# Holography at Work for Nuclear and Hadron Physics

Authors: Youngman Kim; Deokhyun Yi
Journal: Advances in High Energy Physics
 (2011)

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## Abstract

The purpose of this review is to provide basic ingredients of holographic QCD to nonexperts in string theory and to summarize its interesting achievements in nuclear and hadron physics. We focus on results from a less stringy bottom-up approach and review a stringy
top-down model with some calculational details.


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## Body

## 1. Introduction

The approaches based on the Anti de Sitter/conformal field theory (AdS/CFT) correspondence [1–3] find many interesting possibilities to explore strongly interacting systems. The discovery of D-branes in string theory [4] was a crucial ingredient to put the correspondence on a firm footing. Typical examples of the strongly interacting systems are dense baryonic matter, stable/unstable nuclei, strongly interacting quark gluon plasma, and condensed matter systems. The morale is to introduce an additional space, which roughly corresponds to the energy scale of 4D boundary field theory, and try to construct a 5D holographic dual model that captures certain nonperturbative aspects of strongly coupled field theory, which are highly nontrivial to analyze in conventional quantum field theory based on perturbative techniques. There are in general two different routes to modeling holographic dual of quantum chromodynamics (QCD). One way is a top-down approach based on stringy D-brane configurations. The other way is so-called a bottom-up approach to a holographic, in which a 5D holographic dual is constructed from QCD. Despite the fact that this bottom-up approach is somewhat ad hoc, it reflects some important features of the gauge/gravity duality and is rather successful in describing properties of hadrons. However, we should keep in mind that a usual simple, tree-level analysis in the holographic dual model, both top-down and bottom-up, is capturing the leading Nc contributions, and we are bound to suffer from subleading corrections.The goal of this review is twofold. First, we will assemble results mostly from simple bottom-up models in nuclear and hadron physics. Surely we cannot have them all here. We will devote to selected physical quantities discussed in the bottom-up model. The selection of the topics is based on authors’ personal bias. Second, we present some basic materials that might be useful to understand some aspects of AdS/CFT and D-brane models. We will focus on the role of the AdS/CFT in low-energy QCD. Although the correspondence between QCD and gravity theory is not known, we can obtain much insights on QCD by the gauge/gravity duality.We organize this review as follows. Section2 reviews the gauge/gravity. Section 3 briefly discusses developments of holographic QCD and demonstrates how to build up a bottom-up model using the AdS/CFT dictionary. After discussing the gauge/gravity duality and modeling in the bottom-up approach, we proceed with selected physical quantities. In each section, we show results mostly from the bottom-up approach and list some from the top-down model. Section 4 deals with vacuum condensates of QCD in holographic QCD. We will mainly discuss the gluon condensate and the quark-gluon mixed condensate. Section 5 collects some results on hadron spectroscopy and form factors from the bottom-up model. Contents are glueballs, light mesons, heavy quarkonium, and hadron form-factors. Section 6 is about QCD at finite temperature and density. We consider QCD phase transition and dense matter. Section 7 is devoted to some general remarks on holographic QCD and to list a few topics that are not discussed properly in this article. Due to our limited knowledge, we are not able to cover all interesting works done in holographic QCD. To compensate this defect partially, we will list some recent review articles on holographic QCD.In AppendicesA–F, we look back on some basic materials that might be useful for nonexperts in string theory to work in holographic QCD. In Appendix A: we review the relation between the bulk mass and boundary operator dimension. In  Appendix B: we present a D3/D7 model and axial U(1) symmetry in the model. In Appendix C: we discuss non-Abelian chiral symmetry based on D4/D8/D8¯ model.  In  Appendices D and E: we describe how to calculate the Hawking temperature of an AdS black hole. In Appendix F: we encapsulate the Hawking-Page transition and sketch how to calculate Polyakov loop expectation value in thermal AdS and AdS black hole.We close this section with a cautionary remark.Though it is tempting to argue that holographic QCD is dual to real QCD, what we mean by QCD here might be mostly QCD-like or a cousin of QCD.


## 2. Introduction to the AdS/CFT Correspondence

The AdS/CFT correspondence, first suggested by Maldacena [1], is a duality between gravity theory in anti de Sitter space (AdS) background and conformal field theory (CFT). The original conjecture states that there is a correspondence between a weakly coupled gravity theory (type IIB string theory) on AdS5×S5 and the strongly coupled 𝒩=4 supersymmetric Yang-Mills theory on the four-dimensional boundary of AdS5. The strings reside in a higher-dimensional curved spacetime and there exists some well-defined mapping between the objects in the gravity side and the dual objects in the four-dimensional gauge theory. Thus, the conjecture allows the use of non-perturbative methods for strongly coupled theory through its gravity dual.


### 2.1.Dp Brane Dynamics

The duality emerges from a careful consideration of the D-brane dynamics. ADp brane sweeps out (p+1) world-volume in spacetime. Introducing D branes gives open string modes whose endpoints lie on the D branes and the open string spectrum consists of a finite number of massless modes and also an infinite tower of massive modes. The open string end points can move only in the parallel (p+1) directions of the brane, see Figure 1(a), and a Dp brane can be seen as a point along its transverse directions. The dynamics of the Dp brane is described by the Dirac-Born-Infeld (DBI) action [5] and Chern-Simons term:(2.1)SDp=-Tp∫dp+1xe-ϕ-det⁡(P[g]ab+2πα′Fab)+SCS
with a dilaton e-ϕ. Here gab is the induced metric on Dp. P denotes the pullback and Fab is the world-volume field strength. Tp is the tension of the brane which has the following  form:(2.2)Tp=1(2π)pgslsp+1=1(2π)pgsα′(p+1)/2,
and it is the mass per unit spatial volume. Here gs is the string coupling and ls is the string length. α′ is the Regge slope parameter and related to the string length scale as ls=α′. In general, states in the closed string spectrum contain a finite number of massless modes and an infinite tower of massive modes with masses of order ms=ls-1=α′-1/2. Thus, at low energies E≪ms, the higher-order corrections come in powers of α′E2 from integrating out the massive string modes. If there are Nc stack of multiple D branes, the open strings between different branes give a non-Abelian U(Nc) gauge group; see Figure 1(b). In the low-energy limit, we can integrate out the massive modes to obtain non-Abelian gauge theory of the massless fields.The configurations ofNc stack of D3 branes in 10d spacetime. (a) D3 branes sweep (3+1) dimensions in (9+1) space time. (b) Nc = 3 stack of D3 branes and all the possible classes of open strings.
(a)(b)Now, we takep=3 and consider Nc D3 brane stacks in type IIB theory. The low-energy effective action of this configuration gives a non-Abelian gauge theory with U(Nc) gauge group. In addition, this gauge group can be factorized into U(Nc)=SU(Nc)×U(1) and the U(1) part, which describes the center of mass motion of the D3 branes, can be decoupled by the global translational invariance. The remaining subgroup SU(Nc) describes the dynamics of branes from each other. Therefore we see that in the low-energy limit, the massless open string modes on Nc stacks of D3-branes constitute 𝒩=4SU(Nc) Yang-Mills theory [6] with 16 supercharges in (3+1) spacetime. From (2.1) for p=3, we obtain the effective Lagrangian at low energies up to the two-derivative order:(2.3)L=-14πgsTr⁡(14FμνFμν+12DμϕiDμϕi-14[ϕi,ϕj]2+i2Ψ¯IΓμDμΨI-i2Ψ¯IΓi[ϕi,ΨI])
with a gauge field Aμ, six scalar fields ϕi, and four Weyl fermions ΨI.In fact, the original system also contains closed string states. The higher-order derivative corrections for the Lagrangian (2.3) come both in powers of α′E2 from the massive modes and powers of the string coupling gsNc for loop corrections. It is known that the string coupling constant gs is related by the 10-dimensional gravity constant as G(10)~gs2ls8 and thus the dimensionless string coupling is of order G(10)E8, which is negligible in the low-energy limit. Therefore, at low energies closed strings are decoupled from open strings and the physics on the Nc D3 branes is described by the massless 𝒩=4 super Yang-Mills theory with gauge group SU(Nc).


### 2.2.AdS5×S5 Geometry

Now we view the same system from a different angle. Since D branes are massive and carry energy and Ramond-Ramond (RR) charge,Nc D3 branes deform the spacetime around them to make a curved geometry. Note that the total mass of a D3 brane is infinite because it occupies the infinite world-volume of its transverse directions, but the tension, or the mass per unit three-volume of the D3 brane(2.4)T3=1(2π)3gsls4
is finite.In the flat spacetime, the circumference of the circle surrounding an origin at a distancer is 2πr, and it simply shrinks to zero if one approaches the origin. But if there is a stack of D3 branes, it deforms the spacetime and makes throat geometry along its transverse directions. Thus, near the D3 branes, the radius of a circle around the stack approaches a constant R, an asymptotical infinite cylinder structure, or AdS5×S5; see Figure 2(b). The Nc D3 brane stack is located at the infinite end of the throat and this infinite end is called the “horizon”. In the near horizon geometry, a D3 brane is surrounded by a five-dimensional sphere S5.The two descriptions of theNc D3 configuration. (a) Flat spacetime for r≫R. Nc D3 branes deform the spacetime.
(a)(b)To be more specific, let us start with type IIB string theory forp=3. We find a black hole type solution which is carrying charges with respect to the RR four-form potential. The theory has magnetically charged D3 branes, which are electrically charged under the potential dA4 and it is self-dual *F5=F5. The low-energy effective action is(2.5)S=1(2π)7ls8∫d10x-g(e-2ϕ(R+4(∇φ)2)-25!F52).
We assume that the metric is spherically symmetric in seven dimensions with the RR source at the origin; then the Nc parameter appears in terms of the five-form field RR-field strength on the five-sphere as(2.6)∫S5*F5=Nc,
where S5 is the five-sphere surrounding the source for a four-form field C4. Now by using the Euclidean symmetry we get the curved metric solution [7–9] for the D3 brane:(2.7)ds2=f(r)-1/2ημνdxμdxν+f(r)1/2(dr2+r2dΩ52),
where(2.8)f(r)=1+R4r4
with the radius of the horizon R:(2.9)R2=4πgsNcα′=4πgsNcls2.dΩ5 is the five-sphere metric. For r≫R we have f(r)≃1 and the spacetime becomes flat with a small correction R4/r4=4πgsNcls4/r4. This factor can be interpreted as a gravitational potential since G(10)~gs2ls8 and MD3~NcT3~Nc/gsls4  and thus R4/r4~GMD3/r4. In the near horizon limit, this gravitational effect becomes strong and the metric changes into(2.10)ds2=r2R2ημνdxμdxν+R2r2(dr2+r2dΩ52).
This is AdS5×S5.The geometry by the D3 branes is sketched in Figure2(b). Far away from the D3 brane stacks, the spacetime is flat (9+1)-dimensional Minkowski spacetime and the only modes which survive in the low-energy limit are the massless-closed string (graviton) multiplets, and they decouple from each other due to weak interactions. On the other hand, close to the D3 branes, the geometry takes the form AdS5×S5 and the whole tower of massive modes exists there. This is because the excitations seen from an observer at infinity are close to the horizon and a closed string mode in a throat should go over a gravitational potential to meet the asymptotic flat region. Therefore, as we focus on  the lower-energy limit, the excitation modes should be originated deeper in the throat, and then they decouple from the ones in the flat region. Thus in the low-energy limit the interacting sector lives in AdS5×S5 geometry.


