors to a vector, and more generally to an antisymmetric rank pp-tensor. These operations are all of the form

ψϵ(ψ¯Γ a 1a pψ), \psi \mapsto \epsilon \left( \overline{\psi} \Gamma^{a_1 \cdots a_p} \psi \right) \,,

where ϵ\epsilon \in \mathbb{C} is some prefactor, constrained such as to make the whole expression be real, hence such as to make this the components of an element in p d1,1\wedge^p \mathbb{R}^{d-1,1}.

For a Spin(d1,1)Spin(d-1,1) representation VV as in prop. , with real/Majorana structure as in prop. , write

SV S \hookrightarrow V

for the subspace of Majorana spinors, regarded as a real vector space.

Recall, by prop. , that on Majorana spinors the Majorana conjugate () TC(-)^T C coincides with the Dirac conjugate ()¯() Γ 0\overline{(-)} \coloneqq (-)^\dagger \Gamma_0 . Therefore we write ()¯\overline{(-)} in the following for the conjugation of Majorana spinors, unambiguously defined.


For a Spin(d1,1)Spin(d-1,1) representation VV as in prop. , with real/Majorana structure as in prop. , let

()¯Γ():S×S d1,1 \overline{(-)}\Gamma (-) \;\colon\; S \times S \longrightarrow \mathbb{R}^{d-1,1}

be the function that takes Majorana spinors ψ 1\psi_1, ψ 2\psi_2 to the vector with components

ψ¯ 1Γ aψ 2ψ 1 TCΓ aψ 2. \overline{\psi}_1\Gamma^a \psi_2 \coloneqq \psi_1^T C \Gamma^a \psi_2 \,.

Now the crucial property for the construction of spacetime supersymmetry super Lie algebras below is the following


For a Spin(d1,1)Spin(d-1,1) representation VV as in prop. , with real/Majorana structure as in prop. , then spinor to vector pairing operation of def. satisfies the following properties: it is

  1. symmetric:

    ψ¯ 1Γψ 2=ψ¯ 2Γψ 1 \overline{\psi}_1 \Gamma \psi_2 = \overline{\psi}_2 \Gamma \psi_1
  2. component-wise real-valued (i.e. it indeed takes values in d d\mathbb{R}^d \subset \mathbb{C}^d);

  3. Spin(d1,1)Spin(d-1,1)-equivariant: for gSpin(d1,1)g \in Spin(d-1,1) then

    g()¯Γ(g())=g(()¯Γ()). \overline{g(-)}\Gamma(g(-)) = g(\overline{(-)}\Gamma(-)) \,.

Regarding the first point, we need to show that for all aa then CΓ aC \Gamma_a is a symmetric matrix. Indeed:

(CΓ a) T =Γ a TC T =±Γ a TC =±±CΓ a =CΓ a, \begin{aligned} (C \Gamma_a)^T & = \Gamma_a^T C^T \\ & = \pm \Gamma_a^T C \\ & = \pm \pm C \Gamma_a \\ & = C \Gamma_a \end{aligned} \,,

where the first sign picked up is from C T=±CC^T = \pm C, while the second is from Γ a TC=±CΓ a\Gamma_a^T C = \pm C \Gamma_a (according to prop. ). Imposing the condition in prop. one finds that these signs agree, and hence cancel out.

(In van Proeyen99 this is part of table 1, in (Castellani-D’Auria-Fré) this is implicit in equation (II.2.32a).)

With this the second point follows together with the relation ψ TC=ψ Γ 0\psi^T C = \psi^\dagger \Gamma_0 for Majorana spinors ψ\psi (prop. ) and using the conjugate-symmetry of the hermitian form ,=() Γ 0()\langle -,-\rangle = (-)^\dagger \Gamma_0 (-) as well as the hermiticity of Γ 0Γ a\Gamma_0 \Gamma_a (both from prop. ):

(ψ¯ 1Γ aψ 2) * =(ψ 1 TCΓ aψ 2) * =(ψ 1 Γ 0Γ aψ 2) * =ψ 2 (Γ 0Γ a) ψ 1 =ψ 2 Γ 0Γ aψ 1 =ψ¯ 2Γ aψ 1. \begin{aligned} (\overline{\psi}_1 \Gamma_a \psi_2)^\ast &= (\psi_1^T C \Gamma_a \psi_2)^\ast \\ & = (\psi_1^\dagger \Gamma_0 \Gamma^a \psi_2)^\ast \\ & = \psi_2^\dagger (\Gamma_0 \Gamma^a)^\dagger \psi_1 \\ & = \psi_2^\dagger \Gamma_0 \Gamma^a \psi_1 \\ & = \overline{\psi}_2 \Gamma_a \psi_1 \end{aligned} \,.

Regarding the third point: By prop. and prop. we get

(g(ψ 1),Γ ag(ψ 2)) =g(ψ 1),Γ ag(ψ 2) =ψ 1,(Γ 0 1g Γ 0)Γ agψ 2 =ψ 1(g 1Γ ag)ψ 2, \begin{aligned} (g(\psi_1), \Gamma_a g(\psi_2)) & = \langle g(\psi_1),\Gamma_a g(\psi_2)\rangle \\ & = \langle \psi_1, (\Gamma_0^{-1}g^\dagger\Gamma_0) \Gamma_a g \psi_2 \rangle \\ & = \langle \psi_1 (g^{-1} \Gamma_a g) \psi_2 \rangle \end{aligned} \,,

where we used that Γ 0 1() Γ 0\Gamma_0^{-1}(-)^\dagger \Gamma_0 is the adjoint with respect to the hermitian form ,=() Γ 0()\langle -,-\rangle = (-)^\dagger \Gamma_0 (-) (prop. ) and that gg is unitary with respect to this hermitian form by prop. .

(In (Castellani-D’Auria-Fré) this third statement implicit in equations (II.2.35)-(II.2.39).)


Proposition implies that adding a copy of SS to the Poincaré Lie algebra in odd degree, then the pairing of def. is a consistent extension of the Lie bracket of the latter to a super Lie algebra. This is the super Poincaré Lie algebra, to which we come below.


For a Spin(d1,1)Spin(d-1,1) representation VV as in prop. , with real/Majorana structure as in prop. , let

()¯ΓΓ():S×S 2 d \overline{(-)}\Gamma\Gamma (-) \;\colon\; S \times S \longrightarrow \wedge^2 \mathbb{C}^d

be the function that takes Majorana spinors ψ 1\psi_1, ψ 2\psi_2 to the skew-symmetric rank 2-tensor with components

ψ¯ 1Γ abψ 2iψ 1 TC12(Γ aΓ bΓ bΓ a)ψ 2, \overline{\psi}_1\Gamma^{a b} \psi_2 \coloneqq i \psi_1^T C \tfrac{1}{2}(\Gamma^a \Gamma^b - \Gamma^b \Gamma^a) \psi_2 \,,

For ψ 1=ψ 2=ψ\psi_1 = \psi_2 = \psi then the pairing in prop. is real

a,biψ¯Γ abψ \underset{a,b}{\forall} \;\;\;\; i \overline{\psi} \Gamma^{a b} \psi \in \mathbb{R} \subset \mathbb{C}

and Spin(d1,1)Spin(d-1,1)-equivariant.


The equivariance follows exactly as in the proof of prop. .

The reality is checked by direct computation as follows:

(ψ¯ 1Γ aΓ bψ 2) * =(ψ 1 Γ aΓ bψ 2) * =ψ 2 (Γ 0Γ aΓ b) ψ 1 =ψ 2 Γ 0Γ aΓ bψ 1 =ψ¯ 2Γ aΓ bψ 1, \begin{aligned} (\overline{\psi}_1 \Gamma_a \Gamma_b \psi_2)^\ast & = (\psi_1^\dagger \Gamma_a \Gamma_b \psi_2)^\ast \\ & = \psi_2^\dagger (\Gamma_0 \Gamma_a \Gamma_b)^\dagger \psi_1 \\ & = -\langle \psi_2^\dagger \Gamma_0 \Gamma_a \Gamma_b \psi_1 \rangle \\ & = -\overline{\psi}_2 \Gamma_a \Gamma_b \psi_1 \end{aligned} \,,

where for the identification (Γ 0Γ aΓ b) =Γ 0Γ aΓ b(\Gamma_0 \Gamma_a \Gamma_b)^\dagger = - \Gamma_0 \Gamma_a \Gamma_b we used the properties in prop. .

Hence for ψ 1=ψ 2\psi_1 = \psi_2 then

(ψ¯Γ aΓ bψ) *=ψ¯Γ aΓ bψ (\overline{\psi} \Gamma_a \Gamma_b \psi)^\ast = - \overline{\psi} \Gamma_a \Gamma_b \psi

and so this sign cancels against the sign in i *=ii^\ast = -i.



The following pairings are real and Spin(d1,1)Spin(d-1,1)-equivariant:

ψ¯Γ aψ i ψ¯Γ a 1a 2ψ i ψ¯Γ a 1a 2a 3ψ ψ¯Γ a 1a 4ψ ψ¯Γ a 1a 5ψ i ψ¯Γ a 1a 6ψ i ψ¯Γ a 1a 7ψ . \begin{aligned} & \overline{\psi} \Gamma_a \psi \\ i & \overline{\psi}\Gamma_{a_1 a_2} \psi \\ i & \overline{\psi} \Gamma_{a_1 a_2 a_3} \psi \\ & \overline{\psi} \Gamma_{a_1 \cdots a_4} \psi \\ & \overline{\psi} \Gamma_{a_1 \cdots a_5} \psi \\ i & \overline{\psi} \Gamma_{a_1 \cdots a_6} \psi \\ i & \overline{\psi} \Gamma_{a_1 \cdots a_7} \psi \\ & \vdots \end{aligned} \,.

The equivariance follows as in the proof of prop. .

Regarding reality:

Using that all the Γ a\Gamma_a are hermitian with respect ()¯()\overline{(-)}(-) (prop. ) we have

(ψ¯Γ a 1a pψ) * =ψ¯Γ a pa 1ψ =(1) p(p1)/2ψ¯Γ a 1a pψ. \begin{aligned} \left( \overline{\psi} \Gamma_{a_1 \cdots a_p} \psi \right)^\ast & = \overline{\psi} \Gamma_{a_p \cdots a_1} \psi \\ &= (-1)^{p(p-1)/2} \overline{\psi} \Gamma_{a_1 \cdots a_p} \psi \end{aligned} \,.

Example: Majorana spinors in dimensions 11, 10, and 9

We spell out some of the above constructions and properties for Majorana spinors in Minkowski spacetimes of dimensions 11, 10 and 9, and discuss some relations between these. These spinor structures are relevant for spinors in 11-dimensional supergravity and type II supergravity in 10d and 9d, as well as to the relation between these via Kaluza-Klein compactification and T-duality.


Let {γ a}\{\gamma_a\} be any Dirac representation of Spin(8,1)Spin(8,1) according to prop. . By the same logic as in the proof of prop. we get from this the Dirac representations in dimensions 9+1 and 10+1 by setting

Γ a8(0 γ a γ a 0),Γ 9(0 id id 0),Γ 10(iid 0 0 iid). \Gamma_{a \leq 8} \coloneqq \left( \array{ 0 & \gamma_a \\ \gamma_a & 0 } \right) \;\,,\;\; \Gamma_{9} \coloneqq \left( \array{ 0 & id \\ -id & 0 } \right) \;\,,\;\; \Gamma_{10} \coloneqq \left( \array{ i id & 0 \\ 0 & -i id } \right) \,.

By prop. the Dirac representation in d=11d = 11 has complex dimension 2 10/2=2 5=322^{10/2} = 2^{5} = 32. By prop. and prop. this representation carries a real structure and hence gives a real/Majorana spin representation S 32S \hookrightarrow \mathbb{C}^{32} of Spin(10,1)Spin(10,1) of real dimension 32. This representation often just called “32\mathbf{32}”. This way the corresponding super-Minkowski spacetime (remark ) is neatly written as

10,1|32 \mathbb{R}^{10,1\vert \mathbf{32}}

which thus serves to express both, the real dimension of the space of odd-graded coordinate functions at every point on it, as well as the way that the Spin(10,1)Spin(10,1)-cover of the Lorentz group SO(10,1)SO(10,1) acts on these. This is the local model space for super spacetimes in 11-dimensional supergravity.

As we regard 32\mathbb{C}^{32} instead as the Dirac representation of Spin(9,1)Spin(9,1) via def. , then it decomposes into to chiral halfs, each of complex dimension 16. This is the direct sum decomposition in terms of which the block decomposition of the above Clifford matrices is given.

Since in 10d the Weyl condition is compatible with the Majorana condition (by prop. ), the real Majorana representation 32\mathbf{32} correspondingly decomposes as a direct sum of two real representations of dimension 16, which often are denoted 16\mathbf{16} and 16¯\overline{\mathbf{16}}. Hence as real/Majorana Spin(9,1)Spin(9,1)-representations there is a direct sum decomposition

321616¯. \mathbf{32} \simeq \mathbf{16} \oplus \overline{\mathbf{16}} \,.

The corresponding super Minkowski spacetime (remark )

9,1|16+16¯ \mathbb{R}^{9,1\vert \mathbf{16} + \overline{\mathbf{16}}}

is said to be of “type IIA”, since this is the local model space for superspacetimes in type IIA supergravity. This is as opposed to 9,1|1616\mathbb{R}^{9,1\vert \mathbf{16}\oplus \mathbf{16}}, which is “type IIB” and in contrast to 9,1|16\mathbb{R}^{9,1\vert \mathbf{16}} which is “heterotic” (the local model space for heterotic supergravity).

Now the Dirac-Weyl representation for Spin(8,1)Spin(8,1) is of complex dimension d=2 8/2=2 4=16d = 2^{8/2} = 2^4 = 16. By prop. and prop. this also admits real structure, and hence gives a Majorana representation for Spin(8,1)Spin(8,1), accordingly denoted 16\mathbf{16}. Notice that this is Majorana-Weyl.

We want to argue that both the 16\mathbf{16} and the 16¯\overline{\mathbf{16}} of Spin(9,1)Spin(9,1) become isomorphic to the single 16\mathbf{16} of Spin(8,1)Spin(8,1) under forming the restricted representation along the inclusion Spin(8,1)Spin(9,1)Spin(8,1)\hookrightarrow Spin(9,1) (the one fixed by the above choice of components).

