nLab Yang-Mills instanton



Quantum Field Theory

algebraic quantum field theory (perturbative, on curved spacetimes, homotopical)



field theory:

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Topological physics

Chern-Weil theory



In SU(n)SU(n)-Yang-Mills theory an instanton is a field configuration with non-vanishing second Chern class that minimizes the Yang-Mills energy.


Let (X,g)(X,g) be a compact Riemannian manifold of dimension 4. Let GG be a compact Lie group.

A field configuration of GG-Yang-Mills theory on (X,g)(X,g) is a GG-principal bundle PXP \to X with connection \nabla.

For G=SU(n)G = SU(n) the special unitary group, there is canonically an associated complex vector bundle E=P× G nE = P \times_G \mathbb{C}^n.

Write F Ω 2(X,End(E))F_\nabla \in \Omega^2(X,End(E)) for the curvature 2-form of \nabla.

One says that \nabla is an instanton configuration if F F_\nabla is Hodge-self dual

F =F , \star F_\nabla = F_\nabla \,,

where :Ω k(X)Ω 4k(X)\star : \Omega^k(X) \to \Omega^{4-k}(X) is the Hodge star operator induced by the Riemannian metric gg.

The second Chern class of PP, which by the Chern-Weil homomorphism is given by

c 2(E)= XTr(F F )=kH 4(X,) c_2(E) = \int_X Tr(F_\nabla \wedge F_\nabla) = k \in H^4(X, \mathbb{Z})

is called the instanton number or the instanton sector of \nabla.

Notice that therefore any connection, even if not self-dual, is in some instanton sector, as its underlying bundle has some second Chern class, meaning that it can be obtained from shifting a self-dual connection. The self-dual connections are a convenient choice of “base point” in each instanton sector.


SU(2)SU(2)-instantons from the correct maths to the traditional physics story

The traditional story in the physics textbooks (copied endlessly from one textbook to the next, over generations) of SU(2)-instantons (BPST instantons) tends to fail to highlight some key global points, without which the whole construction really collapses. The following text means to explain the correct description (using the mathematics of Cech cohomology cocycles via the clutching construction on the one-point compactification of Minkowski spacetime) but presented and phrased such that the folklore physics story becomes fully visible – including its crucial fix.



For GG any gauge group (a Lie group), a GG-Yang-Mills instanton on some 4-dimensional spacetime (a pseudo-Riemannian manifold) XX is a GG-principal bundle on XX (whose class is going to be the “instanton sector”), equipped with a GG-principal connection \nabla (this is the actual gauge field) such that this is self-dual, in that its curvature form F F_\nabla satisfies F =F F_\nabla = \star F_{\nabla}, where \star denotes the Hodge star operator for the given metric (which represents the field of gravity).

The ADHM construction gives an essentially unique self-dual principal connection \nabla on every isomorphism class of SU(2)SU(2)-principal bundles (ie.: “in every instanton sector”). Therefore classifying and counting instantons amounts to classifying and counting GG-principal bundles.

This is what makes instantons a “topological” structure in the parlance of physics, meaning that they do not depend on Riemannian metric information, after all.

We now take the spacetime to be Minkowski spacetime 3,1\mathbb{R}^{3,1} and the gauge group to be the special unitary group SU(2)SU(2). This is the case of “BPST-instantons”. Of course other choices are possible and may lead to richer situations, but this simple case is what the physics textbooks tend to focus on (not the least, of course, because these choices are relevant for phenomenology of the weak nuclear force as seen in accelerator experiments) and it already serves to highlight key points. Since, as we just said, we may disregard all metric properties, this means now that we regard spacetime to be just the Cartesian space 4\mathbb{R}^4.

Or almost. At this point one needs to be careful with the boundary conditions in order to get the topology right.


Vanishing at infinity

The physical energy condition on an instanton is that the field strength F F_\nabla vanishes “at infinity”. Mathematically, a continuous function on a locally compact topological space XX vanishes at infinity precisely if it extends to the “one-point compactificationX +X{}X^+ \coloneqq X \cup \{\infty\} of XX – where we literally adjoin the “point at infinity” “\infty” and glue it to XX by defining a suitable topology on X{}X \cup \{\infty\}.