### 2.3. The Gauge/Gravity Duality

So far, we have considered two seemingly different descriptions of theNc D3 brane configuration. As we mentioned, each of the D3 branes carries the gravitational degrees of freedom in terms of its tension, or the string coupling gs as in (2.4). So the strength of the gravity effect due to Nc stacks of D3 branes depends on the parameter gsNc.IfgsNc≪1, from (2.9) we see that R≪ls and therefore the throat geometry effect is less than string length scale. Thus the spacetime is nearly flat and the fluctuations of the D3 branes are described by open string states. In this regime the string coupling gs is small and the closed strings are decoupled from the open strings. Here the closed string description is inapplicable since one needs to know about the geometry below the string length scale. If we take the low-energy limit, the effective theory, which describes the open string modes, is 𝒩=4 super Yang-Mills theory with SU(Nc) gauge group.On the other hand, ifgsNc≫1, then the back-reaction of the branes on the background becomes important and spacetime will be curved. In this limit the closed string description reduces to classical gravity which is supergravity theory in the near horizon geometry. Here the open string description is not feasible because gsNc is related with the loop corrections and one has to deal with the strongly coupled open strings. Again, if we take the low-energy limit, the interaction is described by the type IIB string theory in the near-horizon geometry, AdS5×S5.The gauge/gravity correspondence is nothing but the conjecture connecting these two descriptions ofNc D3 branes in the low-energy limit. It is a duality between the 𝒩=4 super-Yang-Mills theory with gauge group SU(Nc) and the type IIB closed string theory in AdS5×S5; see Figure 3. The relation between the Yang-Mills coupling gYM and the string coupling strength gs is given by(2.11)gYM2=4πgs,(Rls)4=4πgsNc.
Then, the ‘t Hooft coupling λ=gYM2Nc can be expressed in terms of the string length scale:(2.12)λ=Rls44.
Therefore, the dependence of gsNc becomes the question of whether the ‘t Hooft coupling is large or small, or the gauge theory is strongly or weakly coupled.Figure 3
The sketch of the AdS/CFT correspondence.The two descriptions can be viewed as two extremes ofr. For the sake of convenience, we use the coordinate z=R2/r. Then, the AdS5×S5 metric (2.10) becomes(2.13)ds2=R2z2(ημνdxμdxν+dz2)+R2dΩ52,
which shows the conformal equivalence between AdS5 and flat spacetime more clearly. In (2.13), each z-slice of AdS5 is isometric to four-dimensional Minkowski spacetime. In this coordinate, z=0 is the boundary of AdS5, where Yang-Mills theory lives, with identifying xμ as the coordinates of the gauge theory. If z→∞, the determinant of the metric goes to zero and it is the Poincaré horizon. Here the factor R2/z2 also has some relation with the energy scales. If the gauge theory side has a certain energy scale E, the corresponding energy in the gravity side is (z/R)E. In other words, a gauge theory object with an energy scale E is involved with a bulk side one localized in the z-direction at z~1/E [1, 10, 11]. Therefore, the UV or high-energy limit corresponds to z→0 (or r→∞) and the IR or low-energy limit corresponds to z→∞ (or r→0).The operator-field correspondence between operators in the four-dimensional gauge theory and corresponding dual fields in the gravity side was given in [2, 3]. Then the AdS/CFT correspondence can be stated as follows,(2.14)〈Te∫d4xϕ0(x)O(x)〉CFT=Zsugra,
where ϕ0(x)=ϕ(x,u→∞) and the string theory partition function Zsugra at the boundary specified by ϕ0 has the form(2.15)Zsugra=e-Ssugra(ϕ(x,u))|u→∞.
The relation (2.14) implies that the generating functional of gauge-invariant operators in CFT can be matched with the generating functional for tree diagrams in supergravity.


## 2.1.Dp Brane Dynamics

The duality emerges from a careful consideration of the D-brane dynamics. ADp brane sweeps out (p+1) world-volume in spacetime. Introducing D branes gives open string modes whose endpoints lie on the D branes and the open string spectrum consists of a finite number of massless modes and also an infinite tower of massive modes. The open string end points can move only in the parallel (p+1) directions of the brane, see Figure 1(a), and a Dp brane can be seen as a point along its transverse directions. The dynamics of the Dp brane is described by the Dirac-Born-Infeld (DBI) action [5] and Chern-Simons term:(2.1)SDp=-Tp∫dp+1xe-ϕ-det⁡(P[g]ab+2πα′Fab)+SCS
with a dilaton e-ϕ. Here gab is the induced metric on Dp. P denotes the pullback and Fab is the world-volume field strength. Tp is the tension of the brane which has the following  form:(2.2)Tp=1(2π)pgslsp+1=1(2π)pgsα′(p+1)/2,
and it is the mass per unit spatial volume. Here gs is the string coupling and ls is the string length. α′ is the Regge slope parameter and related to the string length scale as ls=α′. In general, states in the closed string spectrum contain a finite number of massless modes and an infinite tower of massive modes with masses of order ms=ls-1=α′-1/2. Thus, at low energies E≪ms, the higher-order corrections come in powers of α′E2 from integrating out the massive string modes. If there are Nc stack of multiple D branes, the open strings between different branes give a non-Abelian U(Nc) gauge group; see Figure 1(b). In the low-energy limit, we can integrate out the massive modes to obtain non-Abelian gauge theory of the massless fields.The configurations ofNc stack of D3 branes in 10d spacetime. (a) D3 branes sweep (3+1) dimensions in (9+1) space time. (b) Nc = 3 stack of D3 branes and all the possible classes of open strings.
(a)(b)Now, we takep=3 and consider Nc D3 brane stacks in type IIB theory. The low-energy effective action of this configuration gives a non-Abelian gauge theory with U(Nc) gauge group. In addition, this gauge group can be factorized into U(Nc)=SU(Nc)×U(1) and the U(1) part, which describes the center of mass motion of the D3 branes, can be decoupled by the global translational invariance. The remaining subgroup SU(Nc) describes the dynamics of branes from each other. Therefore we see that in the low-energy limit, the massless open string modes on Nc stacks of D3-branes constitute 𝒩=4SU(Nc) Yang-Mills theory [6] with 16 supercharges in (3+1) spacetime. From (2.1) for p=3, we obtain the effective Lagrangian at low energies up to the two-derivative order:(2.3)L=-14πgsTr⁡(14FμνFμν+12DμϕiDμϕi-14[ϕi,ϕj]2+i2Ψ¯IΓμDμΨI-i2Ψ¯IΓi[ϕi,ΨI])
with a gauge field Aμ, six scalar fields ϕi, and four Weyl fermions ΨI.In fact, the original system also contains closed string states. The higher-order derivative corrections for the Lagrangian (2.3) come both in powers of α′E2 from the massive modes and powers of the string coupling gsNc for loop corrections. It is known that the string coupling constant gs is related by the 10-dimensional gravity constant as G(10)~gs2ls8 and thus the dimensionless string coupling is of order G(10)E8, which is negligible in the low-energy limit. Therefore, at low energies closed strings are decoupled from open strings and the physics on the Nc D3 branes is described by the massless 𝒩=4 super Yang-Mills theory with gauge group SU(Nc).