For this it is sufficient to see that Γ 9\Gamma_9, which as a complex linear map goes Γ 9:1616¯\Gamma_9 \colon \mathbf{16} \longrightarrow \overline{\mathbf{16}} constitutes an isomorphism when regarded as a morphism in the category of representations of Spin(8,1)Spin(8,1).

Clearly it is a linear isomorphism, so it is sufficient that it is a homomorphism of Spin(8,1)Spin(8,1)-representations at all. But that’s clear since by the Clifford algebra relations Γ 9\Gamma_9 commutes with all Γ aΓ b\Gamma_a \Gamma_b for a,b{0,,8}a,b \in \{0,\cdots, 8\}.

Hence the branching rule for restricting the Weyl representation in 11d along the sequence of inclusions

Spin(8,1)Spin(9,1)Spin(10,1) Spin(8,1) \hookrightarrow Spin(9,1) \hookrightarrow Spin(10,1)


32Rep(Spin(10,1))1616¯Rep(Spin(9,1))1616Rep(Spin(8,1)). \underset{\in Rep(Spin(10,1))}\underbrace{\mathbf{32}} \mapsto \underset{\in Rep(Spin(9,1))}\underbrace{\mathbf{16} \oplus \overline{\mathbf{16}}} \mapsto \underset{\in Rep(Spin(8,1))}\underbrace{\mathbf{16} \oplus \mathbf{16}} \,.

If we write a Majorana spinor in 32\mathbf{32} as ϑ 32\vartheta \in \mathbb{C}^{32}, decomposed as a (1×32)(1 \times 32)-matrix as

ϑ=(ψ 1 ψ 2). \vartheta = \left( \array{ \psi_1 \\ \psi_2 } \right) \,.

and if we write for short

ψ 1=(ψ 1 0),ψ 2=(0 ψ 2) \psi_1 = \left( \array{ \psi_1 \\ 0 } \right) \,,\;\;\; \psi_2 = \left( \array{ 0 \\ \psi_2 } \right)

then this says that after restriction to Spin(9,1)Spin(9,1)-action then ψ 1\psi_1 becomes a Majorana spinor in the 16\mathbf{16}, and ψ 2\psi_2 a Majorana spinor in the 16¯\overline{\mathbf{16}}, and after further restriction to Spin(8,1)Spin(8,1)-action then either comes a Majorana spinor in one copy of 16\mathbf{16}.

The type IIA spinor-to-vector pairing is just that of 11d under this re-interpretation. We find:


The type IIA spinor-to-vector pairing is given by

(ψ 1 ψ 2)¯Γ a IIA(ψ 1 ψ 2) ={ψ¯ 1γ aψ 1+ψ¯ 2γ aψ 2 |a8 ψ¯ 1ψ 1ψ¯ 2ψ 2 |a=9. \begin{aligned} \overline{\left(\array{\psi_1 \\ \psi_2}\right)} \Gamma_a^{IIA} \left( \array{\psi_1 \\ \psi_2} \right) & = \left\{ \array{ \overline{\psi}_1 \gamma_a \psi_1 + \overline{\psi}_2 \gamma_a \psi_2 & \vert \; a \leq 8 \\ \overline{\psi}_1 \psi_1 - \overline{\psi}_2 \psi_2 & \vert \; a = 9 } \right. \end{aligned} \,.

Using that on Majorana spinors the Majorana conjugate coincides with the Dirac conjugate (prop. ) and applying prop. we compute:

(ψ 1 ψ 2)¯Γ a IIA(ψ 1 ψ 2) (ψ 1 ψ 2)¯Γ a(ψ 1 ψ 2) =(ψ 1 ψ 2) Γ 0Γ a(ψ 1 ψ 2) ={(ψ 1 ψ 2) (γ 0γ a 0 0 γ 0γ a)(ψ 1 ψ 2) |a8 (ψ 1 ψ 2) (γ 0 0 0 γ 0)(ψ 1 ψ 2) |a=9 ={ψ¯ 1γ aψ 1+ψ¯ 2γ aψ 2 |a8 ψ¯ 1ψ 1ψ¯ 2ψ 2 |a=9. \begin{aligned} \overline{\left(\array{\psi_1 \\ \psi_2}\right)} \Gamma_a^{IIA} \left( \array{\psi_1 \\ \psi_2} \right) &\coloneqq \overline{\left(\array{\psi_1 \\ \psi_2}\right)} \Gamma_a \left( \array{\psi_1 \\ \psi_2} \right) \\ & = \left(\array{\psi_1 \\ \psi_2}\right)^\dagger \Gamma_0 \Gamma_a \left( \array{\psi_1 \\ \psi_2} \right) \\ & = \left\{ \array{ \left(\array{\psi_1 \\ \psi_2}\right)^\dagger \left( \array{ \gamma_0 \gamma_a & 0 \\ 0 & \gamma_0 \gamma_a } \right) \left( \array{\psi_1 \\ \psi_2} \right) & \vert \; a\leq 8 \\ \left(\array{\psi_1 \\ \psi_2}\right)^\dagger \left( \array{ \gamma_0 & 0 \\ 0 & -\gamma_0 } \right) \left( \array{\psi_1 \\ \psi_2} \right) & \vert \; a = 9 } \right. \\ & = \left\{ \array{ \overline{\psi}_1 \gamma_a \psi_1 + \overline{\psi}_2 \gamma_a \psi_2 & \vert \; a \leq 8 \\ \overline{\psi}_1 \psi_1 - \overline{\psi}_2 \psi_2 & \vert \; a = 9 } \right. \end{aligned} \,.

The type IIB spinor-to-vector pairing is

(ψ 1 ψ 2)¯Γ a IIB(ψ 1 ψ 2) ={ψ¯ 1γ aψ 1+ψ¯ 2γ aψ 2 |a8 ψ¯ 1ψ 1+ψ¯ 2ψ 2 |a=9 \begin{aligned} \overline{\left(\array{\psi_1 \\ \psi_2}\right)} \Gamma_a^{IIB} \left(\array{\psi_1 \\ \psi_2}\right) & = \left\{ \array{ \overline{\psi}_1 \gamma_a \psi_1 + \overline{\psi}_2 \gamma_a \psi_2 & \vert \; a \leq 8 \\ \overline{\psi}_1 \psi_1 + \overline{\psi}_2 \psi_2 & \vert \; a = 9 } \right. \end{aligned}

The type II pairing spinor-to-vector pairing is obtained from the type IIA pairing of prop. by replacing all bottom right matrix entries (those going 16¯16¯\overline{\mathbf{16}}\to \overline{\mathbf{16}} by the corresponding top left entries (those going 1616\mathbf{16} \to \mathbf{16} )). Notice that in fact all these block entries are the same, except for the one at a=9a = 9, where they simply differ by a sign. This yields the claim.

Notice also the following relation between the different pairing in dimensions 11, 10 and 9:


The (9,10)(9,10)-component of the spinor-to-bivector pairing (def. ) in 11d equals the 9-component of the type IIB spinor-to-vector pairing

i(ψ 1 ψ 2)¯Γ 9Γ 10(ψ 1 ψ 2) =(ψ 1 ψ 2)¯Γ 9 IIB(ψ 1 ψ 2) \begin{aligned} i \overline{\left(\array{\psi_1 \\ \psi_2}\right)} \Gamma_9 \Gamma_{10} \left(\array{\psi_1 \\ \psi_2}\right) & = \overline{\left(\array{\psi_1 \\ \psi_2}\right)} \Gamma_9^{IIB} \left(\array{\psi_1 \\ \psi_2}\right) \end{aligned}

Using prop. and prop. we compute:

i(ψ 1 ψ 2)¯Γ 9Γ 10(ψ 1 ψ 2) =i(ψ 1 ψ 2) Γ 0Γ 9Γ 10(ψ 1 ψ 2) =i(ψ 1 ψ 2) (iγ 0 0 0 iγ 0)(ψ 1 ψ 2) =ψ¯ 1ψ 1+ψ¯ 2ψ 2 =(ψ 1 ψ 2)¯Γ 9 IIB(ψ 1 ψ 2) \begin{aligned} i\, \overline{\left(\array{\psi_1 \\ \psi_2}\right)} \Gamma_9 \Gamma_{10} \left(\array{\psi_1 \\ \psi_2}\right) & = i \left(\array{\psi_1 \\ \psi_2}\right)^\dagger \Gamma_0\Gamma_9 \Gamma_{10} \left(\array{\psi_1 \\ \psi_2}\right) \\ & = i \, \left(\array{\psi_1 \\ \psi_2}\right)^\dagger \left( \array{ -i \gamma_0 & 0 \\ 0 & -i \gamma_0 } \right) \left(\array{\psi_1 \\ \psi_2}\right) \\ & = \overline{\psi}_1 \psi_1 + \overline{\psi}_2 \psi_2 \\ & = \overline{\left(\array{\psi_1 \\ \psi_2}\right)} \Gamma_9^{IIB} \left(\array{\psi_1 \\ \psi_2}\right) \end{aligned}

The following is an evident variant of the extensions considered in (CAIB 99, FSS 13).


We have

  1. The 11d N=1N = 1 super-Minkowski spacetime 10,1|32\mathbb{R}^{10,1\vert \mathbf{32}} (def. ) is the central super Lie algebra extension of the 10d type IIA super-Minkowski spacetime 9,1|16+16¯\mathbb{R}^{9,1\vert \mathbf{16}+ \overline{\mathbf{16}}} by the 2-cocycle

    c 2ψ¯Γ 10ψCE( 9,1|16+16¯) c_2 \coloneqq \overline{\psi} \wedge \Gamma_{10} \psi \;\;\; \in CE(\mathbb{R}^{9,1\vert \mathbf{16} + \overline{\mathbf{16}}})

    (from def. ).

  2. The 10d type IIA super-Minkowski spacetime 9,1|16+16¯\mathbb{R}^{9,1\vert \mathbf{16}+ \overline{\mathbf{16}}} is central super Lie algebra extension of th 9d N=2N = 2 super-Minkowski spacetime by the 2-cocycle given by the type IIA spinor-to-vector pairing

    c 2 IIA(ψ 1 ψ 2)¯Γ 9 IIA(ψ 1 ψ 2)CE( 8,1|16+16) c_2^{IIA} \;\coloneqq\; \overline{\left(\array{\psi_1 \\ \psi_2}\right)} \wedge \Gamma_9^{IIA} \left( \array{\psi_1 \\ \psi_2} \right) \;\;\;\in CE(\mathbb{R}^{8,1\vert \mathbf{16}+ \mathbf{16}})

    (from prop. ).

  3. The 10d type IIB super-Minkowski spacetime 9,1|16+16\mathbb{R}^{9,1\vert \mathbf{16}+ \mathbf{16}} is central super Lie algebra extension of th 9d N=2N = 2 super-Minkowski spacetime by the 2-cocycle given by the type IIB spinor-to-vector pairing

    c 2 IIB(ψ 1 ψ 2)¯Γ 9 IIB(ψ 1 ψ 2)CE( 8,1|16+16) c_2^{IIB} \;\coloneqq\; \overline{\left(\array{\psi_1 \\ \psi_2}\right)} \wedge \Gamma_9^{IIB} \left( \array{\psi_1 \\ \psi_2} \right) \;\;\; \in CE(\mathbb{R}^{8,1\vert \mathbf{16} + \mathbf{16}})

    (from prop. ).

In summary, we have the following diagram in the category of super L-infinity algebras

10,1|32 9,1|16+16 9,1|16+16¯ c 2 B 8,1|16+16 c 2 IIB c 2 IIA B B, \array{ && && \mathbb{R}^{10,1\vert \mathbf{32}} \\ && && \downarrow \\ \mathbb{R}^{9,1\vert \mathbf{16} + \mathbf{16}} && && \mathbb{R}^{9,1\vert \mathbf{16} + \overline{\mathbf{16}}} &\overset{c_2}{\longrightarrow}& B \mathbb{R} \\ & \searrow && \swarrow \\ && \mathbb{R}^{8,1\vert \mathbf{16} + \mathbf{16}} \\ & {}^{\mathllap{c_2^{IIB}}}\swarrow && \searrow^{\mathrlap{c_2^{IIA}}} \\ B \mathbb{R} && && B \mathbb{R} } \,,

where BB\mathbb{R} denotes the line Lie 2-algebra, and where each “hook”

𝔤^ 𝔤 ω 2 B \array{ \widehat{\mathfrak{g}} \\ \downarrow \\ \mathfrak{g} &\overset{\omega_2}{\longrightarrow}& B\mathbb{R} }

is a homotopy fiber sequence (because homotopy fibers of super L L_\infty-algebra cocycles are the corresponding extension that they classify, see at L-infinity algebra cohomology).


To see that the given 2-forms are indeed cocycles: they are trivially closed (by def. ), and so all that matters is that we have a well defined super-2-form in the first place. Since the ψ α\psi^\alpha are in bidegree (1,odd)(1,odd), they all commutes with each other (see at signs in supergeometry) and hece the condition is that the pairing is symmetric. This is the case by prop. .

Now to see the extensions. Notice that for 𝔤\mathfrak{g} any (super) Lie algebra (of finite dimension, for convenience), and for ω 2𝔤 *\omega \in \wedge^2\mathfrak{g}^\ast a Lie algebra 2-cocycle on it, then the Lie algebra extension 𝔤^\widehat{\mathfrak{g}} that this classifies is neatly characterized in terms of its dual Chevalley-Eilenberg algebra: that is simply the original CE algebra with one new generator ee (in degree (1,even)(1,even)) adjoined, and with the differential of ee taking to be ω\omega:

CE(𝔤^)=(CE(𝔤)e),de=ω). CE(\widehat{\mathfrak{g}}) = (CE(\mathfrak{g}) \otimes \langle e\rangle), d e = \omega) \,.

Hence in the case of ω=c 2 IIA\omega = c_2^{IIA} we identify the new generator with e 9e^9 and see that the equation de 9=c 2 IIAd e^9 = c_2^{IIA} is precisely what distinguishes the CE-algebra of 8,1|16+16\mathbb{R}^{8,1\vert \mathbf{16}+ \mathbf{16}} from that of 9,1|16+16¯\mathbb{R}^{9,1\vert \mathbf{16} + \overline{\mathbf{16}}}, by prop. and the fact that both spin representation have the same underlying space, by remark .