Hence we may formalize the boundary condition by saying that our SU(2)SU(2)-principal connection actually exists on the one-point compactification of 4\mathbb{R}^4. This is the 4-sphere

S 4 4{}. S^4 \simeq \mathbb{R}^4 \cup \{\infty\} \,.

It is this passage from 4\mathbb{R}^4 to S 4S^4 which takes care of the subtleties that often tend to be glossed over.

With the boundary conditions “at infinity” taken care of this way, an SU(2)SU(2)-instanton now is the isomorphism class of an SU(2)SU(2)-principal bundle on the 4-sphere S 4S^4.


Instanton number

There is a classifying space for SU(2)SU(2)-principal bundles, denoted B SU ( 2 ) B SU(2) . This being a classifying space means that for XX any paracompact topological space, we have that the isomorphism classes of SU(2)SU(2)-principal bundles on XX are in bijection with homotopy classes of continuous functions from XX to B SU ( 2 ) B SU(2) .

{XBSU(2)} /homotopy{SU(2)-principal bundles} /gauge transformations. \{X \to B SU(2)\}_{/homotopy} \;\simeq\; \{ SU(2)\text{-principal bundles}\}_{/\text{gauge transformations}} \,.

Now for X=S 4X = S^4, it follows that

{S 4BSU(2)} /homotopyπ 4(BSU(2)) \{S^4 \to B SU(2)\}_{/homotopy} \simeq \pi_4(B SU(2)) \simeq \mathbb{Z}

is the 4th homotopy group of the classifying space B SU ( 2 ) B SU(2) .

This is an integer, and so this integer labels isomorphism classes of instantons on S 4S^4 (hence on 4\mathbb{R}^4). We see below that Chern-Weil theory identifies this number with the instanton number.

But this counting of instantons works more generally, if we use a suitable counting function. First of all, there is a topological space whose only non-trivial homotopy group is π 4\pi_4, and such that this is the group of integers. This is the Eilenberg-MacLane space K(,4)K(\mathbb{Z},4):

{S 4K(,4)} /homotopyand{S d4K(,4)} /homotopy0. \{S^4 \to K(\mathbb{Z},4)\}_{/homotopy} \simeq \mathbb{Z} \;\;\;\;\; \text{and} \;\;\;\;\; \{S^{d\neq 4} \to K(\mathbb{Z},4)\}_{/homotopy} \simeq 0 \,.

This space has the following remarkable property: homotopy classes of continuous functions into it compute ordinary cohomology with integer coefficients:

{XK(,4)} /homotopyH 4(X,). \left\{ X \longrightarrow K(\mathbb{Z},4) \right\}_{/homotopy} \;\simeq\; H^4(X,\mathbb{Z}) \,.

Now there is a continuous function

c 2:BSU(2)K(,4) c_2 \;\colon\; B SU(2) \longrightarrow K(\mathbb{Z},4)

called the universal second Chern class. This hence sends SU(2)SU(2)-principal bundles, classified by some map

χ:XBSU(2) \chi \;\colon\; X \longrightarrow B SU(2)

to classes in degree-4 cohomology

c 2(χ):XBSU(2)c 2K(,4). c_2(\chi) \;\colon\; X \longrightarrow B SU(2) \overset{c_2}{\longrightarrow} K(\mathbb{Z},4) \,.

This cohomology class

c 2(χ)H 4(X,) c_2(\chi) \in H^4(X,\mathbb{Z})

is hence called the second Chern class of the SU(2)SU(2)-principal bundle.

This is one in a whole sequence of characteristic classes which are carried by SU(2)SU(2)-principal bundles, the Chern classes.

But in the special case that the base space XX is 4-dimensional, we have that only a single one of these classes may be non-trivial, namely the second Chern class c 2c_2. Therefore this class completely characterizes SU(2)SU(2)-principal bundles in 4d.

In conclusion: Where a BPST-instanton is manifested by an SU(2)-principal bundle on a 4-dimensional manifold, the instanton number is the second Chern class of this bundle.


Constructing instantons from gauge transformations

We may construct the bundles, that are classified this way, explicitly by using Cech cohomology. This says that we get such a bundle by

  1. choosing an open cover {U iX}\{U_i \to X\} of XX (by “charts”)

  2. choosing transition functions g ij:U iU jSU(2)g_{i j} \colon U_i \cap U_j \to SU(2) on each double overlap of two charts

  3. such that on triple overlaps the cocycle condition g ijg jk=g ikg_{i j} \cdot g_{j k} = g_{i k} holds.