## 2.2.AdS5×S5 Geometry

Now we view the same system from a different angle. Since D branes are massive and carry energy and Ramond-Ramond (RR) charge,Nc D3 branes deform the spacetime around them to make a curved geometry. Note that the total mass of a D3 brane is infinite because it occupies the infinite world-volume of its transverse directions, but the tension, or the mass per unit three-volume of the D3 brane(2.4)T3=1(2π)3gsls4
is finite.In the flat spacetime, the circumference of the circle surrounding an origin at a distancer is 2πr, and it simply shrinks to zero if one approaches the origin. But if there is a stack of D3 branes, it deforms the spacetime and makes throat geometry along its transverse directions. Thus, near the D3 branes, the radius of a circle around the stack approaches a constant R, an asymptotical infinite cylinder structure, or AdS5×S5; see Figure 2(b). The Nc D3 brane stack is located at the infinite end of the throat and this infinite end is called the “horizon”. In the near horizon geometry, a D3 brane is surrounded by a five-dimensional sphere S5.The two descriptions of theNc D3 configuration. (a) Flat spacetime for r≫R. Nc D3 branes deform the spacetime.
(a)(b)To be more specific, let us start with type IIB string theory forp=3. We find a black hole type solution which is carrying charges with respect to the RR four-form potential. The theory has magnetically charged D3 branes, which are electrically charged under the potential dA4 and it is self-dual *F5=F5. The low-energy effective action is(2.5)S=1(2π)7ls8∫d10x-g(e-2ϕ(R+4(∇φ)2)-25!F52).
We assume that the metric is spherically symmetric in seven dimensions with the RR source at the origin; then the Nc parameter appears in terms of the five-form field RR-field strength on the five-sphere as(2.6)∫S5*F5=Nc,
where S5 is the five-sphere surrounding the source for a four-form field C4. Now by using the Euclidean symmetry we get the curved metric solution [7–9] for the D3 brane:(2.7)ds2=f(r)-1/2ημνdxμdxν+f(r)1/2(dr2+r2dΩ52),
where(2.8)f(r)=1+R4r4
with the radius of the horizon R:(2.9)R2=4πgsNcα′=4πgsNcls2.dΩ5 is the five-sphere metric. For r≫R we have f(r)≃1 and the spacetime becomes flat with a small correction R4/r4=4πgsNcls4/r4. This factor can be interpreted as a gravitational potential since G(10)~gs2ls8 and MD3~NcT3~Nc/gsls4  and thus R4/r4~GMD3/r4. In the near horizon limit, this gravitational effect becomes strong and the metric changes into(2.10)ds2=r2R2ημνdxμdxν+R2r2(dr2+r2dΩ52).
This is AdS5×S5.The geometry by the D3 branes is sketched in Figure2(b). Far away from the D3 brane stacks, the spacetime is flat (9+1)-dimensional Minkowski spacetime and the only modes which survive in the low-energy limit are the massless-closed string (graviton) multiplets, and they decouple from each other due to weak interactions. On the other hand, close to the D3 branes, the geometry takes the form AdS5×S5 and the whole tower of massive modes exists there. This is because the excitations seen from an observer at infinity are close to the horizon and a closed string mode in a throat should go over a gravitational potential to meet the asymptotic flat region. Therefore, as we focus on  the lower-energy limit, the excitation modes should be originated deeper in the throat, and then they decouple from the ones in the flat region. Thus in the low-energy limit the interacting sector lives in AdS5×S5 geometry.


## 2.3. The Gauge/Gravity Duality

So far, we have considered two seemingly different descriptions of theNc D3 brane configuration. As we mentioned, each of the D3 branes carries the gravitational degrees of freedom in terms of its tension, or the string coupling gs as in (2.4). So the strength of the gravity effect due to Nc stacks of D3 branes depends on the parameter gsNc.IfgsNc≪1, from (2.9) we see that R≪ls and therefore the throat geometry effect is less than string length scale. Thus the spacetime is nearly flat and the fluctuations of the D3 branes are described by open string states. In this regime the string coupling gs is small and the closed strings are decoupled from the open strings. Here the closed string description is inapplicable since one needs to know about the geometry below the string length scale. If we take the low-energy limit, the effective theory, which describes the open string modes, is 𝒩=4 super Yang-Mills theory with SU(Nc) gauge group.On the other hand, ifgsNc≫1, then the back-reaction of the branes on the background becomes important and spacetime will be curved. In this limit the closed string description reduces to classical gravity which is supergravity theory in the near horizon geometry. Here the open string description is not feasible because gsNc is related with the loop corrections and one has to deal with the strongly coupled open strings. Again, if we take the low-energy limit, the interaction is described by the type IIB string theory in the near-horizon geometry, AdS5×S5.The gauge/gravity correspondence is nothing but the conjecture connecting these two descriptions ofNc D3 branes in the low-energy limit. It is a duality between the 𝒩=4 super-Yang-Mills theory with gauge group SU(Nc) and the type IIB closed string theory in AdS5×S5; see Figure 3. The relation between the Yang-Mills coupling gYM and the string coupling strength gs is given by(2.11)gYM2=4πgs,(Rls)4=4πgsNc.
Then, the ‘t Hooft coupling λ=gYM2Nc can be expressed in terms of the string length scale:(2.12)λ=Rls44.
Therefore, the dependence of gsNc becomes the question of whether the ‘t Hooft coupling is large or small, or the gauge theory is strongly or weakly coupled.Figure 3
The sketch of the AdS/CFT correspondence.The two descriptions can be viewed as two extremes ofr. For the sake of convenience, we use the coordinate z=R2/r. Then, the AdS5×S5 metric (2.10) becomes(2.13)ds2=R2z2(ημνdxμdxν+dz2)+R2dΩ52,
which shows the conformal equivalence between AdS5 and flat spacetime more clearly. In (2.13), each z-slice of AdS5 is isometric to four-dimensional Minkowski spacetime. In this coordinate, z=0 is the boundary of AdS5, where Yang-Mills theory lives, with identifying xμ as the coordinates of the gauge theory. If z→∞, the determinant of the metric goes to zero and it is the Poincaré horizon. Here the factor R2/z2 also has some relation with the energy scales. If the gauge theory side has a certain energy scale E, the corresponding energy in the gravity side is (z/R)E. In other words, a gauge theory object with an energy scale E is involved with a bulk side one localized in the z-direction at z~1/E [1, 10, 11]. Therefore, the UV or high-energy limit corresponds to z→0 (or r→∞) and the IR or low-energy limit corresponds to z→∞ (or r→0).The operator-field correspondence between operators in the four-dimensional gauge theory and corresponding dual fields in the gravity side was given in [2, 3]. Then the AdS/CFT correspondence can be stated as follows,(2.14)〈Te∫d4xϕ0(x)O(x)〉CFT=Zsugra,
where ϕ0(x)=ϕ(x,u→∞) and the string theory partition function Zsugra at the boundary specified by ϕ0 has the form(2.15)Zsugra=e-Ssugra(ϕ(x,u))|u→∞.
The relation (2.14) implies that the generating functional of gauge-invariant operators in CFT can be matched with the generating functional for tree diagrams in supergravity.