The other two cases are directly analogous.

Recall the following (e.g. from FSS 16 and references given there):


The cocycle for the higher WZW term of the Green-Schwarz sigma-model for the M2-brane is

μ M2iϑ¯Γ aΓ bϑe ae bCE( 10,1|32) \mu_{M2} \coloneqq i\,\overline{\vartheta} \wedge \Gamma_a \Gamma_b \vartheta \wedge e^a \wedge e^b \;\;\; \in CE(\mathbb{R}^{10,1\vert \mathbf{32}})

obtained from the spinor-to-bivector pairing of def. . (Here and in the following we are using the nation from remark .)

The cocycle for the WZW term of the Green-Schwarz sigma-model for the type IIA superstring is

μ IIAiϑ¯Γ aΓ 10ϑe a=i(ψ 1 ψ 2)¯Γ aΓ 10(ψ 1 ψ 2)CE( 9,1|16+16¯), \mu_{IIA} \coloneqq i\,\overline{\vartheta} \wedge \Gamma_a \Gamma_{10} \vartheta \wedge e^a = i\, \overline{\left( \array{\psi_1 \\ \psi_2} \right)} \Gamma_a \Gamma_{10} \left( \array{\psi_1 \\ \psi_2} \right) \;\;\; \in CE(\mathbb{R}^{9,1\vert \mathbf{16} + \overline{\mathbf{16}}}) \,,

i.e. this is the the e 10e^{10}-component of μ M2\mu_{M2} (“double dimensional reductionFSS 16):

μ IIA=(π 10) *μ M2. \mu_{IIA} = (\pi_{10})_\ast \mu_{M2} \,.

The e 9e^9-component of the cocycle for the IIA-superstring (def. ), regarded as an element in CE( 8,1|16+16)CE(\mathbb{R}^{8,1}\vert \mathbf{16} + \mathbf{16}), equals the 2-cocycle that defines the type IIB extension, according to prop. :

(π 9) *μ IIA=c 2 IIB. (\pi_9)_\ast \mu_{IIA} = c_2^{IIB} \,.

We have

(π 9) *μ IIA =i(ψ 1 ψ 2)¯Γ 9Γ 10(ψ 1 ψ 2) =(ψ 1 ψ 2)¯Γ 9 IIB(ψ 1 ψ 2) =c 2 IIB \begin{aligned} (\pi_9)_\ast \mu_{IIA} & = i\, \overline{\left( \array{\psi_1 \\ \psi_2} \right)} \Gamma_9 \Gamma_{10} \left( \array{\psi_1 \\ \psi_2} \right) \\ & = \overline{\left(\array{\psi_1 \\ \psi_2}\right)} \Gamma_9^{IIB} \left(\array{\psi_1 \\ \psi_2}\right) \\ & = c_2^{IIB} \end{aligned}

where the first equality is by def. , the second is the statement of prop. , while the third is from prop. .

Real spinor representations via Real alternative division algebras

We discuss a close relation between real spin representations and division algebras, due to Kugo-Townsend 82, Sudbery 84 and others: The real spinor representations in dimensions 3,4,6,103,4,6, 10 happen to have a particularly simple expression in terms of 2-by-2 Hermitian matrices (generalized Pauli matrices) over the four real normed division algebras: the real numbers \mathbb{R} themselves, the complex numbers \mathbb{C}, the quaternions \mathbb{H} and the octonions 𝕆\mathbb{O}. Derived from this also the real spinor representations in dimensions 4,5,7,114,5,7,11 have a fairly simple corresponding expression. We follow the streamlined discussion in Baez-Huerta 09 and Baez-Huerta 10.

Real alternative division algebras

To amplify the following pattern and to fix our notation for algebra generators, recall these definitions:


(complex numbers)

The complex numbers \mathbb{C} is the commutative algebra over the real numbers \mathbb{R} which is generated from one generators {e 1}\{e_1\} subject to the relation

  • (e 1) 2=1(e_1)^2 = -1.


The quaternions \mathbb{H} is the associative algebra over the real numbers which is generated from three generators {e 1,e 2,e 3}\{e_1, e_2, e_3\} subject to the relations

quaternion multiplication table
  1. for all ii

    (e i) 2=1(e_i)^2 = -1

  2. for (i,j,k)(i,j,k) a cyclic permutation of (1,2,3)(1,2,3) then

    1. e ie j=e ke_i e_j = e_k

    2. e je i=e ke_j e_i = -e_k

(graphics grabbed from Baez 02)



The octonions 𝕆\mathbb{O} is the nonassociative algebra over the real numbers which is generated from seven generators {e 1,,e 7}\{e_1, \cdots, e_7\} subject to the relations

octonion multiplication table
  1. for all ii

    (e i) 2=1(e_i)^2 = -1

  2. for e ie je ke_i \to e_j \to e_k an edge or circle in the diagram shown (a labeled version of the Fano plane) then

    1. e ie j=e ke_i e_j = e_k

    2. e je i=e ke_j e_i = -e_k

    and all relations obtained by cyclic permutation of the indices in these equations.

(graphics grabbed from Baez 02)

One defines the following operations on these real algebras:


For 𝕂{,,,𝕆}\mathbb{K} \in \{\mathbb{R}, \mathbb{C}, \mathbb{H}, \mathbb{O}\}, let

() *:𝕂𝕂 (-)^\ast \;\colon\; \mathbb{K} \longrightarrow \mathbb{K}

be the antihomomorphism of real algebras

(ra) *=ra * ,forr,a𝕂 (ab) *=b *a * ,fora,b𝕂 \begin{aligned} (r a)^\ast = r a^\ast &, \text{for}\;\; r \in \mathbb{R}, a \in \mathbb{K} \\ (a b)^\ast = b^\ast a^\ast &,\text{for}\;\; a,b \in \mathbb{K} \end{aligned}

given on the generators of def. , def. and def. by

(e i) *=e i. (e_i)^\ast = - e_i \,.

This operation makes 𝕂\mathbb{K} into a star algebra. For the complex numbers \mathbb{C} this is called complex conjugation, and in general we call it conjugation.

Let then

Re:𝕂 Re \;\colon\; \mathbb{K} \longrightarrow \mathbb{R}

be the function

Re(a)12(a+a *) Re(a) \;\coloneqq\; \tfrac{1}{2}(a + a^\ast)

(“real part”) and

Im:𝕂 Im \;\colon\; \mathbb{K} \longrightarrow \mathbb{R}

be the function

Im(a)12(aa *) Im(a) \;\coloneqq \; \tfrac{1}{2}(a - a^\ast)

(“imaginary part”).

It follows that for all a𝕂a \in \mathbb{K} then the product of a with its conjugate is in the real center of 𝕂\mathbb{K}

aa *=a *a𝕂 a a^\ast = a^\ast a \;\in \mathbb{R} \hookrightarrow \mathbb{K}

and we write the square root of this expression as

|a|aa * {\vert a\vert} \;\coloneqq\; \sqrt{a a^\ast}

called the norm or absolute value function

||:𝕂. {\vert -\vert} \;\colon\; \mathbb{K} \longrightarrow \mathbb{R} \,.

This norm operation clearly satisfies the following properties (for all a,b𝕂a,b \in \mathbb{K})

  1. |a|0\vert a \vert \geq 0;

  2. |a|=0a=0{\vert a \vert } = 0 \;\;\;\;\; \Leftrightarrow\;\;\;\;\;\; a = 0;

  3. |ab|=|a||b|{\vert a b \vert } = {\vert a \vert} {\vert b \vert}

and hence makes 𝕂\mathbb{K} a normed algebra.

Since \mathbb{R} is a division algebra, these relations immediately imply that each 𝕂\mathbb{K} is a division algebra, in that

ab=0a=0orb=0. a b = 0 \;\;\;\;\;\; \Rightarrow \;\;\;\;\;\; a = 0 \;\; \text{or} \;\; b = 0 \,.

Hence the conjugation operation makes 𝕂\mathbb{K} a real normed division algebra.


Sending each generator in def. , def. and def. to the generator of the same name in the next larger algebra constitutes a sequence of real star-algebra homomorphisms

𝕆. \mathbb{R} \hookrightarrow \mathbb{C} \hookrightarrow \mathbb{H} \hookrightarrow \mathbb{O} \,.

(Hurwitz theorem: \mathbb{R}, \mathbb{C}, \mathbb{H} and 𝕆\mathbb{O} are the normed real division algebras)

The four algebras of real numbers \mathbb{R}, complex numbers \mathbb{C}, quaternions \mathbb{H} and octonions 𝕆\mathbb{O} from def. , def. and def. respectively, which are real normed division algebras via def. , are, up to isomorphism, the only real normed division algebras that exist.


While hence the sequence from remark

𝕆 \mathbb{R} \hookrightarrow \mathbb{C} \hookrightarrow \mathbb{H} \hookrightarrow \mathbb{O}

is maximal in the category of real normed non-associative division algebras, there is a pattern that does continue if one disregards the division algebra property. Namely each step in this sequence is given by a construction called forming the Cayley-Dickson double algebra. This continues to an unbounded sequence of real nonassociative star-algebras

𝕆𝕊 \mathbb{R} \hookrightarrow \mathbb{C} \hookrightarrow \mathbb{H} \hookrightarrow \mathbb{O} \hookrightarrow \mathbb{S} \hookrightarrow \cdots

where the next algebra 𝕊\mathbb{S} is called the sedenions.

What actually matters for the following relation of the real normed division algebras to real spin representations is that they are also alternative algebras:


(alternative algebra)

Given any non-associative algebra AA, then the trilinear map

[,,]AAAA [-,-,-] \;-\; A \otimes A \otimes A \longrightarrow A

given on any elements a,b,cAa,b,c \in A by

[a,b,c](ab)ca(bc) [a,b,c] \coloneqq (a b) c - a (b c)

is called the associator (in analogy with the commutator [a,b]abba[a,b] \coloneqq a b - b a ).

If the associator is completely antisymmetric (in that for any permutation σ\sigma of three elements then [a σ 1,a σ 2,a σ 3]=(1) |σ|[a 1,a 2,a 3][a_{\sigma_1}, a_{\sigma_2}, a_{\sigma_3}] = (-1)^{\vert \sigma\vert} [a_1, a_2, a_3] for |σ|\vert \sigma \vert the signature of the permutation) then AA is called an alternative algebra.

If the characteristic of the ground field is different from 2, then alternativity is readily seen to be equivalent to the conditions that for all a,bAa,b \in A then

(aa)b=a(ab)and(ab)b=a(bb). (a a)b = a (a b) \;\;\;\;\; \text{and} \;\;\;\;\; (a b) b = a (b b) \,.

We record some basic properties of associators in alternative star-algebras that we need below:


(properties of alternative star algebras)

Let AA be an alternative algebra (def. ) which is also a star algebra. Then

  1. the associator vanishes when at least one argument is real

    [Re(a),b,c] [Re(a),b,c]
  2. the associator changes sign when one of its arguments is conjugated

    [a,b,c]=[a *,b,c]; [a,b,c] = -[a^\ast,b,c] \,;
  3. the associator vanishes when one of its arguments is the conjugate of another:

    [a,a *,b]=0; [a,a^\ast, b] = 0 \,;
  4. the associator is purely imaginary

    Re([a,b,c])=0. Re([a,b,c]) = 0 \,.

That the associator vanishes as soon as one argument is real is just the linearity of an algebra product over the ground ring.

Hence in fact

[a,b,c]=[Im(a),Im(b),Im(c)]. [a,b,c] = [Im(a), Im(b), Im(c)] \,.

This implies the second statement by linearity. And so follows the third statement by skew-symmetry:

[a,a *,b]=[a,a,b]=0. [a,a^\ast,b] = -[a,a,b] = 0 \,.

The fourth statement finally follows by this computation:

[a,b,c] * =[c *,b *,a *] =[c,b,a] =[a,b,c]. \begin{aligned} [a,b,c]^\ast & = -[c^\ast, b^\ast, a^\ast] \\ & = -[c,b,a] \\ & = -[a,b,c] \end{aligned} \,.

Here the first equation follows by inspection and using that (ab) *=b *a *(a b)^\ast = b^\ast a^\ast, the second follows from the first statement above, and the third is the ant-symmetry of the associator.

It is immediate to check that:


(\mathbb{R}, \mathbb{C}, \mathbb{H} and 𝕆\mathbb{O} are real alternative algebras)

The real algebras of real numbers, complex numbers, def. ,quaternions def. and octonions def. are alternative algebras (def. ).


Since the real numbers, complex numbers and quaternions are associative algebras, their associator vanishes identically. It only remains to see that the associator of the octonions is skew-symmetric. By linearity it is sufficient to check this on generators. So let e ie je ke_i \to e_j \to e_k be a circle or a cyclic permutation of an edge in the Fano plane. Then by definition of the octonion multiplication we have

(e ie j)e j =e ke j =e je k =e i =e i(e je j) \begin{aligned} (e_i e_j) e_j &= e_k e_j \\ &= - e_j e_k \\ & = -e_i \\ & = e_i (e_j e_j) \end{aligned}

and similarly

(e ie i)e j =e j =e ke i =e ie k =e i(e ie j). \begin{aligned} (e_i e_i ) e_j &= - e_j \\ &= - e_k e_i \\ &= e_i e_k \\ &= e_i (e_i e_j) \end{aligned} \,.

The analog of the Hurwitz theorem (prop. ) is now this:


(\mathbb{R}, \mathbb{C}, \mathbb{H} and 𝕆\mathbb{O} are precisely the alternative real division algebras)

The only division algebras over the real numbers which are also alternative algebras (def. ) are the real numbers themselves, the complex numbers, the quaternions and the octonions from prop. .

This is due to (Zorn 30).

For the following, the key point of alternative algebras is this equivalent characterization:


(alternative algebra detected on subalgebras spanned by any two elements)

A nonassociative algebra is alternative, def. , precisely if the subalgebra? generated by any two elements is an associative algebra.

This is due to Emil Artin, see for instance (Schafer 95, p. 18).

Proposition is what allows to carry over a minimum of linear algebra also to the octonions such as to yield a representation of the Clifford algebra on 9,1\mathbb{R}^{9,1}. This happens in the proof of prop. below.