Now comes the major fact which makes this general theory look like the structures that appear in the physics books: the clutching construction.

Namely, for general XX one needs the charts {U iX}\{U_i \to X\} to form a good open cover in order to guarantee that all isomorphism classes of gauge bundles are captured by the construction via transition functions.

But the clutching construction says that whenever XX happens to be a sphere, then it is sufficient to cover it by two hemispheres that overlap a little:

U +upper hemisphere with a little rim U_+ \coloneqq \text{upper hemisphere with a little rim}
U lower hemisphere with a little rim. U_- \coloneqq \text{lower hemisphere with a little rim} \,.

But since everything is topological now, it doesn’t matter that these charts are literally hemi-spheres in the metric sense. In order to get the standard picture we instead make U +U_+ maximally large and take it to cover all of S 4S^4 except the “north pole” (which is really the “point at infinity”, due to the one-point compactification above), while we take U U_- to be a tiny open neighbourhood of that point, sitting there as a tiny ice cap around the north pole. So

U +S 4{point at infinity} U_+ \coloneqq S^4 - \{\text{point at infinity}\}
U tiny neighbourhood of the point at infinity. U_- \coloneqq \text{tiny neighbourhood of the point at infinity} \,.

Topologically this is homeomorphic to the situation before, and hence just as good.

So now back to the general prescription of building principal bundles via Cech cohomology, we are to choose transition functions on all overlaps of charts. But thanks to the clutching construction, there is now just a single such overlap, namely

U +U =tiny annulus around the point at infinity. U_+ \cap U_- = \text{tiny annulus around the point at infinity} \,.

Moreover, there is no non-trivial triple overlap, hence no cocycle condition that our transition function is to satisfy.

In conclusion, Cech cohomology and the clutching construction jointly now say that SU(2)SU(2)-instantons are classified by maps

U +U SU(2) U_+ \cap U_- \longrightarrow SU(2)

But notice here that the intersection of two hemi-4-spheres that overlap slightly is just a 3-sphere times a slight thickening:

U +U S 3×{ϵ,ϵ}. U_+ \cap U_- \simeq S^3 \times \{-\epsilon, \epsilon\} \,.

Moreover, the thickening direction here is trivial, and so one finds that instantons are classified by homotopy classes of maps

g:S 3SU(2) g \;\colon\; S^3 \longrightarrow SU(2)

This is where this “gauge transformation at infinity” in the physics textbooks really comes from. The key point here is that it is indeed gauge transforming, namely the restriction of our bundle to U +U_+ is gauge transformed to its restriction to the “neighbourhood of infinity” U U_-.

But, as topological spaces, SU(2)S 3SU(2) \simeq S^3 and so

{U +U SU(2)} /homotopyπ 3(S 3). \{ U_+ \cap U_- \to SU(2) \}_{/homotopy} \simeq \pi_3(S^3) \simeq \mathbb{Z} \,.

This way we get the same classification from Cech cohomology that we got from classifying space theory, as it must be. But now we know how that second Chern class may be concretely embodied in the “gauge transformation at infinity”.


Gauge fields vanishing at infinity

Now we bring in connections. As discussed before, we may just as well consider any principal connection. In the Cech cohomology picture and still using the clutching construction this now means to choose

  1. 𝔰𝔲(2)\mathfrak{su}(2)-valued 1-forms A ±A_{\pm} on U ±U_\pm

  2. such that on the tiny intersection of the two charts at infinity we have

    A =gA *g 1+gdg 1 A_- = g A_* g^{-1} + g d g^{-1}

Now observe that in the given situation of using the clutching construction on a sphere, we may always choose

A =0. A_- = 0 \,.

Because then the above just says that A +A_+ on U +U_+ becomes gauge trivial “at infinity”, by the given gauge transformation g:S 3SU(2)g \colon S^3 \to SU(2) (the one whose winding number counts our instantons).

(In the present case this is really simple; but also generally there is a partition of unity-argument to construct connections on any bundle in generalization of this simple situation, see here.)