## 3. Holographic QCD

Ever since the advent of the AdS/CFT correspondence, there have been many efforts, based on the correspondence, to study nonperturbative physics of strongly coupled gauge theories in general and QCD in particular.Witten proposed [12] that we can extend the correspondence to non-supersymmetric theories by considering the AdS black hole and showed that this supergravity treatment qualitatively well describes strong coupled QCD (or QCD-like) at finite temperature: for instance, the area law behavior of Wilson loops, confinement/deconfinement transition of pure gauge theory through the Hawking-Page transition, and the mass gap for glueball states. In [13] symmetry breaking by expectation values of scalar fields were analyzed in the context of the AdS/CFT correspondence, which is essential to encode the spontaneous breaking of chiral symmetry in a holographic QCD model. Regular supergravity backgrounds with less supersymmetries corresponding to dual confining 𝒩=1 super-Yang-Mills theories were proposed in [14, 15]. It has been shown by Polchinski and Strassler [16] that the scaling of high-energy QCD scattering amplitudes can be obtained from a gravity dual description in a sliced AdS geometry whose IR cutoff is determined by the mass of the lightest glueball. Important progress towards flavor physics of QCD has been made by adding flavor degrees of freedom in the fundamental representation of a gauge group to the gravity dual description [17]. Chiral symmetry breaking and meson spectra were studied in a nonsupersymmetric gravity model dual to large Nc nonsupersymmetric gauge theories [18], where flavor quarks are introduced by a D7-brane probe on deformed AdS backgrounds. Using a D4/D6 brane configuration, the authors of [19] explored the meson phenomenology of large Nc QCD together with U(1)A chiral symmetry breaking. They showed that the chiral condensate scales as 1/mq for large mq. A remarkable observation made in [19] is that in addition to the confinement/deconfinement phase transition the model exhibits a possibility that another transition set by Tfund could happen in deconfined phase, T>Tdeconf, where Tdeconf=MKK/(2π). Since the value of MKK is around 1 GeV, we can estimate Tdeconf~160MeV. In this case for Tdeconf<T<Tfund there exist free unbound quarks and meson bound sates of heavy quarks and above Tfund the meson states dissociate into free quarks, which in some sense mimics the dissociation of heavy quarkonium in quark-gluon plasma (QGP). However, we should note that meson bound states in Dp/Dq systems are deeply bound, while the heavy quarkonia in QCD are shallow bound states. In this sense the bound state that disappears above Tfund could be that of strange quarks rather than charmonium or bottomonium [20].To attain a realistic gravity dual description of (largeNc) QCD, non-Abelian chiral symmetry is an essential ingredient together with confinement. Holographic QCD models, which are equipped with the correct structure for the problem, namely, chiral symmetry and confinement, have been suggested in top-down and bottom-up approaches. They are found to be rather successful for various hadronic observables and for certain processes dominated by large Nc. Based on a D4/D8/D8¯ model, Sakai and Sugimoto studied hadron phenomenology in the chiral limit mq=0, and the chiral symmetry breaking geometrically [21, 22]. More phenomenological holographic QCD models were proposed [23–25]. In [23, 24], chiral symmetry breaking is realized by a nonzero chiral condensate whose value is fitted to meson data from experiments. Hadronic spectra and light-front wave functions were studied in [26] based on the “Light-Front Holography” which maps amplitudes in extra dimension to a Lorentz invariant impact separation variable ζ in Minkowski space at fixed light-front time. Light-Front Holography has led to many successful applications in hadron physics including light-quark hadron spectra, meson and baryon form factors, the nonperturbative QCD coupling, and light-front wave-functions; see [27–29] for a review on this topic. In [30], a relation between a bottom-up holographic QCD model and QCD sum rules was analyzed.Now, we demonstrate how to construct a bottom-up holographic QCD model by looking at a low-energy QCD. For illustration purposes, we compare our approach with the (gauged) linear sigma model. The D3/D7 model is summarized in AppendixB with some calculational details. For a review of the linear sigma model, we refer to [31]. Some material in this section is taken from [32]. Suppose  that we are interested in two-flavor QCD at low energy, roughly below 1 GeV. In this regime usually we resort to the effective models or theories of QCD for analytic studies since the QCD lagrangian does not help much.To construct the holographic QCD model dual to two flavor low-energy QCD with chiral symmetry, we first choose relevant fields. To do this, we consider composites of quark fields that have the same quantum numbers with the hadrons of interest. For instance, in the linear sigma model we introduce pion-like and sigma-like fields:π⃗~q¯τ⃗γ5q and σ~q¯q, where τ⃗ is the Pauli matrix for isospin. In the AdS/CFT dictionary, this procedure may be dubbed operator/field correspondence: one-to-one mapping between gauge-invariant local operators in gauge theory and bulk fields in gravity sides. Then we introduce (3.1)q¯LγμtaqL⟷ALμa(x,z),q¯RγμtaqR⟷ARμa(x,z),q¯RαqLβ⟷(2z)Xαβ(x,z).
An interesting point here is that the 5D mass of the bulk field is not a free parameter of the model. This bulk mass is determined by the dimension Δ and spin p of the dual 4D operator in AdSd+1. For instance, consider a bulk field X(x,z) dual to q¯(x)q(x). The bulk mass of X(x,z) is given by mX2=(Δ-p)(Δ+p-d) with Δ=3, p=0, and d=4, and so mX2=-3. For more details, see Appendix A.To write down the Lagrangian of the linear sigma model, we consider (global) chiral symmetry of QCD. Since the mass of light quark~10 MeV is negligible compared to the QCD scale ΛQCD~200 MeV, we may consider the exact chiral symmetry of QCD and treat quark mass effect in a perturbative way. Under the axial transformation, q→e-iγ5τ⃗·θ⃗/2q, the pion-like and sigma-like states transform as π⃗→π⃗+θ⃗σ and σ→σ-θ⃗·π⃗. From this, we can obtain terms that respect chiral symmetry such as π⃗2+σ2. Similarly we ask the holographic QCD model to respect chiral symmetry of QCD. In AdS/CFT, however, a global symmetry in gauge theory corresponds local symmetry in the bulk, and therefore the corresponding holographic QCD model should posses local chiral symmetry. This way vector and axial-vector fields naturally fit into chiral Lagrangian in the bulk as the gauge boson of the local chiral symmetry.We keep the chiral symmetry in the Lagrangian since it will be spontaneously broken. Then we should ask how to realize the spontaneous chiral symmetry breaking. In the linear sigma model, we have a potential term like((π⃗2+σ2)-c2)2 that leads to spontaneous chiral symmetry breaking due to a nonzero vacuum expectation value of the scalar field σ, 〈σ〉=c. In this case the explicit chiral symmetry due to the small quark mass could be mimicked by adding a term -ϵσ to the potential which induces a finite mass of the pion, mπ2~ϵ/c. In a holographic QCD model, the chiral symmetry breaking is encoded in the vacuum expectation value of a bulk scalar field dual to q¯q. For instance, in the hard wall model [23, 24], it is given by 〈X〉=mqz+ζz3, where mq and ζ are proportional to the quark mass and the chiral condensate in QCD. In the D3/D7 model, chiral symmetry breaking can be realized by the embedding solution as shown in Appendix B.The last step to get to the gravity dual to two flavor low-energy QCD is to ensure the confinement to have discrete spectra for hadrons. The simplest way to realize it might be to truncate the extra dimension atz=zm such that the radial direction z of dual gravity runs from zero to zm. Since the radial direction corresponds to an energy scale of a boundary gauge theory, 1/zm maps to ΛQCD.Putting things together, we could arrive at the following bulk Lagrangian with localSU(2)L×SU(2)R, the hard wall model [23, 24]:(3.2)SHW=∫d4x∫dzgTr⁡[-14g52(FL2+FR2)+|DX|2+3|X|2],
where DμX=∂μX-iALμX+iXARμ and AL,R=AL,Rata with Tr(tatb)=(1/2)δab. The bulk scalar field is defined by X=X0e2iπata, where X0≡〈X〉. Here g5 is the five-dimensional gauge coupling, g52=12π2/Nc. The background is given by(3.3)ds2=1z2(dt2-dx⃗2-dz2),0≤z≤zm.
Instead of the sharp IR cutoff in the hard wall mode, we may introduce a bulk potential that plays a role of a smooth cutoff. In [33], this smooth cutoff is introduced by a factor e-Φ with Φ(z)=z2 in the bulk action, the soft wall model. The form Φ(z)=z2 in the AdS would ensure the Regge-like behavior of the mass spectrum mn2~n. The action is given by(3.4)SSW=∫d4x∫dze-ΦgTr⁡[-14g52(FL2+FR2)+|DX|2+3|X|2].
Here we briefly show how to obtain the 4D vector meson mass in the soft wall model. The vector field is defined by V=AL+AR. With the Kaluza-Klein decomposition Vμa(x,z)=g5∑nvn(z)ρμa(x), we obtain (3.5)∂z(e-B∂zvn)+mn2e-Bvn=0,
where B=Φ(z)-A(z)=z2+logz in the AdS geometry (3.3). With vn=eBψn, we transform the equation of motion into the form of a Schrödinger equation: (3.6)ψn′′-V(z)ψn=-mn2ψn,
where V(z)=z2+3/(4z2). Here mn is the mass of the vector resonances, and ρ meson corresponds to n=0. The solution is well known in quantum mechanics and the eigenvalue mn2 is given by [33] (3.7)mn2=4c(n+1),
where c is introduced to restore the energy dimension.The finite temperature could be neatly introduced by a black hole inAdSd+1, where d is the dimension of the boundary gauge theory. The background is given by(3.8)ds2=1z2(f(z)dt2-dx⃗2-dz2f(z)),
where f(z)=1-zd/zhd. The temperature of the boundary gauge theory is identified with the Hawking temperature of the black hole T=d/(4πzh). In Appendices D and E, we try to explain in a comprehensive manner how to calculate the Hawking temperature of a black hole.Now we move on to dense matter. According to the AdS/CFT dictionary, a chemical potential in boundary gauge theory is encoded in the boundary value of the time component of the bulkU(1) gauge field. To be more specific on this, we first consider the chemical potential term in gauge theory:(3.9)Lμ=μqq†q.
Then, we introduce a bulk U(1) gauge field Aμ which is dual to q¯γμq. According to the dictionary, A0(z→0)~c1zd-Δ-p+c2zΔ-p, we have A0(z→0)~μq. In the hard wall model, the solution of the bulk U(1) vector field is given by(3.10)At(z)=μ+ρz2,
where μ and ρ are related to quark chemical potential and quark (or baryon) number density in boundary gauge theory. It is interesting to notice that in chiral perturbation theory, a chemical potential is introduced as the time component of a gauge field by promoting the global chiral symmetry to a local gauge one [34, 35].


## 4. Vacuum Structures

At low energy or momentum scales roughly smaller than 1 GeV,r>1 fm, QCD exhibits confinement and a nontrivial vacuum structure with condensates of quarks and gluons. In this section, we discuss the gluon condensate and quark-gluon mixed condensate.The gluon condensate〈GμνaGaμν〉 was first introduced, at zero temperature, in [36] as a measure for nonperturbative physics in QCD. The gluon condensate characterizes the scale symmetry breaking of massless QCD at quantum level. Under the infinitesimal scale transformation (4.1)xμ′=(1+δλ)xμ,Aμ′=(1-δλ)Aμ,q′=(1-32δλ)q,
the trace of the energy momentum tensor reads schematically (4.2)∂μJDμ=Tμμ≃-〈αsπGμνaGaμναsπ〉.
Here JDμ is the dilatation current, αs is the gauge coupling, and Tμν is the energy-momentum tensor of QCD. Due to Lorentz invariance, we can write 〈Tμν〉=-ϵvacημν, where ϵvac is the energy of the QCD vacuum. Therefore, the value of the gluon condensate sets the scale of the QCD vacuum energy. In addition, the gluon condensate is important in the QCD sum rule analysis since it enters in the operator product expansion (OPE) of the hadronic correlators [36]. At high temperature, the gluon condensate is useful to study the nonperturbative nature of the QGP. For instance, lattice QCD results on the gluon condensate at finite temperature [37, 38] indicate that the value of the gluon condensate shows a drastic change around Tc regardless of the number of quark flavors. The change in the gluon condensate could lead to a dropping of the heavy quarkonium mass around Tc [39].In holographic QCD, the gluon condensate figures in a dilaton profile according to the AdS/CFT since the dilaton is dual to the scalar gluon operatorTr(GμνGμν). The 5D gravity action with the dilaton is given by (4.3)S=γ12κ2∫d5xg[R+12R2-12∂Mϕ∂Mϕ],
where γ=+1 for Minkowski metric, and γ=-1 for Euclidean signature. We work with Minkowski metric for most cases in this paper. The solution of this system is discovered in [40, 41] by solving the coupled dilaton equation of motion and the Einstein equation: (4.4)ds2=(Rz)2(1-c2z8ημνdxμdxν+dz2),
and the corresponding dilaton profile is given by (4.5)ϕ(z)=32log⁡(1+cz41-cz4)+ϕ0,
where ϕ0 is a constant. At z=1/c1/4 there exists a naked singularity that might be resolved in a full string theory consideration. Near the boundary z→0, (4.6)ϕ(z)~cz4.
Therefore, c is nothing but the gluon condensate up to a constant. Unfortunately, however, c is an integration constant of the coupled dilaton equation of motion and the Einstein equation, and therefore, it will be determined by matching with physical observables. In [41], the value of the gluon condensate is estimated by the glueball mass. An interesting idea based on the circular Wilson loop calculation in gravity side is proposed to calculate the value of the gluon condensate G2≡〈(αs/π)GμνaGμνa〉 [42]. The value is determined to be G2=0.010±0.0023 GeV at zero temperature [42]. A phenomenological estimation of the gluon condensate in QCD sum rules gives 〈(αs/π)GμνaGaμν〉≃0.012 GeV4 [36].Now we consider quark-gluon-mixed condensate〈q¯σμνGμνq〉, which can be regarded as an additional order parameter for the spontaneous chiral symmetry breaking since the quark chirality flips via the quark-gluon operator. Thus, it is naturally expressed in terms of the quark condensate as (4.7)〈q¯σμνGμνq〉=m02〈q¯q〉.
In [43], an extended hard wall model is proposed to calculate the value of m02. The bulk action of the extended model is given by(4.8)S=∫d5xgTr⁡[|DX|2+3|X|2-14g52(FL2+FR2)+|DΦ|2-5Φ2],
where Φ is a bulk scalar field dual to the 4D operator on the left-hand-side of (4.7). Then the chiral condensate and the mixed condensate are encoded in the vacuum expectation value of the two scalar fields:(4.9)〈X(x,z)〉=12(m̂z+σz3),〈Φ(x,z)〉=16(c1z-1+σMz5),
where c1 is the source term for the mixed condensate and σM represents the mixed condensate σM=〈q¯RσμνGμνqL〉. Taking c1=0, a source-free condition to study only spontaneous symmetry breaking, we determine the value of the mixed condensate or m02 by considering various hadronic observables. In this sense the mixed condensate is not calculated but fitted to experimental data like the chiral condensate in the hard wall model. The favored value of the m02 in [43] is 0.72 GeV2. A new method to estimate the value of m02 in is suggested in [44], where a nonperturbative gauge invariant correlator (the nonlocal condensate) is calculated in dual gravity description to obtain m02. With inputs from the slopes for the Regge trajectory of vector mesons and the linear term of the Cornell potential, they obtained m02=0.70 GeV2, which is comparable to that from the QCD sum rules, 0.8 GeV2 [45].