So we will be looking at a fragment of linear algebra over these four normed division algebras. To that end, fix the following notation and terminology:


(hermitian matrices with vaues in real normed division algebras)

Let 𝕂\mathbb{K} be one of the four real normed division algebras from prop. , hence equivalently one of the four real alternative division algebras from prop. .

Say that an n×nn \times n matrix with coefficients in 𝕂\mathbb{K}, AMat n×n(𝕂)A\in Mat_{n\times n}(\mathbb{K}) is a hermitian matrix if the transpose matrix (A t) ijA ji(A^t)_{i j} \coloneqq A_{j i} equals the componentwise conjugated matrix (def. ):

A t=A *. A^t = A^\ast \,.

Hence with the notation

() (() t) * (-)^\dagger \coloneqq ((-)^t)^\ast

then AA is a hermitian matrix precisely if

A=A . A = A^\dagger \,.

We write Mat 2×2 her(𝕂)Mat_{2 \times 2}^{her}(\mathbb{K}) for the real vector space of hermitian matrices.


(trace reversal)

Let AMat 2×2 her(𝕂)A \in Mat_{2 \times 2}^{her}(\mathbb{K}) be a hermitian 2×22 \times 2 matrix as in def. . Its trace reversal is the result of subtracting its trace times the identity matrix:

A˜A(trA)1 n×n. \tilde A \;\coloneqq\; A - (tr A) 1_{n\times n} \,.

Spacetime in dimensions 3,4,6 and 10

We discuss how Minkowski spacetime of dimension 3,4,6 and 10 is naturally expressed in terms of the real normed division algebras 𝕂\mathbb{K} from prop. , equivalently the real alternative division algebras from prop. .


(Minkowski spacetime via hermitian matrices in real normed division algebras)

Let 𝕂\mathbb{K} be one of the four real normed division algebras from prop. , hence one of the four real alternative division algebras from prop. .

There is a isomorphism (of real inner product spaces) between Minkowski spacetime (def. ) of dimension

d=2+dim (𝕂) d = 2 + dim_{\mathbb{R}}(\mathbb{K})


  1. 2,1\mathbb{R}^{2,1} for 𝕂=\mathbb{K} = \mathbb{R};

  2. 3,1\mathbb{R}^{3,1} for 𝕂=\mathbb{K} = \mathbb{C};

  3. 5,1\mathbb{R}^{5,1} for 𝕂=\mathbb{K} = \mathbb{H};

  4. 9,1\mathbb{R}^{9,1} for 𝕂=𝕆\mathbb{K} = \mathbb{O}.

and the real vector space of 2×22 \times 2 hermitian matrices over 𝕂\mathbb{K} (def. ) equipped with the inner product whose norm-square is the negative of the determinant operation on matrices:

dim (𝕂)+1,1(Mat 2×2 her(𝕂),det). \mathbb{R}^{dim_{\mathbb{R}}(\mathbb{K})+1,1} \;\simeq\; \left(Mat_{2 \times 2}^{her}(\mathbb{K}), -det \right) \,.

As a linear map this is given by

(x 0,x 1,,x d1)(x 0+x 1 y y * x 0x 1)withyx 21+x 3e 1+x 4e 2++x 2+dim (𝕂)e dim (𝕂)1. (x_0, x_1, \cdots, x_{d-1}) \;\mapsto\; \left( \array{ x_0 + x_1 & y \\ y^\ast & x_0 - x_1 } \right) \;\;\; \text{with}\; y \coloneqq x_2 1 + x_3 e_1 + x_4 e_2 + \cdots + x_{2 + dim_{\mathbb{R}(\mathbb{K})}} \,e_{dim_{\mathbb{R}}(\mathbb{K})-1} \,.

Under this identification the operation of trace reversal from def. corresponds to time reversal in that

(x 0+x 1 y y * x 0x 1)˜=(x 0+x 1 y y * x 0x 1). \widetilde{ \left( \array{ x_0 + x_1 & y \\ y^\ast & x_0 - x_1 } \right) } \;=\; \left( \array{ -x_0 + x_1 & y \\ y^\ast & -x_0 - x_1 } \right) \,.

This is immediate from the nature of the conjugation operation () *(-)^\ast from def. :

det(x 0+x 1 y y * x 0x 1) =(x 0+x 1)(x 0x 1)+yy * =(x 0) 2+a=1d1(x a) 2. \begin{aligned} - det \left( \array{ x_0 + x_1 & y \\ y^\ast & x_0 - x_1 } \right) & = -(x_0 + x_1)(x_0 - x_1) + y y^\ast \\ & = -(x_0)^2 + \underoverset{a = 1}{d-1}{\sum} (x_a)^2 \end{aligned} \,.

By direct computation one finds:


In terms of the trace reversal operation ()˜\widetilde{(-)} from def. , the determinant operation on hermitian matrices (def. ) has the following alternative expression

det(A) =AA˜ =A˜A. \begin{aligned} -det(A) & = A \tilde A \\ & = \tilde A A \end{aligned} \,.

and the Minkowski inner product has the alternative expression

η(A,B)=12Re(tr(AB˜))=12Re(tr(A˜B)). \eta(A,B) = \tfrac{1}{2}Re(tr(A \tilde B)) = \tfrac{1}{2} Re(tr(\tilde A B)) \,.

(Baez-Huerta 09, prop. 5)

Real spinors in dimensions 3, 4, 6 and 10

We now discuss how real spin representations in dimensions 3,4, 6 and 10 are naturally induced from linear algebra over the four real alternative division algebras.

In particular we establish the following table of exceptional isomorphisms of spin groups:

exceptional spinors and real normed division algebras

AA\phantom{AA}spin groupnormed division algebra\,\, brane scan entry
3=2+13 = 2+1Spin(2,1)SL(2,)Spin(2,1) \simeq SL(2,\mathbb{R})A\phantom{A} \mathbb{R} the real numberssuper 1-brane in 3d
4=3+14 = 3+1Spin(3,1)SL(2,)Spin(3,1) \simeq SL(2, \mathbb{C})A\phantom{A} \mathbb{C} the complex numberssuper 2-brane in 4d
6=5+16 = 5+1Spin(5,1)Spin(5,1) \simeq SL(2,H)A\phantom{A} \mathbb{H} the quaternionslittle string
10=9+110 = 9+1Spin(9,1) {\simeq}SL(2,O)A\phantom{A} 𝕆\mathbb{O} the octonionsheterotic/type II string

Prop. immediately implies that for 𝕂{,,}\mathbb{K} \in \{\mathbb{R}, \mathbb{C}, \mathbb{H}\} then there is a monomorphism from the special linear group SL(2,𝕂)SL(2,\mathbb{K}) (see at SL(2,H) for the definition in the quaternionic case) to the spin group in the given dimension:

SL(2,𝕂)Spin(dim (𝕂)+1,1) SL(2,\mathbb{K}) \hookrightarrow Spin(dim_{\mathbb{R}(\mathbb{K} )} +1 ,1)

given by

AA()A . A \mapsto A(-)A^\dagger \,.

This preserves the determinant, and hence the Lorentz form, by the multiplicative property of the determinant:

det(A()A )=det(A)=1det()det(A)=1 *=det(). det(A(-)A^\dagger) = \underset{=1}{\underbrace{det(A)}} det(-) \underset{= 1}{\underbrace{det(A)}}^\ast = det(-) \,.

Hence it remains to show that this is surjective, and to define this action also for 𝕂\mathbb{K} being the octonions, where general matrix calculus does not apply, due to non-associativity.


(Clifford algebra via normed division algebra)

Let 𝕂\mathbb{K} be one of the four real normed division algebras from prop. , hence one of the four real alternative division algebras from prop. .

Define a real linear map

Γ: dim (𝕂)+1,1End (𝕂 4) \Gamma \;\colon\; \mathbb{R}^{dim_{\mathbb{R}}(\mathbb{K})+1,1} \simeq End_{\mathbb{R}}(\mathbb{K}^4)

from (the real vector space underlying) Minkowski spacetime to real linear maps on 𝕂 4\mathbb{K}^4

Γ(A)(ψ ϕ)(A˜ϕ Aψ). \Gamma(A) \left( \array{ \psi \\ \phi } \right) \;\coloneqq\; \left( \array{ \tilde A \phi \\ A \psi } \right) \,.

Here on the right we are using the isomorphism from prop. for identifying a spacetime vector with a 2×22 \times 2-matrix, and we are using the trace reversal (˜)\widetilde(-) from def. .


Each operation of Γ(A)\Gamma(A) in def. is clearly a linear map, even for 𝕂\mathbb{K} being the non-associative octonions. The only point to beware of is that for 𝕂\mathbb{K} the octonions, then the composition of two such linear maps is not in general given by the usual matrix product.


(real spin representations via normed division algebras)

The map Γ\Gamma in def. gives a representation of the Clifford algebra Cl( dim (𝕂+1,1))Cl(\mathbb{R}^{dim_{\mathbb{R}}}(\mathbb{K}+1,1) ) (def. ), i.e of

  1. Cl( 2,1)Cl(\mathbb{R}^{2,1}) for 𝕂=\mathbb{K} = \mathbb{R};

  2. Cl( 3,1)Cl(\mathbb{R}^{3,1}) for 𝕂=\mathbb{K} = \mathbb{C};

  3. Cl( 5,1)Cl(\mathbb{R}^{5,1}) for 𝕂=\mathbb{K} = \mathbb{H};

  4. Cl( 9,1)Cl(\mathbb{R}^{9,1}) for 𝕂=𝕆\mathbb{K} = \mathbb{O}.

Hence this Clifford representation induces representations of the spin group Spin(dim (𝕂)+1,1)Spin(dim_{\mathbb{R}}(\mathbb{K})+1,1) on the real vector spaces

S ±𝕂 2. S_{\pm } \coloneqq \mathbb{K}^2 \,.

(Baez-Huerta 09, p. 6)


We need to check that the Clifford relation

(Γ(A)) 2=η(A,A)1 (\Gamma(A))^2 = -\eta(A,A)1

is satisfied. Now by definition, for any (ϕ,ψ)𝕂 4(\phi,\psi) \in \mathbb{K}^4 then

(Γ(A)) 2(ϕ ψ)=(A˜(Aϕ) A(A˜ψ)), (\Gamma(A))^2 \left( \array{ \phi \\ \psi } \right) \;=\; \left( \array{ \tilde A(A \phi) \\ A(\tilde A \psi) } \right) \,,

where on the right we have in each component ordinary matrix product expressions.

Now observe that both expressions on the right are sums of triple products that involve either one real factor or two factors that are conjugate to each other:

A(A˜ψ) =(x 0+x 1 y y * x 0x 1)((x 0+x 1)ϕ 1+yϕ 2 y *ϕ 1(x 0+x 1)ϕ 2) =((x 0 2+x 1 2)ϕ 1+(x 0+x 1)(yϕ 2)+y(y *ϕ 1)y((x 0+x 1)ϕ 2) ). \begin{aligned} A (\tilde A \psi) & = \left( \array{ x_0 + x_1 & y \\ y^\ast & x_0 - x_1 } \right) \cdot \left( \array{ (-x_0 + x_1) \phi_1 + y \phi_2 \\ y^\ast \phi_1 - (x_0 + x_1)\phi_2 } \right) \\ & = \left( \array{ (-x_0^2 + x_1^2) \phi_1 + (x_0 + x_1)(y \phi_2) + y (y^\ast \phi_1) - y( (x_0 + x_1) \phi_2 ) \\ \cdots } \right) \end{aligned} \,.

Since the associators of triple products that involve a real factor and those involving both an element and its conjugate vanish by prop. (hence ultimately by Artin’s theorem, prop. ). In conclusion all associators involved vanish, so that we may rebracket to obtain

(Γ(A)) 2(ϕ ψ)=((A˜A)ϕ (AA˜)ψ). (\Gamma(A))^2 \left( \array{ \phi \\ \psi } \right) \;=\; \left( \array{ (\tilde A A) \phi \\ (A \tilde A) \psi } \right) \,.

This implies the statement via the equality AA˜=A˜A=det(A)A \tilde A = \tilde A A = -det(A) (prop. ).


(index notation for generalized Pauli matrices)

Prop. says that the isomorphism of prop. is that given by forming generalized Pauli matrices. In standard physics notation these matrices are written as

Γ(x a)=(γ αα˙ a). \Gamma(x^a) = (\gamma^a_{\alpha \dot \alpha}) \,.

The spin representations given via prop. by the Clifford representation of def. are the following:

  1. for 𝕂=\mathbb{K} = \mathbb{R} the Majorana representation of Spin(2,1)Spin(2,1) on S +S S_+ \simeq S_-;

  2. for 𝕂=\mathbb{K} = \mathbb{C} the Majorana representation of Spin(3,1)Spin(3,1) on S +S S_+ \simeq S_-;

  3. for 𝕂=\mathbb{K} = \mathbb{H} the Weyl representation of Spin(5,1)Spin(5,1) on S +S_+ and on S S_-;

  4. for 𝕂=𝕆\mathbb{K} = \mathbb{O} the Majorana-Weyl representation of Spin(9,1)Spin(9,1) on S +S_+ and on S S_-.


(spinor bilinear pairings)

Under the identification of prop. the bilinear pairings

()¯():S +S \overline{(-)}(-) \;\colon\; S_+ \otimes S_-\longrightarrow \mathbb{R}


()¯Γ():S ±S ±V \overline{(-)}\Gamma (-) \;\colon\; S_\pm \otimes S_{\pm}\longrightarrow V

from above are given, respectively, by forming the real part of the canonical 𝕂\mathbb{K}-inner product

()¯():S +S \overline{(-)}(-) \colon S_+\otimes S_- \longrightarrow \mathbb{R}
(ψ,ϕ)ψ¯ϕRe(ψ ϕ) (\psi,\phi)\mapsto \overline{\psi} \phi \coloneqq Re(\psi^\dagger \cdot \phi)

and by forming the product of a column vector with a row vector to produce a matrix, possibly up to trace reversal (def. ):

S +S +V S_+ \otimes S_+ \longrightarrow V
(ψ,ϕ)ψ¯Γϕψϕ +ϕψ ˜ (\psi , \phi) \mapsto \overline{\psi}\Gamma \phi \coloneqq \widetilde{\psi \phi^\dagger + \phi \psi^\dagger}


S S V S_- \otimes S_- \longrightarrow V
(ψ,ϕ)ψϕ +ϕψ (\psi , \phi) \mapsto {\psi \phi^\dagger + \phi \psi^\dagger}

For AVA \in V the AA-component of this map is

η(ψ¯Γϕ,A)=Re(ψ (Aϕ)). \eta(\overline{\psi}\Gamma \phi, A) = Re (\psi^\dagger (A\phi)) \,.