Counting instantons by integrating tr(F F )tr(F_\nabla \wedge F_\nabla)

So far we have derived the physics picture of an instanton: An SU(2)SU(2)-gauge field which becomes gauge trivial “at infinity”, witnessed by a gauge transformation on the “annulus at infinity” S 3SU(2)S^3 \to SU(2), whose winding number is the instanton number. But the key point is that we see that the little neighbourhood of infinity U U_- is part of the picture, and that is necessary now to understand the Chern 4-form.

Namely to every 𝔰𝔲(2)\mathfrak{su}(2)-valued 1-form AA we may assign the ordinary (abelian)differential 4-form

F AF Atr(F AF A), \langle F_A \wedge F_A\rangle \coloneqq tr(F_A \wedge F_A) \,,

where F AF_A is the curvature form of AA.

Now a general fact of Chern-Weil theory is that the 4-form built this way from a single AA is always exact, a potential is given by the Chern-Simons form CS(A)CS(A):

dCS(A)=F AF A. d CS(A) = \langle F_A \wedge F_A\rangle \,.

But beware that this is only true on a single chart. And just because our chart U +U_+ covers “everything except infinity”, we must not forget that there is a second chart U U_-, the “neighbourhood of infinity”.

Namely the 4-form F F \langle F_\nabla \wedge F_\nabla\rangle that is defined on the whole of the 4-sphere S 4S^4, this 4-form is only locally exact (as every closed 4-form is, by the Poincare lemma). Generally, we define it chart-wise by

F F | U +F A +F A + \langle F_{\nabla} \wedge F_\nabla \rangle |_{U_+} \coloneqq \langle F_{A_+} \wedge F_{A_+}\rangle


F F | U F A F A . \langle F_{\nabla} \wedge F_\nabla \rangle |_{U_-} \coloneqq \langle F_{A_-} \wedge F_{A_-}\rangle \,.

This does indeed give a globally defined 4-form, no matter what the local connection forms A ±A_{\pm} are, as long as they satisfy the required condition that they are related by a gauge transformation on the overlap U +U U_+ \cap U_-. Because the 4-form is gauge invariant.

So the 4-form thus defined

F F Ω cl 4(S 4) \langle F_{\nabla} \wedge F_{\nabla} \rangle \in \Omega^4_{cl}(S^4)


  1. a potential 3-form CS(A +)CS(A_+) when restricted to U +U_+;

  2. a potential 3-form CS(A )CS(A_-) when restricted to U U_-;

but it does not have a potential 3-form on all of S 4S^4, unless the instanton number vanishes.

Put this way this should be very obvious now. But it is easy to get confused about this, due to the sheer convenience of the clutching construction used above: we actually were allowed to choose A =0A_- = 0!

This might make it superficially look like there is only a single local gauge potential A +A_+ around, and that the 4-form is locally exact. But this is not the case: there are two local gauge potentials on two different charts, and just because one of them happens to be equal to zero still does not mean that the extension of the 4-form to all of the 4-sphere has a global potential. It has a potential CS(A +)CS(A_+) only after restricting it to U +U_+, even if this means “only removing the point at infinity”.

With it thus being understood that F F \langle F_\nabla \wedge F_\nabla\rangle is not globally exact, it becomes believable that the integral

S 4F F \int_{S^4} \langle F_\nabla \wedge F_\nabla\rangle \;\in\; \mathbb{N}

is generally non-vanishing, and is in fact yet another incarnation of the same integer that we had before, the instanton number. That this is so is given to us by Chern-Weil theory.


Outlook: Chern-Simons 2-Gerbes

In fact the full story is nicer still. Namely the local Chern-Simons 3-forms CS(A ±)CS(A_\pm) together with the gauge transformation at infinity form a Cech cohomology cocycle for a circle 3-bundle with connection (a bundle 2-gerbe). This is the Chern-Simons 2-gerbe of the gauge field. And the fourth incarnation of the instanton number is: the Dixmier-Douady class of this 2-gerbe


As gradient flows between flat connections.

We discuss how Yang-Mills instantons may be understood as trajectories of the gradient flow of the Chern-Simons theory action functional.

Let (Σ,g Σ)(\Sigma,g_\Sigma) be a compact 3-dimensional Riemannian manifold .

Let the cartesian product

X=Σ× X = \Sigma \times \mathbb{R}

of Σ\Sigma with the real line be equipped with the product metric of gg with the canonical metric on \mathbb{R}.