## 5. Spectroscopy and Form Factors

Any newly proposed models or theories in physics are bound to confront experimental data, for instance, hadron masses, decay constants, and form factors. In this section, we consider the spectroscopy of the glueball, light meson, heavy quarkonium, and hadron form factors in hard wall model, soft wall model, and their variants.


### 5.1. Glueballs

Glueballs are made up of gluons with no constituent quarks in them. The glueball states are in general mixed with conventionalq¯q states; so in experiments we may observe these mixed states only. Their existence was expected from the early days of QCD [46, 47]. For theoretical and experimental status of glueballs, we refer to [48, 49].The spectrum of glueballs is one of the earliest QCD quantities calculated based on the AdS/CFT duality. In [12], Witten confirmed the existence of the mass gap in the dilaton equation of motion on a black hole background, implying a discrete glueball spectrum with a finite gap. Extensive studies on the glueball spectrum were done in [50, 51] and also comparisons between the supergravity results and lattice gauge theory results were made.Now we consider a scalar glueball (0++) on R3×S1 as an example [50]. When the radius of the circle S1 is very small R→0, only the gauge degrees of freedom remain and the gauge theory is effectively the same as pure QCD3 [12, 50]. Using the operator/field correspondence, we first find operators that have the quantum numbers with glueball states of interest and then introduce a corresponding bulk field to obtain the glueball masses. In this case we are to solve an equation of motion for a bulk scalar field ϕ, which is dual to tr F2 in the AdS5 Euclidean black hole background. The equation of motion for ϕ is given by (5.1)∂μ(g∂νϕgμν)=0,
and the metric is (5.2)ds2=(ρ2-b4ρ2)-1dρ2+(ρ2-b4ρ2)dτ2+ρ2+(dx⃗)2+dΩ52,
where τ is for the compactified imaginary time direction. For simplicity, we assume that ϕ is independent of τ [12, 50] and seek a solution of the form ϕ(ρ,x)=f(ρ)ek·x, where k⃗ is the momentum in R3. Then the equation of motion for f(ρ) reads (5.3)ρ-1ddρ((ρ4-b4)ρdfdρ)+m2=0,
where m2 is the three-dimensional glueball mass, m2=-k2 [12, 50]. By solving this eigenvalue equation with suitable boundary conditions, regularity at the horizon (ρ=b), and normalizability f~ρ-4 at the boundary (ρ→∞), we can obtain discrete eigenvalues, the three-dimensional glueball masses. In the context of a sliced AdS background of the Polchinski and Strassler set up [16], which is dual to confining gauge theory, the mass ratios of glueballs are studied in [52, 53].More realistic or phenomenology-oriented approaches follow the earlier developments. In the soft wall model the mass spectra of scalar and vector glueballs and their dependence on the bulk geometry and the shape of the soft wall are studied in [54]. The exact glueball correlators are calculated in [55], where the decay constants as well as the mass spectrum of the glueball are also obtained in both hard wall and soft wall models. Here we briefly summarize the scalar glueball properties in the soft wall model [54, 55]. Following a standard path to construct a bottom-up mode, we introduce a massless bulk scalar field ϕ dual to the scalar gluon operator Tr(FμνFμν) to write down the bulk action as [54] (5.4)S~∫d5ge-ΦgMN(∂Mϕ)(∂Nϕ),
where Φ=z2 as in the soft wall model. The equation of motion for ϕ(q,z) can be transformed to a one-dimensional Schrödinger form: (5.5)ψ′′-V(z)ψ=q2ψ,
where ψ=e-(Φ+3lnz)/2ϕ with q2=m2 [54]. The glueball mass spectrum is then given as the eigenvalue of the Schrödinger type equation with regular eigenfunction at z=0 and z=∞: (5.6)mn2=4(n+2)c̃,
where n is an integer, n=0,1,2,…. c̃ is introduced to make the exponent Φ dimensionless, Φ=c̃z2, and it will be fit to hadronic data. Since the vector meson mass in the soft wall model is mn2=4(n+1)c̃, we calculate the ratio of the lightest (n=0) scalar glueball mass mG02 to the ρ meson mass to obtain mG02/mρ2=2 [54]. The properties of the glueball at finite temperature are studied in the hard wall model [56] and also in the soft wall model [56, 57] by calculating the spectral function of the glueball in the AdS black hole backgroudn. The spectral function is related to various Green functions, and it can be defined by the two-point retarded Green function as ρ(ω,q⃗)=-2ImGR(ω,q⃗). The retarded function can be computed in the real-time AdS/CFT, following the prescription proposed in [58]. Both studies using the soft wall model predicted that the dissociation temperature of scalar glueballs is far below the deconfinement (Hawking-Page) transition temperature of the soft wall model. See Section 6.1 and Appendix F for more on the Hawking-Page transition. Note that below the Hawking-Page transition temperature, the AdS black hole is unstable. In [56], the melting temperature of the scalar glueball from the spectral functions is about 40–60 MeV, while the deconfinement temperature of the soft wall model is about 190 MeV [59]. This implies that we have to build a more refined holographic QCD model to have a realistic melting temperature [56, 57].


### 5.2. Light Mesons

There have been an armful of works in holographic QCD that studied light meson spectroscopy. Here we will try to summarize results from the hard wall model, soft wall model, and their variants.In Table1, we list some hadronic observables from hard wall models to see if the results are stable against some deformation of the model. In the table, * means input data and the model with no * is a fit to all seven observables: Model A and Model B from the hard wall model [23], Model I from a hard wall model in a deformed AdS geometry [60], and Model II from a hard wall model with the quark-gluon-mixed condensate [43]. In [24], the following deformed AdS background is considered:(5.7)ds2=π2zmsin⁡[πz/(2zm)](dt2-dxidxi-dz2),0≤z≤zm,
and it is stated that the correction from the deformation is less than 10%. The backreaction on the AdS metric due to quark mass and chiral condensate is investigated in [60]. One of the deformed backgrounds obtained in [60] phenomenologically reads(5.8)ds2=1z2e-2B(z)(dt2-dxidxi-dz2),0≤z≤zm,
where B(z)=(mq2/24)z2+(mqσ/16)z4+(σ2/24)z6. In Table 1, we quote some results from this deformed background. Dynamical (back-reacted) holographic QCD model with area-law confinement and linear Regge trajectories was developed in [61].Table 1
Meson spectroscopy from the hard-wall model and from its variations: Model I [60], Model II [43], Model A [23],  and Model B [23]. The experimental data listed in the last column are taken from the particle data group [62]. All results are given in units of MeV except for the condensate and the ratio of two condensates.
Model IModel IIModel AModel BExperimentmρ775.8*775.8775.8*832775.49±0.34ma113481244136312201230±40fπ92.4*80.592.4*84.092.4±0.35Fρ1/2334330329353345±8Fa11/2481459486440433±13mπ139.6*139.3139.6*141139.57±0.00035gρππ4.464.874.485.296.03±0.07We remark that the sensitivity of calculated hadronic observables to the details of the hard wall model was studied in [63] by varying the infrared boundary conditions, the 5D gauge coupling, and scaling dimension of q¯q operator. It turns out that predicted hadronic observables are not sensitive to varying scaling dimension of q¯q operator, while they are rather sensitive to the IR boundary conditions and the 5D gauge coupling [63].In addition to mesons, baryons were also studied in the hard wall model [64–67]. It is pointed out in [65, 66] that one has to use the same IR cutoff of the hard wall model zm for both meson and baryon sectors.Now we collect some results from the soft wall model [33]. There were two nontrivial issues to be resolved in the original soft wall model. Firstly, so called, the dilaton factor Φ~z2 is introduced phenomenologically to explain mn2~n. The dilaton factor is supposed to be a solution of gravity-dilaton equations of motion. Secondly, the chiral symmetry breaking in the model is a bit different from QCD since the chiral condensate is proportional to the quark mass in the soft wall model. In QCD, in the chiral limit, where the quark mass is zero, the chiral condensate is finite that characterizes spontaneous chiral symmetry breaking. Several attempts have made to improve these aspects and to fit experimental values better [68–71]. In [69], a quartic term in the potential for the bulk scalar X dual to q¯q is introduced to the soft wall model to incorporate chiral symmetry breaking with independent sources for spontaneous and explicit breaking; thereby the chiral condensate remains finite in the chiral limit. Then, the authors of [69] parameterized the vev of the bulk scalar X0 such that it satisfies constraints from the AdS/CFT at UV and from phenomenology at IR: X0~mqz+σz3 as z→0 and X0~z as z→∞. The constraint at IR is due to the observation [72] that chiral symmetry is not restored in the highly excited mesons. Note that X0~z keeps the mass difference between vector and axial-vector mesons constant as z→∞. With the parameterized X0, they obtained a dilaton factor Φ(z) [69]. We list some of results of [69] in Table 2. An extended soft wall model with a finite UV cutoff was discussed in [73, 74]. In [75], the authors studied a dominant tetra-quark component of the lightest scalar mesons in the soft wall model, where a rather generic lower bound on the tetra-quark mass was derived.Table 2
Meson spectroscopy from the modified soft wall model [69]. We show the center values of experimental data. In [69] the experimental data are mostly taken from the particle data group [62], while ρ(1282) is from [76]. All results are given in units of MeV.
nρ-mesonρ experimenta1-mesona1 experiment1475775.51185123021129128215911647314291465190019304167417202101209651884190922792270620722149——722432265——As long as confinement and non-Abelian chiral symmetry are concerned, the Sakai-Sugimoto model [21, 22] based on a D4/D8/D8¯ brane configuration (see Appendix C) is the only available stringy model. In this model, properties of light mesons and baryons have been greatly studied [21, 22, 77–84].In a simple bottom-up model with the Chern-Simons term, it was also shown that baryons arise as stable solitons which are the 5D analogs of 4D skyrmions and the properties of the baryons are studied [85].