(Baez-Huerta 09, prop. 8, prop. 9).


(real spin representation in d=2+1d = 2+1)

Consider the case 𝕂=\mathbb{K} = \mathbb{R} of real numbers.

Now V=Mat 2×2 symm()V= Mat^{symm}_{2\times 2}(\mathbb{R}) is the space of symmetric 2x2-matrices with real numbers.

V={(t+x y y tx)|t,x,y} V = \left\{ \left. \left( \array{ t + x & y \\ y & t - x } \right) \right\vert t,x,y\in \mathbb{R} \right\}

The “light-cone”-basis for this space would be

{v +(1 0 0 0),v (0 0 0 1),v y(0 1 1 0)} \left\{ v^+ \coloneqq \left( \array{ 1 & 0 \\ 0 & 0 } \right) \,, \; v^- \coloneqq \left( \array{ 0 & 0 \\ 0 & 1 } \right) \,, \; v^y \coloneqq \left( \array{ 0 & 1 \\ 1 & 0 } \right) \right\}

Its trace reversal (def. ) is

{v˜ +(0 0 0 1),v˜ (1 0 0 0),v˜ y(0 1 1 0)} \left\{ \tilde{v}^+ \coloneqq \left( \array{ 0 & 0 \\ 0 & -1 } \right) \,, \; \tilde v^- \coloneqq \left( \array{ -1 & 0 \\ 0 & 0 } \right) \,, \; \tilde v^y \coloneqq \left( \array{ 0 & 1 \\ 1 & 0 } \right) \right\}

Hence the Minkowski metric of prop. in this basis has the components

η=(0 1 0 1 0 0 0 0 2). \eta = \left( \array{ 0 & -1 & 0 \\ -1 & 0 & 0 \\ 0 & 0 & 2 } \right) \,.

As vector spaces S ±= 2S_{\pm} = \mathbb{R}^2.

The bilinear spinor pairing ()¯():S +S \overline{(-)}(-) \colon S_+ \otimes S_- \to \mathbb{R} is given by

ψ¯ϕ =ψ tϕ =ψ 1ϕ 1+ψ 2ϕ 2. \begin{aligned} \overline{\psi}\phi &= \psi^t \cdot \phi \\ & = \psi_1 \phi_1 + \psi_2 \phi_2 \end{aligned} \,.

The spinor pairing S +S +V *S_+ \otimes S_+ \otimes V^\ast \to \mathbb{R} from prop. is given on an A=A +v ++A v +A yv yA = A_+ v^+ + A_- v^- + A_y v^y (A +,A ,A yA_+, A_-, A_y \in \mathbb{R}) by the components

η(ψ¯Γϕ,A) =ψ tAϕ =ψ 1ϕ 1A ++ψ 2ϕ 2A +(ψ 1ϕ 2+ψ 2ϕ 1)A y \begin{aligned} \eta(\overline{\psi}\Gamma\phi,A) &= \psi^t \cdot A \cdot \phi \\ & = \psi_1 \phi_1 A_+ + \psi_2 \phi_2 A_- + (\psi_1 \phi_2 + \psi_2 \phi_1) A_y \end{aligned}

and S S V *S_- \otimes S_- \otimes V^\ast \to \mathbb{R} is given by the components

η(ψ¯Γϕ,A) =ψ tA˜ϕ =ψ 1ϕ 1A +ψ 2ϕ 2A +(ψ 1ϕ 2+ψ 2ϕ 1)A y. \begin{aligned} \eta(\overline{\psi}\Gamma\phi,A) &= \psi^t \cdot \tilde A \cdot \phi \\ &= -\psi_1 \phi_1 A_+ - \psi_2 \phi_2 A_- + (\psi_1 \phi_2 + \psi_2 \phi_1) A_y \end{aligned} \,.

and so, in view of the above metric components, in terms of dual bases {ψ i}\{\psi^i\} this is

μ=ψ 1ψ 1A ψ 2ψ 2A ++12(ψ 1ψ 2ψ 2ψ 1)A y \mu = - \psi^1 \otimes \psi^1 \otimes A_- - \psi^2 \otimes \psi^2 \otimes A_+ + \frac{1}{2} (\psi^1 \otimes \psi^2 \oplus \psi^2 \otimes\psi^1) \otimes A_y

So there is in particular the 2-dimensional space of isomorphisms of super Minkowski spacetime super translation Lie algebras

2,1|2 2,1|2¯ \mathbb{R}^{2,1|\mathbf{2}} \stackrel{\simeq}{\longrightarrow} \mathbb{R}^{2,1|\bar\mathbf{2}}

(not though of the corresponding super Poincaré Lie algebras, because for them the difference in the Spin-representation does matter) spanned by

(θ 1,θ 2,e)(θ 1,θ 2,e) (\theta_1,\theta_2, \vec e) \mapsto (\theta_1, -\theta_2, -\vec e)

and by

(θ 1,θ 2,e)(θ 1,θ 2,e). (\theta_1,\theta_2, \vec e) \mapsto (-\theta_1, \theta_2, -\vec e) \,.

Hence there is a 1-dimensional space of non-trivial automorphism

2,1|2 2,1|2 \mathbb{R}^{2,1|\mathbf{2}} \stackrel{\simeq}{\longrightarrow} \mathbb{R}^{2,1|\mathbf{2}}

spanned by

(θ 1,θ 2,e)(θ 1,θ 2,e). (\theta_1,\theta_2, \vec e) \mapsto (-\theta_1, -\theta_2, \vec e) \,.

Real spinors in dimensions 4,5,7 and 11


Write VMat 2×2 hermitian(𝕂)V \coloneqq Mat^{hermitian}_{2\times 2}(\mathbb{K}) \oplus \mathbb{R}.

Write S𝕂 4S \coloneqq \mathbb{K}^4. Define a real linear map

Γ:VEnd(S) \Gamma \;\colon\; V\longrightarrow End(S)

given by left matrix multiplication

Γ(A,a)(a1 2×2 A˜ A a1 2×2). \Gamma(A,a) \coloneqq \left( \array{ a \cdot 1_{2\times 2} & \tilde A \\ A & -a \cdot 1_{2\times 2} } \right) \,.

The real vector space VV in def. equipped with the inner product η(,)\eta(-,-) given by

η((A,a),(A,a))=det(A)+a 2 \eta((A,a), (A,a)) = det(A) + a^2

is Minkowski spacetime

  1. 3,1\mathbb{R}^{3,1} for 𝕂=\mathbb{K} = \mathbb{R};

  2. 4,1\mathbb{R}^{4,1} for 𝕂=\mathbb{K} = \mathbb{C};

  3. 6,1\mathbb{R}^{6,1} for 𝕂=\mathbb{K} = \mathbb{H};

  4. 10,1\mathbb{R}^{10,1} for 𝕂=𝕆\mathbb{K} = \mathbb{O}.


The map Γ\Gamma in def. gives a representation of the Clifford algebra of

  1. 3,1\mathbb{R}^{3,1} for 𝕂=\mathbb{K} = \mathbb{R};

  2. 4,1\mathbb{R}^{4,1} for 𝕂=\mathbb{K} = \mathbb{C};

  3. 6,1\mathbb{R}^{6,1} for 𝕂=\mathbb{K} = \mathbb{H};

  4. 10,1\mathbb{R}^{10,1} for 𝕂=𝕆\mathbb{K} = \mathbb{O}.

Under restriction along Spin(n+2,1)Cl(n+2,1)Spin(n+2,1) \hookrightarrow Cl(n+2,1) this is isomorphic to

  1. for 𝕂=\mathbb{K} = \mathbb{R} the Majorana representation of Spin(3,1)Spin(3,1) on SS;

  2. for 𝕂=\mathbb{K} = \mathbb{C} the Dirac representation of Spin(4,1)Spin(4,1) on SS;

  3. for 𝕂=\mathbb{K} = \mathbb{H} the Dirac representation of Spin(6,1)Spin(6,1) on SS;

  4. for 𝕂=𝕆\mathbb{K} = \mathbb{O} the Majorana representation of Spin(10,1)Spin(10,1) on SS.

(Baez-Huerta 10, p. 10, prop. 8, prop. 9)


Γ 0(0 1 2x2 1 2×2 0). \Gamma^0 \coloneqq \left( \array{ 0 & - 1_{2x2} \\ 1_{2\times 2} & 0 } \right) \,.

Under the identification of prop. of the bilinear pairings

()¯():SS \overline{(-)}(-) \;\colon\; S \otimes S \longrightarrow \mathbb{R}


()¯Γ():SSV \overline{(-)}\Gamma (-) \;\colon\; S \otimes S \longrightarrow V

of remark , the first is given by

(ψ,ϕ)ψ¯ϕRe(ψ Γ 0ϕ) (\psi,\phi) \mapsto \overline\psi\phi \coloneqq Re(\psi^\dagger \Gamma^0 \phi)

and the second is characterized by

η(ψ¯Γϕ,A) =ψ¯(Γ(A)ϕ) =Re(ψ Γ 0Γ(A)ϕ). \begin{aligned} \eta \left( \overline{\psi}\Gamma\phi, A \right) &= \overline{\psi}(\Gamma(A)\phi) \\ & = Re( \psi^\dagger \Gamma^0 \Gamma(A)\phi) \end{aligned} \,.

(Baez-Huerta 10, prop. 10, prop. 11).

Real pinor representations (including spacetime reflection)

We discuss Pin(10,1)-representations when using the spinor conventions from CDF, II.7.1, as in Prop. above.

The statement is Prop. below. We need the following facts from the above discussion:

Dirac representation. From Prop. 2.18 we have

(1)Γ 0 =+Γ 0AAAAΓ a =Γ aAAAforaspatial \Gamma_0^\dagger = + \Gamma_0 \phantom{AAAA} \Gamma_{a}^\dagger = - \Gamma_{a} \phantom{AAA} \text{for} \; a \; \text{spatial}

Charge conjugation matrix. From Prop. 2.22 and Remark 2.23 we get for d=11d = 11 (see the table there) that the charge conjugation matrix satisfies

(2)Γ a TC=CΓ a \Gamma_a^T C = - C \Gamma_a

Majorana condition. From Prop. 2.29 we have that the Majorana condition on ψ\psi is equivalent to

(3)ψ Γ 0=ψ tC \psi^\dagger \Gamma_0 = \psi^t C

(Pin group-representation on Majorana spinors)

For p+1=10+1p+1 = 10+1 and with choice of Dirac representation from Prop. , multiplication by Γ a\Gamma_{a} does not preserves the Majorana spinor condition (3). But multiplication by ±iΓ a\pm i \Gamma_{a} does.

Hence to get on Majorana spinors not just a Spin group-representation, but even a Pin group-representation (incluing spacetime reflections on top of rotations) in the spinor convention of Prop. , one needs to use the Cliffor algebra generated from {iΓ a} a0 p\{i \Gamma_a\}_{a - 0}^p


First assume that the index is spatial. Suppose ψ\psi is Majorana, then we compute as follows:

(Γ aψ) TC =ψ TΓ a TC =ψ TCΓ a =ψ Γ 0Γ a =+ψ Γ aΓ 0 =ψ Γ a Γ 0 =(Γ aψ) Γ 0 \begin{aligned} (\Gamma_a \psi)^T C & = \psi^T \Gamma_a^T C \\ & = - \psi^T C \Gamma_a \\ & = - \psi^\dagger \Gamma_0 \Gamma_a \\ & = + \psi^\dagger \Gamma_a \Gamma_0 \\ & = - \psi^\dagger \Gamma_a^\dagger \Gamma_0 \\ & = - (\Gamma_a \psi)^\dagger \Gamma_0 \end{aligned}


  • the first equality is that transposition is an algbra anti-homomorphism,

  • the second equality is the charge conjugation relation (2),

  • the third equality is the Majorana condition (3) on ψ\psi,

  • the fourth equality is the Clifford relation {Γ 0,Γ a}=0\{\Gamma_0,\Gamma_{a}\} = 0,

  • the fifth equality is the anti-hermiticity (1),

  • the sixth equality is that \dagger is algebra anti-homomorphism.

In conclusion, the overall minus sign between the first and the last term means that Γ aψ\Gamma_{a} \psi fails the Majorana condition (3).

But with iΓ ai \Gamma_{a} instead of Γ a\Gamma_{a}, the same computation applies, except that there is one extra sign when \dagger is applied, from i =i *=ii^\dagger = i^\ast = -i. Hence this fixes the sign.

Finally, in the case that a=0a = 0 the same computation goes through once more, except for two extra signs: One from the difference of sign under \dagger from (1), the other due to the difference of sign in commuting through Γ 0\Gamma_0. Hence the conclusion remains the same.

Spacetime supersymmetry

We have seen in example that super-extensions of the symmetries of Minkowski spacetime are given by real spin representations, and then we constructed and classified these (above).

Hence every real spin representation of Spin(d1,1)Spin(d-1,1) induces a super Lie algebra extension of the Poincaré Lie algebra ℑ𝔰𝔬( d1,1)\mathfrak{Iso}(\mathbb{R}^{d-1,1}) in that dimension, i.e. of the Lie algebra of the isometry group of the Minkowski spacetime (def. ) in that dimension. These are the supersymmetry algebras in physics.

Since we may recover a Minkowski spacetime from its Poincaré Lie algebra as the (vector space underlying the) coset of the Poincaré Lie algebra by the Lie algebra 𝔰𝔬(d1,1)\mathfrak{so}(d-1,1) of the spin group (the orthogonal Lie algebra in Lorentzian signature)

d1,1ℑ𝔰𝔬( d1,1)/𝔰𝔬(d1,1) \mathbb{R}^{d-1,1} \simeq \mathfrak{Iso}(\mathbb{R}^{d-1,1})/\mathfrak{so}(d-1,1)

(namely as the Lie algebra of translations along itself), every super Lie algebra extension of the Poincaré Lie algebra defines a super Lie algebra extension of Minkowski spacetime. These extensions are the super Minkowski spacetimes d1,1|N\mathbb{R}^{d-1,1\vert N} which in the physics literature are often just called “superspace”.