Consider field configurations \nabla of Yang-Mills theory over Σ×\Sigma \times \mathbb{R} with finite Yang-Mills action

S YM()= Σ×F F <. S_{YM}(\nabla) = \int_{\Sigma \times \mathbb{R}} F_\nabla \wedge \star F_\nabla \,\,\lt \infty \,.

These must be such that there is t 1<t 2t_1 \lt t_2 \in \mathbb{R} such that F (t<t 1)=0F_\nabla(t \lt t_1) = 0 and F (T>t 2)=0F_\nabla(T \gt t_2) = 0, hence these must be solutions interpolating between two flat connections t 1\nabla_{t_1} and t 2\nabla_{t_2}.

For AΩ 1(U×,𝔤)A \in \Omega^1(U\times \mathbb{R}, \mathfrak{g}) the Lie algebra valued 1-form corresponding to \nabla, we can always find a gauge transformation such that A t=0A_{\partial_t} = 0 (“temporal gauge”). In this gauge we may hence equivalently think of AA as a 1-parameter family

tA(t)Ω 1(Σ,𝔤) t \mapsto A(t) \in \Omega^1(\Sigma, \mathfrak{g})

of connections on Σ\Sigma. Then the self-duality condition on a Yang-Mills instanton

F =F F_\nabla = - \star F_\nabla

reads equivalently

ddtA= gF AΩ 1(Σ,𝔤). \frac{d}{d t} A = -\star_{g} F_A \,\,\, \in \Omega^1(\Sigma, \mathfrak{g}) \,.

On the linear configuration space Ω 1(Σ,𝔤)\Omega^1(\Sigma, \mathfrak{g}) of Lie algebra valued forms on Σ\Sigma define the Hodge inner product metric

G(α,β):= Σα gβ, G(\alpha, \beta) := \int_{\Sigma} \langle \alpha \wedge \star_g \beta \rangle \,,

where ,:𝔤𝔤\langle-,-\rangle : \mathfrak{g} \otimes \mathfrak{g} \to \mathbb{R} is the Killing form invariant polynomial on the Lie algebra 𝔤\mathfrak{g}.


The instanton equation

ddtA= gF A \frac{d}{d t} A = -\star_{g} F_A

is the equation characterizing trajectories of the gradient flow of the Chern-Simons action functional

S CS:Ω 1(Σ,𝔤) S_{CS} : \Omega^1(\Sigma, \mathfrak{g}) \to \mathbb{R}
A ΣCS(A) A \mapsto \int_\Sigma CS(A)

with respect to the Hodge inner product metric on Ω 1(Σ,𝔤)\Omega^1(\Sigma,\mathfrak{g}).


The variation of the Chern-Simons action is

δS CS(A)= ΣδAF A \delta S_{CS}(A) = \int_\Sigma \langle \delta A \wedge F_A\rangle

(see Chern-Simons theory for details).

In other words, we have the 1-form on Ω 1(Σ,𝔤)\Omega^1(\Sigma,\mathfrak{g}):

δS CS() A= ΣF A. \delta S_{CS}(-)_A = \int_\Sigma \langle - \wedge F_A \rangle \,.

The corresponding gradient vector field

S CS:=G 1δS CS \nabla S_{CS} := G^{-1} \delta S_{CS}

is uniquely defined by the equation

δS CS() =G(,S CS) Σ,S CS. \begin{aligned} \delta S_{CS}(-) & = G(-,\nabla S_{CS}) \\ \int_\Sigma \langle - , \star \nabla S_{CS}\rangle \end{aligned} \,.

With the formula (see Hodge star operator)

A=(1) 1(3+1)A=A \star \star A = (-1)^{1(3+1)} A = A

we find therefore

S CS= gF A. \nabla S_{CS} = \star_g F_A \,.

Hence the gradient flow equation

ddtA+S CSA=0 \frac{d}{d t} A + \nabla S_{CS}_A = 0

is indeed

ddtA= gF A. \frac{d}{d t} A = - \star_g F_A \,.

Since flat connections are the critical loci of S CSS_{CS} this says that a finite-action Yang-Mills instanton on Σ×\Sigma \times \mathbb{R} is a gradient flow trajectory between two Chern-Simons theory vacua .

Often this is interpreted as saying that “a Yang-Mills instanton describes the tunneling between two Chern-Simons theory vacua”.