### 5.3. Heavy Quarkonium

The properties of heavy quark system both at zero and at finite temperature have been the subject of intense investigation for many years. This is so because, at zero temperature, the charmonium spectrum reflects detailed information about confinement and interquark potentials in QCD. At finite temperature, due to the small interaction cross-section of the charmonium in hadronic matter, the charmonium spectrum is expected to carry information about the early hot and dense stages of relativistic heavy ion collisions. In addition, the charmonium states may remain bound even above the critical temperatureTc. This suggests that analyzing the charmonium data from heavy ion collision inevitably requires more detailed information about the properties of charmonium states in QGP. Therefore, it is very important to develop a consistent nonperturbative QCD picture for the heavy quark system both below and above the phase transition temperature. For a recent review on heavy quarkonium see, for example, [86].Now we start with the hard wall model to discuss the heavy quarkonium in a bottom-up approach. A simple way to deal with the heavy quarkonium in the hard wall model was proposed in [87]. Since the typical energy scales involved for light mesons and heavy quarkonia are quite different, we may introduce an IR cutoffs zmH for heavy quarkonia in the hard wall model which is different from the IR cutoff for light mesons, 1/zmL~300 MeV. Note that in the hard wall model there is a one-to-one correspondence between the IR cutoff and the vector meson mass 1/zm~mV. In [87], the lowest vector cc¯ (J/ψ) mass ~3GeV is used as an input to fix the IR cutoff for the charmonium, 1/zmH≃1.32GeV. With this, the mass of the second resonance is predicted to be ~7.2GeV, which is quite different from the experiment mψ′~3.7GeV. This is in a sense generic limitation of the hard wall model whose predicted higher resonances are quite different from experiments. Moreover, having two different IR cutoffs in the hard wall model may cause a problem when we treat light quark and heavy quark systems at the same time. In the soft wall model, the mass spectrum of the vector meson is given by [33] (5.9)mn2=4(n+1)c.
For charmonium system, again the lowest mode (J/ψ) is used to fix c, c≃1.55GeV. Then the mass of the second resonance ψ′ is mψ′≃4.38GeV, which is 20% away from the experimental value of 3.686GeV [87]. Additionally, the mass of heavy quarkonium such as J/ψ at finite temperature is calculated to predict that the mass decreases suddenly at Tc and above Tc it increases with temperature. Furthermore, the dissociation temperature is determined to be around 494MeV in the soft wall model [87].To compare heavy quarkonium properties obtained in a holographic QCD study with lattice QCD, the finite-temperature spectral function in the vector channel within the soft wall model was explored in [88]. The spectral function is related to the two-point retarded Green function by ρ(ω,q⃗)=-2ImGR(ω,q⃗). The retarded function can be computed following the prescription [58]. Thermal spectral functions in a stringy set-up, D3/D7 model, were extensively studied in [89]. To deal with the heavy quarkonium in the soft wall model, two different scales (cρ and cJ/ψ) are introduced. It is observed in [88] that a peak in the spectral function melts with increasing temperature and eventually is flattened at T≃1.2Tc. It is also shown numerically that the mass shift squared is approximately proportional to the width broadening [88]. Another interesting finding in [88] is that the spectral peak diminishes at high momentum, which could be interpreted as the J/ψ suppression under the hot wind [90, 91]. A generalized soft wall mode of charmonium is constructed by considering not only the masses but also the decay constants of the charmonium, J/ψ and ψ′ [92]. They calculated the spectral function as well as the position of the complex singularities (quasinormal frequencies) of the retarded correlator of the charm current at finite temperatures. A predicted dissociation temperature is T≈540 MeV, or 2.8Tc [92].Alternatively, heavy quarkonium properties can be studied in terms of holographic heavy-quark potentials. Since the mass of heavy quarks is much larger than the QCD scale parameterΛQCD~200 MeV, the nonrelativistic Schrödinger equation could be a useful tool to study heavy quark bound states:  (5.10)(-▿22mr+V(r))Ψ(r)=EΨ(r),
where mr is the reduced mass, mr=mQ/2. A tricky point with potential models for quarkonia is which potential is to be used in the Schrödinger equation: the free energy or the internal energy. In the context of the AdS/CFT, there have been a lot of works on holographic heavy quark potentials [93–103]. Hou and Ren calculated the dissociation temperature of heavy quarkonia by solving the Schrödinger equation with holographic potentials [99]. They used two ansätze of the potential model: the F-ansatz (U-ansatz) which identifies the potential in the Schrödinger equation with the free energy (the internal energy), respectively. With the F-ansatz, J/ψ does not survive above Tc, while the dissociation temperature of Υ is (1.3–2.1)Tc. For the U-ansatz, J/ψ dissolves into open charm quarks around (1.2–1.7)Tc and Υ dissociates at about (2.5–4.2)Tc.We finish this subsection with a summary of the discussion in [20] on the usefulness of Dq/Dp systems in studying heavy quark bound states. A Dq/Dp system may be good for ss¯ bound states at high temperature since the mesons in the Dq/Dp system are deeply bounded, while heavy quarkonia are shallow bound states. However, there exist certain properties of heavy quarkonia in the quark-gluon plasma that could be understood in the D4/D6 model such as dissociation temperature.


### 5.4. Form Factors

Form factors are a source of information about the internal structure of hadrons such as the distribution of charge. We take the pion electromagnetic form factor as an example. Consider a pion-electron scattering processπ±+e-→π±+e- through photon exchange. The cross section of this process measured in experiments is different from that of Mott scattering which is for the Coulomb scattering of an electron with a point charge. This deviation is parameterized into the pion form factor Fπ(q2), where q2 is given by the energy and momentum of the photon q2=ω2-q⃗2. If the pion is a structureless point particle, we have Fπ=1. The pion electromagnetic form factor is expressed by, with the use of Lorentz invariance, charge conjugation, and electromagnetic gauge invariance:(5.11)(p1+p2)μFπ(q2)=〈π(p2)|Jμ|π(p1)〉,
where q2=(p2-p1)2 and Jμ is the electromagnetic current, Jμ=∑fefq¯fγμqf. The pion charge radius is determined by (5.12)〈rπ2〉=6∂Fπ(q2)∂q2|q2=0.
In a vector meson dominance model, where the photon interacts with the pion only via vector mesons, especially ρ meson, the pion form factor is given by (5.13)Fπ(q2)=mρ2mρ2-q2-imρΓρ(q2).
Then we obtain the pion charge radius 〈rπ2〉=6/mρ≃0.63 fm. The experimental value is 〈rπ2〉=0.672 fm [104]. To evaluate the form factor, we consider the three-point correlation function of two axial vector currents which contains nonzero projection onto a one pion state and the external electromagnetic current: (5.14)Γμαβ(p1,p2)=-∫dx∫dye(-ip1x+ip2y)〈0|T{Jα5†(x)Jμ(0)Jβ5(y)}|0〉.
Alternatively, we can consider two pseudoscalar currents instead of the axial vector currents. The three-point correlation function can be decomposed into several independent Lorentz structures. Among them we pick up the Lorentz structure corresponding to the pion form factor: (5.15)〈0|Jβ5|p2〉〈p2|Jμ|p1〉〈p1|Jα5†|0〉≃fπ2Fπ(q2)p1αp2β(p1μ+p2μ).
Note that 〈0|Jα5|p〉=ifπpα, where |p〉 is a one pion state. For more details on the form factor, we refer to [105–107].In a holographic QCD approach, we can easily evaluate the three-point correlation function of two axial vector currents (or two pseudoscalar currents) and the external electromagnetic current. In [108], the form factors of vector mesons were calculated in the hard wall model and the electric charge radius of the ρ-meson was evaluated to be 〈rρ2〉=0.53 fm2. The number from the soft wall model is 〈rρ2〉=0.655 fm2 [109]. The approach based on the Dyson-Schwinger equations predicted 〈rρ2〉=0.37 fm2 [110] and 〈rρ2〉=0.54 fm2 [111]. The quark mass (or pion mass) dependence of the charge radius of the ρ-meson was calculated in lattice QCD: for instance, with mπ≃300 MeV, 〈rρ2〉=0.55 fm2 [112]. The pion form factor was studied in the hard wall model [113] and in a model that interpolates between the hard wall and soft wall models [114]. The results obtained are 〈rπ2〉=0.58 fm [113] and in [114] 〈rπ2〉=0.500 fm, 〈rπ2〉=0.576 fm, depending on their parameter choice. The gravitational form factors of mesons were calculated in the hard wall model [115, 116]. The gravitational form factor of the pion is defined by (5.16)〈πb(p′)|Θμν(0)|πa(p)〉=12δab[(gμνq2-qμqν)Θ1(q2)+4PμPνΘ2(q2)],
where Θμν is the energy momentum tensor, q=p′-p, and P=(p′+p)/2. There are also interesting works that studied various form factors in holographic QCD [117–120]. Form factors of vector and axial-vector mesons were calculated in the Sakai-Sugimoto model (Figure 4) [121].Figure 4
QCD phase diagram.