To set the scene, we recall some basics of ordinary spacetime symmetry in

Then in

we specialize to the particular such extensions commonly known as supersymmetries.

Finally we discuss the question of how god-given this common choice is, in

Spacetime symmetry


For dd \in \mathbb{N}, write d1,1\mathbb{R}^{d-1,1} for Minkowski spacetime (def. ), regarded as the inner product space whose underlying vector space is d\mathbb{R}^d and equipped with the bilinear form given in the canonical linear basis of d\mathbb{R}^d by

ηdiag(1,+1,+1,,+1). \eta \coloneqq diag(-1,+1,+1, \cdots, +1) \,.

The Poincaré group Iso( d1,1)Iso(\mathbb{R}^{d-1,1}) is the isometry group of this inner product space. The Poincaré Lie algebra 𝔦𝔰𝔬( d1,1)\mathfrak{iso}(\mathbb{R}^{d-1,1}) is the Lie algebra of this Lie group (its Lie differentiation)

𝔦𝔰𝔬( d1,1)Lie(Iso( d1,1)). \mathfrak{iso}(\mathbb{R}^{d-1,1}) \coloneqq Lie(Iso(\mathbb{R}^{d-1,1})) \,.

The Poincaré group is the semidirect product group

Iso( d1,1) d1,1O(d1,1) Iso(\mathbb{R}^{d-1,1}) \simeq \mathbb{R}^{d-1,1} \rtimes O(d-1,1)

of the Lorentz group O(d1,1)O(d-1,1) (the group of linear isometries of Minkowski spacetime) with the d\mathbb{R}^d regarded as the translation group along itself, via the defining action.

Accordingly, the Poincaré Lie algebra is the semidirect product Lie algebra

𝔦𝔰𝔬( d1,1) d1,1𝔰𝔬 +(d1,1) \mathfrak{iso}(\mathbb{R}^{d-1,1}) \simeq \mathbb{R}^{d-1,1} \rtimes \mathfrak{so}^+(d-1,1)

of the abelian Lie algebra on d\mathbb{R}^d with the (orthochronous) special orthogonal Lie algebra 𝔰𝔬(d1,1)\mathfrak{so}(d-1,1).


For {P a}\{P_a\} the canonical linear basis of d\mathbb{R}^d, and for {L ab=L ba}\{L_{a b} = - L_{b a}\} the corresponding canonical basis of 𝔰𝔬(d1,1)\mathfrak{so}(d-1,1), then the Lie bracket in 𝔦𝔰𝔬( d1,1)\mathfrak{iso}(\mathbb{R}^{d-1,1}) is given as follows:

[P a,P b] =0 [L ab,L cd] =η daL bcη bcL ad+η acL bdη dbL ac [L ab,P c] =η acP bη bcP a \begin{aligned} [P_a, P_b] & = 0 \\ [L_{a b}, L_{c d}] & = \eta_{d a} L_{b c} -\eta_{b c} L_{a d} +\eta_{a c} L_{b d} -\eta_{d b} L_{a c} \\ [L_{a b}, P_c] & = \eta_{a c} P_b -\eta_{bc} P_a \end{aligned}

Since Lie differentiation sees only the connected component of a Lie group, and does not distinguish betwee a Lie group and any of its discrete covering spaces, we may equivalently consider the Lie algebra of the spin group Spin(d1,1)SO +(d1,1)Spin(d-1,1) \to SO^+(d-1,1) (the double cover of the proper orthochronous Lorentz group) and its action on d1,1\mathbb{R}^{d-1,1}.

By the discussion at spin group, the Lie algebra of Spin(d1,1)Spin(d-1,1) is the Lie algebra spanned by the Clifford algebra bivectors

L abΓ aΓ b L_{a b} \leftrightarrow \Gamma_a \Gamma_b

and its action on itself as well as on the vectors, identified with single Clifford generators

P aΓ a P_a \leftrightarrow \Gamma_a

is given by forming commutators in the Clifford algebra:

[L ab,P c]12[Γ ab,Γ c] [L_{a b}, P_c] \leftrightarrow \tfrac{1}{2}[\Gamma_{a b}, \Gamma_c ]
[L ab,L cd]12[Γ ab,Γ cd]. [L_{a b}, L_{c d}] \leftrightarrow \tfrac{1}{2}[\Gamma_{a b}, \Gamma_{c d} ] \,.

Via the Clifford relation

Γ aΓ b+Γ bΓ a=2η ab \Gamma_a \Gamma_b + \Gamma_b \Gamma_a = -2 \eta_{a b}

this yields the claim.


Dually, the Chevalley-Eilenberg algebra CE(𝔦𝔰𝔬( d1)CE(\mathfrak{iso}(\mathbb{R}^{d-1}) is generated from d,1\mathbb{R}^{d,1} and 2 d,1\wedge^2 \mathbb{R}^{d,1}. For {t a}\{t_a\} the standard basis of d1,1\mathbb{R}^{d-1,1} we write {ω ab}\{\omega^{a b}\} and {e a}\{e^a\} for these generators. With (η ab)(\eta_{a b}) the components of the Minkowski metric we write

ω a bω acη cb. \omega^{a}{}_b \coloneqq \omega^{a c}\eta_{c b} \,.

In terms of this the CE-differential that defines the Lie algebra structure is

d CE:ω ab=ω a cω cb d_{CE} \colon \omega^{a b} = \omega^a{}_c \wedge \omega^{c b}
d CE:e aω a bt b d_{CE} \colon e^a \mapsto \omega^{a}{}_b \wedge t^b

Super Poincaré and super Minkowski symmetry

We may now finally make explicit the super-extension of spacetime symmetry according to example :

In all of the following it is most convenient to regard super Lie algebras dually via their Chevalley-Eilenberg algebras:


For 𝔤\mathfrak{g} a super Lie algebra of finite dimension, then its Chevalley-Eilenberg algebra CE(𝔤)CE(\mathfrak{g}) is the super-Grassmann algebra on the dual super vector space

𝔤 * \wedge^\bullet \mathfrak{g}^\ast

equipped with a differential d 𝔤d_{\mathfrak{g}} that on generators is the linear dual of the super Lie bracket

d 𝔤[,] *:𝔤 *𝔤 *𝔤 * d_{\mathfrak{g}} \;\coloneqq\; [-,-]^\ast \;\colon\; \mathfrak{g}^\ast \to \mathfrak{g}^\ast \wedge \mathfrak{g}^\ast

and which is extended to 𝔤 *\wedge^\bullet \mathfrak{g}^\ast by the graded Leibniz rule (i.e. as a graded derivation).


Here all elements are (×/2)(\mathbb{Z} \times \mathbb{Z}/2)-bigraded, the first being the cohomological grading nn in n𝔤 *\wedge^\n \mathfrak{g}^\ast, the second being the super-grading σ\sigma (even/odd).

For α iCE(𝔤)\alpha_i \in CE(\mathfrak{g}) two elements of homogeneous bi-degree (n i,σ i)(n_i, \sigma_i), respectively, the sign rule is

α 1α 2=(1) n 1n 2(1) σ 1σ 2α 2α 1. \alpha_1 \wedge \alpha_2 = (-1)^{n_1 n_2} (-1)^{\sigma_1 \sigma_2}\; \alpha_2 \wedge \alpha_1 \,.

(See at signs in supergeometry for discussion of this sign rule and of an alternative sign rule that is also in use. )

We may think of CE(𝔤)CE(\mathfrak{g}) equivalently as the dg-algebra of left-invariant super differential forms on the corresponding simply connected super Lie group .

The concept of Chevalley-Eilenberg algebras is traditionally introduced as a means to define Lie algebra cohomology:


Given a super Lie algebra 𝔤\mathfrak{g}, then

  1. an nn-cocycle on 𝔤\mathfrak{g} (with coefficients in \mathbb{R}) is an element of degree (n,even)(n,even) in its Chevalley-Eilenberg algebra CE(𝔤)CE(\mathfrak{g}) (def. ) which is d 𝕘d_{\mathbb{g}} closed.

  2. the cocycle is non-trivial if it is not d 𝔤d_{\mathfrak{g}}-exact

  3. hene the super-Lie algebra cohomology of 𝔤\mathfrak{g} (with coefficients in \mathbb{R}) is the cochain cohomology of its Chevalley-Eilenberg algebra

    H (𝔤,)=H (CE(𝔤)). H^\bullet(\mathfrak{g}, \mathbb{R}) = H^\bullet(CE(\mathfrak{g})) \,.

The following says that the Chevalley-Eilenberg algebra is an equivalent incarnation of the super Lie algebra:


The functor

CE:sLieAlg findgAlg op CE \;\colon\; sLieAlg^{fin} \hookrightarrow dgAlg^{op}

that sends a finite dimensional super Lie algebra 𝔤\mathfrak{g} to its Chevalley-Eilenberg algebra CE(𝔤)CE(\mathfrak{g}) (def. ) is a fully faithful functor which hence exibits super Lie algebras as a full subcategory of the opposite category of differential-graded algebras.



d d \in \mathbb{N}

be a spacetime dimension and let

NRep (Spin(d1,1)) N \in Rep_{\mathbb{R}}(Spin(d-1,1))

be a real spin representation of the spin group cover Spin(d1,1)Spin(d-1,1) of the Lorentz group O(d1,1)O(d-1,1) in this dimension. Then the dd-dimensional NN-supersymmetric super-Minkowski spacetime d1,1|N\mathbb{R}^{d-1,1|N} is the super Lie algebra that is characterized by the fact that its Chevalley-Eilenberg algebra CE( d1,1)CE(\mathbb{R}^{d-1,1}) is as follows:

The algebra has generators (as an associative algebra over \mathbb{R})

e adeg=(1,even)andψ αdeg=(1,odd) \underset{deg = (1,even)}{\underbrace{e^a}} \;\;\;\; \text{and} \;\;\;\; \underset{deg = (1,odd)}{\underbrace{\psi^\alpha}}

for a{0,1,2,,9}a \in \{0,1,2, \cdots, 9\} and α{1,2,dim (N)}\alpha \in \{1, 2, \cdots dim_{\mathbb{R}}(N)\} subjects to the relations

e ae b=e be a ψ αψ β=+ψ βψ α e aψ α=ψ αe a \begin{aligned} e^a \wedge e^b = - e^b \wedge e^a \\ \psi^\alpha \wedge \psi^\beta = + \psi^\beta \wedge \psi^\alpha \\ e^a \wedge \psi^\alpha = - \psi^\alpha \wedge e^a \end{aligned}

(see at signs in supergeometry), and the differential d CEd_{CE} acts on the generators as follows:

d d1,1|Nψ α 0 d d1,1|Ne a ψ¯Γ aψ (C ααΓ a α β)ψ αψ β, \begin{aligned} d_{\mathbb{R}^{d-1,1\vert N}} \; \psi^\alpha & \coloneqq 0 \\ d_{\mathbb{R}^{d-1,1\vert N}} \; e^a & \coloneqq \overline{\psi} \wedge \Gamma^a \psi \\ & \coloneqq \left(C_{\alpha \alpha'} {\Gamma^a}^{\alpha'}{}_\beta\right) \psi^\alpha \wedge \psi^\beta \end{aligned} \,,


  1. ψ¯Γ aψ\overline{\psi} \wedge \Gamma^a \psi denotes the aa-component of the Spin(d1,1)Spin(d-1,1)-invariant spinor bilinear pairing NotimeN dN \otime N \to \mathbb{R}^d that comes with every real spin representation applied to ψψ\psi \wedge \psi regarded as an NNN \otimes N-valued form;

  2. hence in components (if NN is a Majorana spinor representation, by prop. :

    1. C=(C αα)C = (C_{\alpha \alpha'}) is the charge conjugation matrix (as discussed at Majorana spinor);

    2. Γ a=((Γ a) α β)\Gamma^a = ((\Gamma^a)^{\alpha}{}_\beta) are the matrices representing the Clifford algebra action on NN in the linear basis {ψ α} α=1 dim (N)\{\psi^\alpha\}_{\alpha = 1}^{dim_{\mathbb{R}}(N)}

  3. summation over paired indices is understood.

That this indeed yields a super Lie algebra follows by the symmetry and equivariance of the bilinear spinor pairing (via prop. .

There is a canonical Lie algebra action of the special orthogonal Lie algebra

Lie(Spin(d1,1))𝔰𝔬(d1,1) Lie(Spin(d-1,1)) \simeq \mathfrak{so}(d-1,1)

on d1,1|1\mathbb{R}^{d-1,1\vert 1}. The NN-supersymmetric super Poincaré Lie algebra 𝔦𝔰𝔬( d1,1|N)\mathfrak{iso}(\mathbb{R}^{d-1,1\vert N}) in dimension dd is the super Lie algebra which is the semidirect product Lie algebra of this Lie algebra action

𝔦𝔰𝔬( d1,1|N)= d1,1|N𝔰𝔬(d1,1). \mathfrak{iso}(\mathbb{R}^{d-1,1\vert N}) = \mathbb{R}^{d-1,1\vert N} \rtimes \mathfrak{so}(d-1,1) \,.