As Dp-D(p+4)-brane bound states

Due to the higher WZW term D p+4C p+1FF\propto \int_{D_{p+4}} C_{p+1} \wedge \langle F \wedge F \rangle in the Green-Schwarz sigma model for D(p+4)-branes, Yang-Mills instantons in the Chan-Paton gauge field on D(p+4)D (p+4)-branes are equivalently Dp-D(p+4)-brane bound states (see e.g. Polchinski 96, 5.4, Tong 05, 1.4).

The lift to M-theory as M5-MO9 brane bound states is due to Strominger 90, Witten 96.

Dp-D(p+4)-brane bound states and Yang-Mills instantons



The ADHM construction:


Introductions and surveys:

A survey in view of the asymptotic nature of the Feynman perturbation series is in

For a fairly comprehensive list of literature see the bibliography of

  • Marcus Hutter, Instantons in QCD: Theory and Application of the Instanton Liquid Model (arXiv:hep-ph/0107098)

Identification of the moduli space of instantons with that of holomorphic maps from the Riemann sphere to the loop group of the gauge group:

Detailed argument for the theta-vacuum structure from chiral symmetry breaking is offered in

The multi-instantons in SU(2)SU(2)-Yang-Mills theory (BPST instantons) were discovered in

  • A. A. Belavin, A.M. Polyakov, A.S. Schwartz, Yu.S. Tyupkin, Pseudoparticle solutions of the Yang-Mills equations, Phys. Lett. B 59 (1), 85-87 (1975) doi

  • A. A. Belavin, V.A. Fateev, A.S. Schwarz, Yu.S. Tyupkin, Quantum fluctuations of multi-instanton solutions, Phys. Lett. B 83 (3-4), 317-320 (1979) doi

See also

  • Michael Atiyah, Nigel Hitchin, J. M. Singer, Deformations of instantons, Proc. Nat. Acad. Sci. U.S. 74, 2662 (1977)

  • Edward Witten, Some comments on the recent twistor space constructions, Complex manifold techniques in theoretical physics (Proc. Workshop, Lawrence, Kan., 1978), pp. 207–218, Res. Notes in Math., 32, Pitman, Boston, Mass.-London, 1979.

Methods of algebraic geometry were introduced in

  • M. F. Atiyah, R. S. Ward, Instantons and algebraic geometry, Comm. Math. Phys. 55, n. 2 (1977), 117-124, MR0494098, euclid

The more general ADHM construction in terms of linear algebra of vector bundles on projective varieties is proposed in

  • M.F. Atiyah, N.J. Hitchin, V.G. Drinfeld, Yu.I. Manin, Construction of instantons, Physics Letters 65 A, 3, 185–187 (1978) pdf

Monographs with the standard material include

  • Dan Freed, Karen Uhlenbeck, Instantons and four-manifolds, Springer-Verlag, (1991)

  • Robbert Dijkgraaf, Topological gauge theories and group cohomology (ps)

  • Nicholas Manton, Paul M. Sutcliffe, Topological solitons, Cambridge Monographs on Math. Physics, gBooks

Yang-Mills instantons on spaces other than just spheres are explicitly discussed in

  • Gabor Kunstatter, Yang-mills theory in a multiply connected three space, Mathematical Problems in Theoretical Physics: Proceedings of the VIth International Conference on Mathematical Physics Berlin (West), August 11-20,1981. Editor: R. Schrader, R. Seiler, D. A. Uhlenbrock, Lecture Notes in Physics, vol. 153, p.118-122 (web)

based on


  • Henrique N. Sá Earp, Instantons on G 2G_2−manifolds PhD thesis (2009) (pdf)

is a discussion of Yang-Mills instantons on a 7-dimensional manifold with special holonomy.

As Dp-D(p+4)-brane bound states

The argument that Yang-Mills instantons in the Chan-Paton gauge field on a D(p+4)-brane are equivalent to Dp-D(p+4) brane bound states goes back to


Review is in:

Discussion specifically of D0-D4-brane bound states:

with emphasis to the resulting configuration spaces of points, as in

Discussion specifically of D1-D5-brane bound states

Discussion specifically of D4-D8-brane bound states:

In the Witten-Sakai-Sugimoto model geometrically engineering QCD, where the D4-branes get interpreted as baryons:

Last revised on April 2, 2024 at 13:46:05. See the history of this page for a list of all contributions to it.