## 5.1. Glueballs

Glueballs are made up of gluons with no constituent quarks in them. The glueball states are in general mixed with conventionalq¯q states; so in experiments we may observe these mixed states only. Their existence was expected from the early days of QCD [46, 47]. For theoretical and experimental status of glueballs, we refer to [48, 49].The spectrum of glueballs is one of the earliest QCD quantities calculated based on the AdS/CFT duality. In [12], Witten confirmed the existence of the mass gap in the dilaton equation of motion on a black hole background, implying a discrete glueball spectrum with a finite gap. Extensive studies on the glueball spectrum were done in [50, 51] and also comparisons between the supergravity results and lattice gauge theory results were made.Now we consider a scalar glueball (0++) on R3×S1 as an example [50]. When the radius of the circle S1 is very small R→0, only the gauge degrees of freedom remain and the gauge theory is effectively the same as pure QCD3 [12, 50]. Using the operator/field correspondence, we first find operators that have the quantum numbers with glueball states of interest and then introduce a corresponding bulk field to obtain the glueball masses. In this case we are to solve an equation of motion for a bulk scalar field ϕ, which is dual to tr F2 in the AdS5 Euclidean black hole background. The equation of motion for ϕ is given by (5.1)∂μ(g∂νϕgμν)=0,
and the metric is (5.2)ds2=(ρ2-b4ρ2)-1dρ2+(ρ2-b4ρ2)dτ2+ρ2+(dx⃗)2+dΩ52,
where τ is for the compactified imaginary time direction. For simplicity, we assume that ϕ is independent of τ [12, 50] and seek a solution of the form ϕ(ρ,x)=f(ρ)ek·x, where k⃗ is the momentum in R3. Then the equation of motion for f(ρ) reads (5.3)ρ-1ddρ((ρ4-b4)ρdfdρ)+m2=0,
where m2 is the three-dimensional glueball mass, m2=-k2 [12, 50]. By solving this eigenvalue equation with suitable boundary conditions, regularity at the horizon (ρ=b), and normalizability f~ρ-4 at the boundary (ρ→∞), we can obtain discrete eigenvalues, the three-dimensional glueball masses. In the context of a sliced AdS background of the Polchinski and Strassler set up [16], which is dual to confining gauge theory, the mass ratios of glueballs are studied in [52, 53].More realistic or phenomenology-oriented approaches follow the earlier developments. In the soft wall model the mass spectra of scalar and vector glueballs and their dependence on the bulk geometry and the shape of the soft wall are studied in [54]. The exact glueball correlators are calculated in [55], where the decay constants as well as the mass spectrum of the glueball are also obtained in both hard wall and soft wall models. Here we briefly summarize the scalar glueball properties in the soft wall model [54, 55]. Following a standard path to construct a bottom-up mode, we introduce a massless bulk scalar field ϕ dual to the scalar gluon operator Tr(FμνFμν) to write down the bulk action as [54] (5.4)S~∫d5ge-ΦgMN(∂Mϕ)(∂Nϕ),
where Φ=z2 as in the soft wall model. The equation of motion for ϕ(q,z) can be transformed to a one-dimensional Schrödinger form: (5.5)ψ′′-V(z)ψ=q2ψ,
where ψ=e-(Φ+3lnz)/2ϕ with q2=m2 [54]. The glueball mass spectrum is then given as the eigenvalue of the Schrödinger type equation with regular eigenfunction at z=0 and z=∞: (5.6)mn2=4(n+2)c̃,
where n is an integer, n=0,1,2,…. c̃ is introduced to make the exponent Φ dimensionless, Φ=c̃z2, and it will be fit to hadronic data. Since the vector meson mass in the soft wall model is mn2=4(n+1)c̃, we calculate the ratio of the lightest (n=0) scalar glueball mass mG02 to the ρ meson mass to obtain mG02/mρ2=2 [54]. The properties of the glueball at finite temperature are studied in the hard wall model [56] and also in the soft wall model [56, 57] by calculating the spectral function of the glueball in the AdS black hole backgroudn. The spectral function is related to various Green functions, and it can be defined by the two-point retarded Green function as ρ(ω,q⃗)=-2ImGR(ω,q⃗). The retarded function can be computed in the real-time AdS/CFT, following the prescription proposed in [58]. Both studies using the soft wall model predicted that the dissociation temperature of scalar glueballs is far below the deconfinement (Hawking-Page) transition temperature of the soft wall model. See Section 6.1 and Appendix F for more on the Hawking-Page transition. Note that below the Hawking-Page transition temperature, the AdS black hole is unstable. In [56], the melting temperature of the scalar glueball from the spectral functions is about 40–60 MeV, while the deconfinement temperature of the soft wall model is about 190 MeV [59]. This implies that we have to build a more refined holographic QCD model to have a realistic melting temperature [56, 57].


## 5.2. Light Mesons

There have been an armful of works in holographic QCD that studied light meson spectroscopy. Here we will try to summarize results from the hard wall model, soft wall model, and their variants.In Table1, we list some hadronic observables from hard wall models to see if the results are stable against some deformation of the model. In the table, * means input data and the model with no * is a fit to all seven observables: Model A and Model B from the hard wall model [23], Model I from a hard wall model in a deformed AdS geometry [60], and Model II from a hard wall model with the quark-gluon-mixed condensate [43]. In [24], the following deformed AdS background is considered:(5.7)ds2=π2zmsin⁡[πz/(2zm)](dt2-dxidxi-dz2),0≤z≤zm,
and it is stated that the correction from the deformation is less than 10%. The backreaction on the AdS metric due to quark mass and chiral condensate is investigated in [60]. One of the deformed backgrounds obtained in [60] phenomenologically reads(5.8)ds2=1z2e-2B(z)(dt2-dxidxi-dz2),0≤z≤zm,
where B(z)=(mq2/24)z2+(mqσ/16)z4+(σ2/24)z6. In Table 1, we quote some results from this deformed background. Dynamical (back-reacted) holographic QCD model with area-law confinement and linear Regge trajectories was developed in [61].Table 1
Meson spectroscopy from the hard-wall model and from its variations: Model I [60], Model II [43], Model A [23],  and Model B [23]. The experimental data listed in the last column are taken from the particle data group [62]. All results are given in units of MeV except for the condensate and the ratio of two condensates.
Model IModel IIModel AModel BExperimentmρ775.8*775.8775.8*832775.49±0.34ma113481244136312201230±40fπ92.4*80.592.4*84.092.4±0.35Fρ1/2334330329353345±8Fa11/2481459486440433±13mπ139.6*139.3139.6*141139.57±0.00035gρππ4.464.874.485.296.03±0.07We remark that the sensitivity of calculated hadronic observables to the details of the hard wall model was studied in [63] by varying the infrared boundary conditions, the 5D gauge coupling, and scaling dimension of q¯q operator. It turns out that predicted hadronic observables are not sensitive to varying scaling dimension of q¯q operator, while they are rather sensitive to the IR boundary conditions and the 5D gauge coupling [63].In addition to mesons, baryons were also studied in the hard wall model [64–67]. It is pointed out in [65, 66] that one has to use the same IR cutoff of the hard wall model zm for both meson and baryon sectors.Now we collect some results from the soft wall model [33]. There were two nontrivial issues to be resolved in the original soft wall model. Firstly, so called, the dilaton factor Φ~z2 is introduced phenomenologically to explain mn2~n. The dilaton factor is supposed to be a solution of gravity-dilaton equations of motion. Secondly, the chiral symmetry breaking in the model is a bit different from QCD since the chiral condensate is proportional to the quark mass in the soft wall model. In QCD, in the chiral limit, where the quark mass is zero, the chiral condensate is finite that characterizes spontaneous chiral symmetry breaking. Several attempts have made to improve these aspects and to fit experimental values better [68–71]. In [69], a quartic term in the potential for the bulk scalar X dual to q¯q is introduced to the soft wall model to incorporate chiral symmetry breaking with independent sources for spontaneous and explicit breaking; thereby the chiral condensate remains finite in the chiral limit. Then, the authors of [69] parameterized the vev of the bulk scalar X0 such that it satisfies constraints from the AdS/CFT at UV and from phenomenology at IR: X0~mqz+σz3 as z→0 and X0~z as z→∞. The constraint at IR is due to the observation [72] that chiral symmetry is not restored in the highly excited mesons. Note that X0~z keeps the mass difference between vector and axial-vector mesons constant as z→∞. With the parameterized X0, they obtained a dilaton factor Φ(z) [69]. We list some of results of [69] in Table 2. An extended soft wall model with a finite UV cutoff was discussed in [73, 74]. In [75], the authors studied a dominant tetra-quark component of the lightest scalar mesons in the soft wall model, where a rather generic lower bound on the tetra-quark mass was derived.Table 2
Meson spectroscopy from the modified soft wall model [69]. We show the center values of experimental data. In [69] the experimental data are mostly taken from the particle data group [62], while ρ(1282) is from [76]. All results are given in units of MeV.
nρ-mesonρ experimenta1-mesona1 experiment1475775.51185123021129128215911647314291465190019304167417202101209651884190922792270620722149——722432265——As long as confinement and non-Abelian chiral symmetry are concerned, the Sakai-Sugimoto model [21, 22] based on a D4/D8/D8¯ brane configuration (see Appendix C) is the only available stringy model. In this model, properties of light mesons and baryons have been greatly studied [21, 22, 77–84].In a simple bottom-up model with the Chern-Simons term, it was also shown that baryons arise as stable solitons which are the 5D analogs of 4D skyrmions and the properties of the baryons are studied [85].