This is characterized by the fact that its Chevalley-Eilenberg algebra CE(𝔦𝔰𝔬( d1,1|N))CE(\mathfrak{iso}(\mathbb{R}^{d-1,1\vert N})) is as follows:

it is generated from elements

e adeg=(1,even)andψ αdeg=(1,odd)andω ab=ω badeg=(1,even) \underset{deg = (1,even)}{\underbrace{e^a}} \;\;\;\; and \;\;\;\; \underset{deg = (1,odd)}{\underbrace{\psi^\alpha}} \;\;\;\; and \;\;\;\; \underset{deg = (1,even)}{\underbrace{\omega^{a b} = - \omega^{b a}}}

with the super vielbein (e a,ψ α)(e^a, \psi^\alpha) as before, and with ω ab\omega^{a b} the dual basis of the induced linear basis for the vector space of skew-symmetric matrices underlying the special orthogonal Lie algebra. The commutation relations are as before, together with the relation that the generators ω ab\omega^{a b} anti-commute with every generator. Finally the differential d 𝔦𝔰𝔬( d1,1|N)d_{\mathfrak{iso}(\mathbb{R}^{d-1,1\vert N})} acts on these generators as follows:

d 𝔦𝔰𝔬( d1,1|N)ψ α (14ω abΓ abψ) α (14(Γ ab) α β)ω abψ β d 𝔦𝔰𝔬( d1,1|N)e a ψ¯Γ aψω a be b (C ααΓ a α β)ψ αψ βω a be b , \begin{aligned} d_{\mathfrak{iso}(\mathbb{R}^{d-1,1\vert N})} \; \psi^\alpha & \coloneqq \left(\tfrac{1}{4}\omega^{a b} \Gamma_{a b} \psi \right)^\alpha \\ & \coloneqq \left(\tfrac{1}{4} (\Gamma_{a b})^\alpha{}_{\beta} \right) \omega^{a b} \wedge \psi^\beta \\ d_{\mathfrak{iso}(\mathbb{R}^{d-1,1\vert N})} \; e^a & \coloneqq \overline{\psi} \wedge \Gamma^a \psi - \omega^a{}_b \wedge e^b \\ & \coloneqq \left( C_{\alpha \alpha'} {\Gamma^a}^{\alpha'}{}_\beta \right) \psi^\alpha \wedge \psi^\beta - \omega^a{}_b \wedge e^b \\ \end{aligned} \,,

where we are shifting spacetime indicices with the Lorentz metric

(η ab)diag(1,1,1,,1). (\eta_{a b}) \coloneqq diag(-1,1,1,\cdots, 1) \,.

The canonical maps between these super Lie algebras, dually between their Chevalley-Eilenberg algebras, that send each generator to itself, if present, or to zero if not, constitute the diagram

d1,1|N 𝔦𝔰𝔬( d1,1|N) 𝔰𝔬(d1,1). \array{ \mathbb{R}^{d-1,1\vert N} &\hookrightarrow& \mathfrak{iso}(\mathbb{R}^{d-1,1\vert N}) \\ && \downarrow \\ && \mathfrak{so}(d-1,1) } \,.

Poincaré connections: Graviton and gravitino field

We may now apply the general discussion of super Lie algebra valued super differential forms, def. , to the case of the super Poincare Lie algebra, def. .

its Chevalley-Eilenberg algebra CE(ℑ𝔰𝔬( d1,1|N))CE(\mathfrak{Iso}(\mathbb{R}^{d-1,1|N})) is generated on

  • elements {e a}\{e^a\} and {ω ab}\{\omega^{ a b}\} of degree (1,even)(1,even)

  • and elements {ψ α}\{\psi^\alpha\} of degree (1,odd)(1,odd)

with the differential defined by

d CEω ab=ω a bω bc d_{CE} \, \omega^{a b} = \omega^a{}_b \wedge \omega^{b c}
d CEe a=ω a be b+i2ψ¯Γ aψ d_{CE} \, e^{a } = \omega^a{}_b \wedge e^b + \frac{i}{2}\overline{\psi} \Gamma^a \psi
d CEψ=14ω abΓ abψ. d_{CE} \, \psi = \frac{1}{4} \omega^{ a b} \Gamma_{a b} \psi \,.

Accordingly its Weil algebra W(ℑ𝔰𝔬( d1,1|N))W(\mathfrak{Iso}(\mathbb{R}^{d-1,1|N})) has these generators together with a further degree-shifted copy of each {t a}\{t^a\}, {r ab}\{r^{a b}\} and {ρ α}\{\rho^{\alpha}\} with differential given by

d Wω ab=ω a bω bc+r ab d_{W} \, \omega^{a b} = \omega^a{}_b \wedge \omega^{b c} + r^{a b}
d We a=ω a be b+i2ψ¯Γ aψ+t a d_{W} \, e^{a } = \omega^a{}_b \wedge e^b + \frac{i}{2} \overline{\psi} \Gamma^a \psi + t^a
d Wψ=14ω abΓ abψ+ρ. d_{W} \, \psi = \frac{1}{4} \omega^{ a b} \Gamma_{a b} \psi + \rho \,.

Differential form data with values in this is a morphism of dg-algebras from the Weil algebra W(ℑ𝔰𝔬( d1,1|N))W(\mathfrak{Iso}(\mathbb{R}^{d-1,1|N})) to the deRham dg-algebra Ω ( p|q)\Omega^\bullet(\mathbb{R}^{p|q}), def.

Ω (X)W(ℑ𝔰𝔬( d1,1|N)):(A,F A). \Omega^\bullet(X) \leftarrow W(\mathfrak{Iso}(\mathbb{R}^{d-1,1|N})) : (A,F_A) \,.

This is ∞-Lie algebroid valued differential form data with ∞-Lie algebroid valued curvature that is explicitly given by:

  • connection forms / field configuration

  • curvature forms / field strengths

    • T=dE+ΩE+Γ(Ψ¯Ψ)Ω 2(X, d1,1)T = d E + \Omega \cdot E + \Gamma(\overline{\Psi} \wedge \Psi) \in \Omega^2(X,\mathbb{R}^{d-1,1}) - the torsion

    • R=dΩ+[ΩΩ]Ω 2(X,𝔰𝔬(10,1))R = d \Omega + [\Omega \wedge \Omega] \in \Omega^2(X, \mathfrak{so}(10,1)) - the Riemann curvature

    • ρ=dΨ+(ΩΨ)Ω 2(X,S)\rho = d \Psi + (\Omega \wedge \Psi) \in \Omega^2(X, S) – the covariant derivative of the gravitino


The Chevalley-Eilenberg algebra CE(𝔦𝔰𝔬( d1,1|N))CE(\mathfrak{iso}(\mathbb{R}^{d-1,1|N})) is generated by

  • elements {e a}\{e^a\} and {ω ab}\{\omega^{ a b}\} of degree (1,even)(1,even)

  • and elements {ψ α}\{\psi^\alpha\} of degree (1,odd)(1,odd)

with the differential defined by

d CEω ab=ω a bω bc d_{CE} \, \omega^{a b} = \omega^a{}_b \wedge \omega^{b c}
d CEe a=ω a be b+ψ¯Γ aψ d_{CE} \, e^{a } = \omega^a{}_b \wedge e^b + \overline{\psi} \Gamma^a \psi
d CEψ=14ω abΓ abψ. d_{CE} \, \psi = \frac{1}{4} \omega^{ a b} \Gamma_{a b} \psi \,.

Discarding the terms involving ω\omega here this is the CE algebra of the super translation algebra underlying super Minkowski spacetime.

In this way the super-Poincaré Lie algebra and its extensions is usefully discussed for instance in (D’Auria-Fré 82) and in (Azcárraga-Townsend 89, CAIB 99). In much of the literature instead the following equivalent notation is popular, which more explicitly involves the coordinates on super Minkowski space.


The abstract generators in def. are identified with left invariant 1-forms on the super-translation group (= super Minkowski space) as follows.

Let (x a,θ α)(x^a, \theta^\alpha) be the canonical coordinates on the supermanifold d|N\mathbb{R}^{d|N} underlying the super translation group. Then the identification is

  • ψ α=dθ α\psi^\alpha = d \theta^\alpha.

  • e a=dx a+θ¯Γ adθe^a = d x^a + \overline{\theta} \Gamma^a d \theta.

Notice that this then gives the above formula for the differential of the super-vielbein in def. as

de a =d(dx a+θ¯Γ adθ) =dθ¯Γ adθ =ψ¯Γ aψ. \begin{aligned} d e^a & = d (d x^a + \overline{\theta} \Gamma^a d \theta) \\ & = d \overline{\theta}\Gamma^a d \theta \\ & = \overline{\psi}\Gamma^a \psi \end{aligned} \,.

The term ψ¯Γ aψ\overline{\psi} \Gamma^a \psi is sometimes called the supertorsion of the super-vielbein ee, because the defining equation

d CEe aω a be b=ψ¯Γ aψ d_{CE} e^{a } -\omega^a{}_b \wedge e^b = \overline{\psi} \Gamma^a \psi

may be read as saying that ee is torsion-free except for that term. Notice that this term is the only one that appears when the differential is applied to “Lorentz scalars”, hence to object in CE(𝔦𝔰𝔬( d1,1|N))CE(\mathfrak{iso}(\mathbb{R}^{d-1,1\vert N})) which have “all indices contracted”. See also at torsion constraints in supergravity.

Notably we have

d(ψ¯Γ a 1a pψe a 1e a p)(ψ¯Γ a 1a pψe a 1e a p1)(Ψ¯Γ a pΨ). d \left( \overline{\psi} \wedge \Gamma^{a_1 \cdots a_p} \psi \wedge e_{a_1} \wedge \cdots \wedge e_{a_p} \right) \propto \left( \overline{\psi} \wedge \Gamma^{a_1 \cdots a_p} \psi \wedge e_{a_1} \wedge \cdots \wedge e_{a_{p-1}} \right) \wedge \left( \overline{\Psi} \wedge \Gamma_{a_p} \Psi \right) \,.

This remaining operation “eΨ 2e \mapsto \Psi^2” of the differential acting on Lorentz scalars is sometimes denoted “t 0t_0”, e.g. in (Bossard-Howe-Stelle 09, equation (8)).

This relation is what governs all of the exceptional super Lie algebra cocycles that appear below: for some combinations of (d,p,N)(d,p,N) a Fierz identity implies that the term

(ψ¯Γ a 1a pψe a 1e a p1)(Ψ¯Γ a pΨ) \left( \overline{\psi} \wedge \Gamma^{a_1 \cdots a_p} \psi \wedge e_{a_1} \wedge \cdots \wedge e_{a_{p-1}} \right) \wedge \left( \overline{\Psi} \wedge \Gamma_{a_p} \Psi \right)

vanishes identically, and hence in these dimensions the term

ψ¯Γ a 1a pψe a 1e a p \overline{\psi} \wedge \Gamma^{a_1 \cdots a_p} \psi \wedge e_{a_1} \wedge \cdots \wedge e_{a_p}

is a cocycle in super Lie algebra cohomology.

Superconformal symmetry

We discuss super Lie algebra extensions of the conformal Lie algebra of d1,1\mathbb{R}^{d-1,1} (equivalently the isometry Lie algebra of anti de Sitter space of dimension d+1d+1, see also at AdS-CFT.)


There exist superconformal extensions of the super Poincaré Lie algebra, (besides dimension 2\leq 2) in dimensions 3,4,5,6 as follows (with notation as at super Lie algebra – classification):

ddNNsuperconformal super Lie algebraR-symmetrybrane worldvolume theory
32k+12k+1B(k,2)B(k,2) \simeq osp(2k+1/4)(2k+1/4)SO(2k+1)SO(2k+1)
32k2kD(k,2)D(k,2)\simeq osp(2k/4)(2k/4)SO(2k)SO(2k)M2-brane
4k+1k+1A(3,k)𝔰𝔩(4/k+1)A(3,k)\simeq \mathfrak{sl}(4/k+1)U(k+1)U(k+1)D3-brane
6kkD(4,k)D(4,k) \simeq osp(8/2k)(8/2k)Sp(k)Sp(k)M5-brane

There exists no superconformal extension of the super Poincaré Lie algebra in dimension d>6d \gt 6.

This is due to (Shnider 88), see also (Nahm 78). Review is in (Minwalla 98, section 4.2). See also the references at super p-brane – As part of the AdS-CFT correspondence.

Proof (sketch)

By realizing the conformal real Lie algebra 𝔰𝔬( d,2)\mathfrak{so}(\mathbb{R}^{d,2}) as a real section of the complexified 𝔰𝔬( d+2)\mathfrak{so}(\mathbb{C}^{d+2}) one is reduced to finding those (finite dimensional) simple super Lie algebras over the complex numbers whose even-graded part extends 𝔰𝔬( d+2)\mathfrak{so}(\mathbb{C}^{d+2}) and such that the implied representation of that on the odd-graded part contains the spin representation.

The complex finite dimensional simple super Lie algebras have been classified, see at super Lie algebra – Classification. By the tables shown there

𝔤\mathfrak{g}𝔤 even\mathfrak{g}_{even}𝔤 even\mathfrak{g}_{even} rep on 𝔤 odd\mathfrak{g}_{odd}
B(m,n)B(m,n)B mC nB_m \oplus C_nvector \otimes vector
D(m,n)D(m,n)D mC nD_m \oplus C_nvector \otimes vector
D(2,1,α)D(2,1,\alpha)A 1A 1A 1A_1 \oplus A_1 \oplus A_1vector \otimes vector \otimes vector
F(4)F(4)B 3A 1B_3\otimes A_1spinor \otimes vector
G(3)G(3)G 2A 1G_2\oplus A_1spinor \otimes vector
Q(n)Q(n)A nA_nadjoint
𝔤\mathfrak{g}𝔤 even\mathfrak{g}_{even}𝔤 even\mathfrak{g}_{even} rep on 𝔤 1\mathfrak{g}_{{-1}}
A(m,n)A(m,n)A mA nCA_m \oplus A_n \oplus Cvector \otimes vector \otimes \mathbb{C}
A(m,m)A(m,m)A mA nA_m \oplus A_nvector \otimes vector
C(n)C(n) 1\mathbb{C}_{-1} \oplus \mathbb{C}vector \otimes \mathbb{C}

the only manifest spinor representation of 𝔰𝔬(2k+1)=B k\mathfrak{so}(2k+1) = B_k or of 𝔰𝔬(2k)=D k\mathfrak{so}(2k) = D_k appears in the exceptional super Lie algebra F(4)F(4), which contains B 3=𝔰𝔬(7)B_3 = \mathfrak{so}(7) in its even parts acting spinorially on its odd part. This hence gives a superconformal super Lie algebra in dimension 72=57-2 = 5, as shown in the proposition.

But other spinor representations may still disguise as vector representations of other Lie algebras under one of the exceptional isomorphisms. These exist only in low dimensions, and hence to conclude the proof it is sufficient to just list all candidates.

First there is the exceptional isomorphism

𝔰𝔬(5)𝔰𝔭(2)=C 2 \mathfrak{so}(5) \simeq \mathfrak{sp}(2) = C_2

with the spinor representation of 𝔰𝔬(5)\mathfrak{so}(5) being the vector representation of 𝔰𝔭(2)=C 2\mathfrak{sp}(2) = C_2. This we find in the above tables as a summand in the even-graded subalgebra of B(m,2)B(m,2) and of D(m,2)D(m,2). Hence these are superconformal super Lie algebras in dimension 52=35-2 = 3, as shown in the statement.

The other exceptional isomorphism of relevance is

𝔰𝔬(6)𝔰𝔲(4)=A 3 \mathfrak{so}(6)\simeq \mathfrak{su}(4) = A_3

with the spinor representation of 𝔰𝔬(6)\mathfrak{so}(6) being the vector representation of 𝔰𝔲(4)=A 3\mathfrak{su}(4) = A_3. By the above tables this appears as a summand in the even-graded subalgebra of the super Lie algebra A(3,k)A(3,k), and so this is the superconformal algebra in dimension 62=46-2 = 4.

Finally by triality the vector representation of 𝔰𝔬(8)=D 4\mathfrak{so}(8) = D_4 is isomorphic to its spinor representation. By the above tables this means that D(4,k)D(4,k) is a superconformal algebra in dimension 82=68-2 = 6. For details on this see (Shnider 88, last paragraphs)


Further constraints follow from requiring super-unitary representations (Minwalla 98, section 4.3). This restricts for instance the 6d superconformal algebra to D(4,1)=𝔬𝔰𝔭(8|2)D(4,1) = \mathfrak{osp}(8|2) and D(4,1)=𝔬𝔰𝔭(8,4)D(4,1) = \mathfrak{osp}(8,4), the latter being (over the reals as 𝔰𝔬(8 *|4)=𝔬𝔰𝔭(6,2|4)\mathfrak{so}(8^\ast|4) = \mathfrak{osp}(6,2|4)) the symmetry algebra of the 6d (2,0)-superconformal QFT on the worldvolume of the M5-brane.

Supersymmetry from the superpoint

Above we have discussed the definition and classification of supersymmetry in the strict sense of “spacetime supersymmetry” and specifically in the strict sense of super-extensions (def. ) of the Poincaré Lie algebra by real spin representations. However, example showed that there are other exotic super-extensions of the Poincaré Lie algebra which are not of this form.

While there are further conditions, motivated from physics, which one may impose to single out the “ordinary” super-extensions from the exotic ones (remark ) this raises the question which fundamental mathematical principle, if any, singles out the “ordinary” super-extensions.

Here we discuss one such principle. Supersymmetry and spin representations emerge from forming consecutive maximal invariant central extensions starting from the superpoint (Huerta-Schreiber 17).

Before giving the general definition and discussion, we consider the simplest case right away:


Consider the superpoint

0|1 \mathbb{R}^{0\vert 1}

regarded as an abelian super Lie algebra, via example .

Its maximal central extension is the N=1N = 1 super-worldline of the superparticle:

0,1|1 0|1. \array{ \mathbb{R}^{0,1\vert \mathbf{1}} \\ \downarrow \\ \mathbb{R}^{0\vert 1} } \,.
  • whose even part is spanned by one generator HH

  • whose odd part is spanned by one generator QQ

  • the only non-trivial bracket is

    {Q,Q}=H \{Q, Q\} = H

Then consider the superpoint

0|2. \mathbb{R}^{0\vert 2} \,.

Its maximal central extension is

the d=3d = 3, N=1N = 1 super Minkowski spacetime

2,1|2 0|2. \array{ \mathbb{R}^{2,1\vert \mathbf{2}} \\ \downarrow \\ \mathbb{R}^{0\vert 2} } \,.
  • whose even part is 3\mathbb{R}^3, spanned by generators P 0,P 1,P 2P_0, P_1, P_2

  • whose odd part is 2\mathbb{R}^2, regarded as

    the Majorana spinor representation 2\mathbf{2}

    of Spin(2,1)SL(2,)Spin(2,1) \simeq SL(2,\mathbb{R})

  • the only non-trivial bracket is the spinor bilinear pairing

    {Q α,Q β}=C ααΓ a α βP a \{Q_\alpha, Q'_\beta\} = C_{\alpha \alpha'} \Gamma_a{}^{\alpha'}{}_\beta \,P^a

where C αβC_{\alpha \beta} is the charge conjugation matrix.


Recall that dd-dimensional central extensions of super Lie algebras 𝔤\mathfrak{g} are classified by 2-cocycles. These are super-skew symmetric bilinear maps

μ 2:𝔤𝔤 d \mu_2 \;\colon\; \mathfrak{g} \wedge\mathfrak{g} \longrightarrow \mathbb{R}^d

satisfying a cocycle condition. The extension 𝔤^\widehat{\mathfrak{g}} that this classifies has underlying super vector space the direct sum

𝔤^𝔤 d \widehat{\mathfrak{g}} \coloneqq \mathfrak{g} \oplus \mathbb{R}^d

an the new super Lie bracket is given on pairs (x,c)𝔤 d(x,c) \in \mathfrak{g} \oplus \mathbb{R}^d


[(x 1,c 1),(x 2,c 2)] μ 2=([x 1,x 2],μ 2(c 1,c 2)). [\; (x_1,c_1), (x_2,c_2)\;]_{\mu_2} \;=\; (\, [x_1,x_2]\,,\, \mu_2(c_1,c_2) \,) \,.

The condition that the new bracket [,] μ 2[-,-]_{\mu_2} satisfies the super Jacobi identity is equivalent to the cocycle condition on μ 2\mu_2.

Now in the case that 𝔤= 0|q\mathfrak{g} = \mathbb{R}^{0\vert q}, the cocycle condition is trivial and a 2-cocycle is just a symmetric bilinear form on the qq fermionic dimensions.

So in the case 𝔤= 0|1\mathfrak{g} = \mathbb{R}^{0\vert 1} there is a unique such, up to scale, namely

μ 2(aQ,bQ)=abP. \mu_2(a Q,b Q) = a b P \,.

But in the case 𝔤= 0|2\mathfrak{g} = \mathbb{R}^{0\vert 2} there is a 3-dimensional space of 2-cocycles, namely

μ 2((Q 1 Q 2),(Q 1 Q 2))={Q 1Q 1, 12(Q 1Q 2+Q 2Q 1), Q 2Q 2 \mu_2 \left( \left( \array{ Q_1 \\ Q_2 }\right), \left( \array{ Q'_1 \\ Q'_2 } \right) \right) = \left\{ \array{ Q_1 Q'_1, & \tfrac{1}{2}\left( Q_1 Q'_2 + Q_2 Q'_1 \right), \\ & Q_2 Q'_2 } \right.

If this is identified with the three coordinates of 3d Minkowski spacetime

2,1(t+x y tx) \mathbb{R}^{2,1} \;\simeq\; \left( \array{ t + x & y \\ & t - x } \right)

then the pairing is the claimed one (see at supersymmetry – in dimensions 3,4,6,10).

On the face of it prop. only produces the super-translation super Lie algebra in 3d, without identifying the fact that its odd components transform as spinors under the spin group (def. ) double cover (prop. ) of the proper orthochronous Lorentz group (def. ). But in fact this information is contained. To see this, consider the following


(external and internal symmetries)

Let 𝔤\mathfrak{g} be a super Lie algebra (def. , prop. ). Its Lie algebra of infinitesimal internal symmetries is the stabilizer of 𝔤 even\mathfrak{g}_{\mathrm{even}} inside the automorphism Lie algebra

𝔦𝔫𝔱(𝔤)Stab 𝔞𝔲𝔱(𝔤) even(𝔤 even), \mathfrak{int}(\mathfrak{g}) \coloneqq \mathrm{Stab}_{\mathfrak{aut}(\mathfrak{g})_{\mathrm{even}}}(\mathfrak{g}_{\mathrm{even}}) \,,

hence is the sub-Lie algebra of derivations Δ\Delta on those which vanish on 𝔤 even𝔤\mathfrak{g}_{\mathrm{even}} \hookrightarrow \mathfrak{g}. This is clearly a normal sub-Lie algebra, so that the quotient

𝔢𝔵𝔱(𝔤)aut(𝔤) even/𝔦𝔫𝔱(𝔤) \mathfrak{ext}(\mathfrak{g}) \coloneqq \mathrm{aut}(\mathfrak{g})_{\mathrm{even}}/\mathfrak{int}(\mathfrak{g})

of all automorphisms by internal ones is again a Lie algebra, the Lie algebra of external symmetries of 𝔤\mathfrak{g}, sitting in a short exact sequence

0𝔦𝔫𝔱(𝔤)𝔞𝔲𝔱(𝔤) even𝔢𝔵𝔱(𝔤)0. 0 \to \mathfrak{int}(\mathfrak{g}) \hookrightarrow \mathfrak{aut}(\mathfrak{g})_{\mathrm{even}} \to \mathfrak{ext}(\mathfrak{g}) \to 0 \,.

Finally, the Lie algebra of simple external automorphisms

𝔢𝔵𝔱 simp(𝔤)𝔢𝔵𝔱(𝔤)𝔞𝔲𝔱(𝔤) \mathfrak{ext}_{\mathrm{simp}}(\mathfrak{g}) \hookrightarrow \mathfrak{ext}(\mathfrak{g}) \hookrightarrow \mathfrak{aut}(\mathfrak{g})

is the maximal semi-simple sub-Lie algebra of the external automorphism Lie algebra.



The internal automorphisms according to def. ) of the super-Minkowski Lie algebra d1,1|N\mathbb{R}^{d-1,1\vert N} (def. ) are called the R-symmetries in the physics literature (e.g. Freed 99, p. 56).


(maximal invariant central extensions)

Let 𝔤\mathfrak{g} be a super Lie algebra (def. , prop. ). Let 𝔥𝔞𝔲𝔱(𝔤) even\mathfrak{h} \hookrightarrow \mathfrak{aut}(\mathfrak{g})_{\mathrm{even}} be a sub-Lie algebra of its automorphism Lie algebra and let

V 𝔤^ 𝔤 \array{ V &\longrightarrow& \widehat{\mathfrak{g}} \\ && \downarrow \\ && \mathfrak{g} }

be a central extension of 𝔤\mathfrak{g} by a vector space VV in even degree. Then we say that: 𝔤^\widehat{\mathfrak{g}} is

  1. an 𝔥\mathfrak{h}-invariant central extension if the 2-cocycles that classify the extension are 𝔥\mathfrak{h}-invariant 2-cocycles,

  2. an invariant central extension if it is 𝔥\mathfrak{h}-invariant and 𝔥=𝔢𝔵𝔱 simp(𝔤)\mathfrak{h} = \mathfrak{ext}_{\mathrm{simp}}(\mathfrak{g}) is the semisimple part of its external automorphism Lie algebra (def. );

1.a maximal 𝔥\mathfrak{h}-invariant central extension if it is an 𝔥\mathfrak{h}-invariant central extension such that the nn-tuple of 𝔥\mathfrak{h}-invariant 2-cocycles that classifies it is a linear basis for the 𝔥\mathfrak{h}-invariant cohomology H 2(𝔤,) 𝔥H^2(\mathfrak{g},\mathbb{R})^{ \mathfrak{h} }


(Huerta-Schreiber 17)

The diagram of super Lie algebras shown on the right

is obtained by consecutively forming maximal invariant central extensions according to def. .

Here d1,1|N\mathbb{R}^{d-1,1\vert \mathbf{N}} is the dd, N\mathbf{N} super-translation supersymmetry algebra from def. .

Moreover, in each case the semisimple part of the external automorphism is the Lie algebra of the corresponding spin group.


That every super Minkowski spacetime is some central extension of some superpoint is elementary. This was highlighted in (Chryssomalakos-Azcárraga-Izquierdo-Bueno 99, 2.1). But most central extensions of superpoints are nothing like super-Minkowksi spacetimes. The point of the above proposition is to restrict attention to iterated invariant central extensions and to find that these single out the super-Minkowski spacetimes.


Hence just from studying iterated invariant central extensions of super Lie algebras, starting with the superpoint, we (re-)discover

  1. Lorentzian geometry,

  2. spin geometry.

  3. super spacetimes.

In the next chapter geometry of physics – fundamental super p-branes we discuss that this process continues through higher central extensions to yield not only super-spacetime, but also the super p-branes propagating on it.

Perhaps we need to understand the nature of time itself better. [...][...] One natural way to approach that question would be to understand in what sense time itself is an emergent concept, and one natural way to make sense of such a notion is to understand how pseudo-Riemannian geometry can emerge from more fundamental and abstract notions such as categories of branes. (G. Moore, p.41 of “Physical Mathematics and the Future”, talk at Strings 2014)


Mathematical discussion of supersymmetry includes

Discussion in the style standard in physics includes

Discussion specifically of (real, Majorana) spin representations includes

The relation between supersymmetry and division algebras was gradually established by a variety of authors, including

  • Taichiro Kugo, Paul Townsend, Supersymmetry and the division algebras, Nuclear Physics B, Volume 221, Issue 2 (1982) p. 357-380. (spires, pdf)

  • A. Sudbery, Division algebras, (pseudo)orthogonal groups and spinors, Jour. Phys. A17 (1984), 939–955.

  • Jonathan Evans, Supersymmetric Yang–Mills theories and division algebras, Nucl. Phys. B298

    (1988), 92–108. Also available as h index?198801412i

  • K.-W. Chung, A. Sudbery, Octonions and the Lorentz and conformal groups of ten-dimensional space-time, Phys. Lett. B 198 (1987), 161–164.

  • Corinne Manogue, A. Sudbery, General solutions of covariant superstring equations of motion, Phys. Rev. D 12 (1989), 4073–4077

  • Jörg Schray, The general classical solution of the superparticle, Class. Quant. Grav. 13 (1996), 27–38. (arXiv:hep-th/9407045)

  • Tevian Dray, J. Janesky, Corinne Manogue, Octonionic hermitian matrices with non-real eigenvalues,

    Adv. Appl. Clifford Algebras 10 (2000), 193–216 (arXiv:math/0006069)

Streamlined proof and exposition regarding supersymmetry and division algebras is in

A neat collection of background on the real normed division algebras themselves is in

  • John Baez, The Octonions, Bull. Amer. Math. Soc. 39 (2002), 145-205. (web)

The derivation of the process of higher invariant extensions that leads from the superpoint to 11-dimensional supergravity:

Last revised on August 9, 2020 at 17:48:01. See the history of this page for a list of all contributions to it.