## 5.3. Heavy Quarkonium

The properties of heavy quark system both at zero and at finite temperature have been the subject of intense investigation for many years. This is so because, at zero temperature, the charmonium spectrum reflects detailed information about confinement and interquark potentials in QCD. At finite temperature, due to the small interaction cross-section of the charmonium in hadronic matter, the charmonium spectrum is expected to carry information about the early hot and dense stages of relativistic heavy ion collisions. In addition, the charmonium states may remain bound even above the critical temperatureTc. This suggests that analyzing the charmonium data from heavy ion collision inevitably requires more detailed information about the properties of charmonium states in QGP. Therefore, it is very important to develop a consistent nonperturbative QCD picture for the heavy quark system both below and above the phase transition temperature. For a recent review on heavy quarkonium see, for example, [86].Now we start with the hard wall model to discuss the heavy quarkonium in a bottom-up approach. A simple way to deal with the heavy quarkonium in the hard wall model was proposed in [87]. Since the typical energy scales involved for light mesons and heavy quarkonia are quite different, we may introduce an IR cutoffs zmH for heavy quarkonia in the hard wall model which is different from the IR cutoff for light mesons, 1/zmL~300 MeV. Note that in the hard wall model there is a one-to-one correspondence between the IR cutoff and the vector meson mass 1/zm~mV. In [87], the lowest vector cc¯ (J/ψ) mass ~3GeV is used as an input to fix the IR cutoff for the charmonium, 1/zmH≃1.32GeV. With this, the mass of the second resonance is predicted to be ~7.2GeV, which is quite different from the experiment mψ′~3.7GeV. This is in a sense generic limitation of the hard wall model whose predicted higher resonances are quite different from experiments. Moreover, having two different IR cutoffs in the hard wall model may cause a problem when we treat light quark and heavy quark systems at the same time. In the soft wall model, the mass spectrum of the vector meson is given by [33] (5.9)mn2=4(n+1)c.
For charmonium system, again the lowest mode (J/ψ) is used to fix c, c≃1.55GeV. Then the mass of the second resonance ψ′ is mψ′≃4.38GeV, which is 20% away from the experimental value of 3.686GeV [87]. Additionally, the mass of heavy quarkonium such as J/ψ at finite temperature is calculated to predict that the mass decreases suddenly at Tc and above Tc it increases with temperature. Furthermore, the dissociation temperature is determined to be around 494MeV in the soft wall model [87].To compare heavy quarkonium properties obtained in a holographic QCD study with lattice QCD, the finite-temperature spectral function in the vector channel within the soft wall model was explored in [88]. The spectral function is related to the two-point retarded Green function by ρ(ω,q⃗)=-2ImGR(ω,q⃗). The retarded function can be computed following the prescription [58]. Thermal spectral functions in a stringy set-up, D3/D7 model, were extensively studied in [89]. To deal with the heavy quarkonium in the soft wall model, two different scales (cρ and cJ/ψ) are introduced. It is observed in [88] that a peak in the spectral function melts with increasing temperature and eventually is flattened at T≃1.2Tc. It is also shown numerically that the mass shift squared is approximately proportional to the width broadening [88]. Another interesting finding in [88] is that the spectral peak diminishes at high momentum, which could be interpreted as the J/ψ suppression under the hot wind [90, 91]. A generalized soft wall mode of charmonium is constructed by considering not only the masses but also the decay constants of the charmonium, J/ψ and ψ′ [92]. They calculated the spectral function as well as the position of the complex singularities (quasinormal frequencies) of the retarded correlator of the charm current at finite temperatures. A predicted dissociation temperature is T≈540 MeV, or 2.8Tc [92].Alternatively, heavy quarkonium properties can be studied in terms of holographic heavy-quark potentials. Since the mass of heavy quarks is much larger than the QCD scale parameterΛQCD~200 MeV, the nonrelativistic Schrödinger equation could be a useful tool to study heavy quark bound states:  (5.10)(-▿22mr+V(r))Ψ(r)=EΨ(r),
where mr is the reduced mass, mr=mQ/2. A tricky point with potential models for quarkonia is which potential is to be used in the Schrödinger equation: the free energy or the internal energy. In the context of the AdS/CFT, there have been a lot of works on holographic heavy quark potentials [93–103]. Hou and Ren calculated the dissociation temperature of heavy quarkonia by solving the Schrödinger equation with holographic potentials [99]. They used two ansätze of the potential model: the F-ansatz (U-ansatz) which identifies the potential in the Schrödinger equation with the free energy (the internal energy), respectively. With the F-ansatz, J/ψ does not survive above Tc, while the dissociation temperature of Υ is (1.3–2.1)Tc. For the U-ansatz, J/ψ dissolves into open charm quarks around (1.2–1.7)Tc and Υ dissociates at about (2.5–4.2)Tc.We finish this subsection with a summary of the discussion in [20] on the usefulness of Dq/Dp systems in studying heavy quark bound states. A Dq/Dp system may be good for ss¯ bound states at high temperature since the mesons in the Dq/Dp system are deeply bounded, while heavy quarkonia are shallow bound states. However, there exist certain properties of heavy quarkonia in the quark-gluon plasma that could be understood in the D4/D6 model such as dissociation temperature.


## 5.4. Form Factors

Form factors are a source of information about the internal structure of hadrons such as the distribution of charge. We take the pion electromagnetic form factor as an example. Consider a pion-electron scattering processπ±+e-→π±+e- through photon exchange. The cross section of this process measured in experiments is different from that of Mott scattering which is for the Coulomb scattering of an electron with a point charge. This deviation is parameterized into the pion form factor Fπ(q2), where q2 is given by the energy and momentum of the photon q2=ω2-q⃗2. If the pion is a structureless point particle, we have Fπ=1. The pion electromagnetic form factor is expressed by, with the use of Lorentz invariance, charge conjugation, and electromagnetic gauge invariance:(5.11)(p1+p2)μFπ(q2)=〈π(p2)|Jμ|π(p1)〉,
where q2=(p2-p1)2 and Jμ is the electromagnetic current, Jμ=∑fefq¯fγμqf. The pion charge radius is determined by (5.12)〈rπ2〉=6∂Fπ(q2)∂q2|q2=0.
In a vector meson dominance model, where the photon interacts with the pion only via vector mesons, especially ρ meson, the pion form factor is given by (5.13)Fπ(q2)=mρ2mρ2-q2-imρΓρ(q2).
Then we obtain the pion charge radius 〈rπ2〉=6/mρ≃0.63 fm. The experimental value is 〈rπ2〉=0.672 fm [104]. To evaluate the form factor, we consider the three-point correlation function of two axial vector currents which contains nonzero projection onto a one pion state and the external electromagnetic current: (5.14)Γμαβ(p1,p2)=-∫dx∫dye(-ip1x+ip2y)〈0|T{Jα5†(x)Jμ(0)Jβ5(y)}|0〉.
Alternatively, we can consider two pseudoscalar currents instead of the axial vector currents. The three-point correlation function can be decomposed into several independent Lorentz structures. Among them we pick up the Lorentz structure corresponding to the pion form factor: (5.15)〈0|Jβ5|p2〉〈p2|Jμ|p1〉〈p1|Jα5†|0〉≃fπ2Fπ(q2)p1αp2β(p1μ+p2μ).
Note that 〈0|Jα5|p〉=ifπpα, where |p〉 is a one pion state. For more details on the form factor, we refer to [105–107].In a holographic QCD approach, we can easily evaluate the three-point correlation function of two axial vector currents (or two pseudoscalar currents) and the external electromagnetic current. In [108], the form factors of vector mesons were calculated in the hard wall model and the electric charge radius of the ρ-meson was evaluated to be 〈rρ2〉=0.53 fm2. The number from the soft wall model is 〈rρ2〉=0.655 fm2 [109]. The approach based on the Dyson-Schwinger equations predicted 〈rρ2〉=0.37 fm2 [110] and 〈rρ2〉=0.54 fm2 [111]. The quark mass (or pion mass) dependence of the charge radius of the ρ-meson was calculated in lattice QCD: for instance, with mπ≃300 MeV, 〈rρ2〉=0.55 fm2 [112]. The pion form factor was studied in the hard wall model [113] and in a model that interpolates between the hard wall and soft wall models [114]. The results obtained are 〈rπ2〉=0.58 fm [113] and in [114] 〈rπ2〉=0.500 fm, 〈rπ2〉=0.576 fm, depending on their parameter choice. The gravitational form factors of mesons were calculated in the hard wall model [115, 116]. The gravitational form factor of the pion is defined by (5.16)〈πb(p′)|Θμν(0)|πa(p)〉=12δab[(gμνq2-qμqν)Θ1(q2)+4PμPνΘ2(q2)],
where Θμν is the energy momentum tensor, q=p′-p, and P=(p′+p)/2. There are also interesting works that studied various form factors in holographic QCD [117–120]. Form factors of vector and axial-vector mesons were calculated in the Sakai-Sugimoto model (Figure 4) [121].Figure 4
QCD phase diagram.


## 6. Phases of QCD

Understanding the QCD phase structure is one of the important problems in modern theoretical physics; see [122–126] for some recent reviews. However, a quantitative calculation of the phase diagram from the first principle is extraordinarily difficult.Basic order parameters for the QCD phase transitions are the Polyakov loop which characterizes the deconfinement transition in the limit of infinitely large quark mass and the chiral condensate for chiral symmetry in the limit of zero quark mass. The expectation value of the Polyakov loop is loosely given by(6.1)〈L〉≈lim⁡r→∞e-βV(r),
where V(r) is the potential between a static quark-antiquark pair at a distance r, and β~1/T. The expectation value of the Polyakov loop is zero in confined phase, and it is finite in deconfined phase, while the chiral condensate, which is the simplest order parameter for the chiral symmetry, is nonzero with broken chiral symmetry, vanishing with a restored chiral symmetry. Apart from these order parameters, there are thermodynamic quantities that are relevant to study the QCD phase transition. The equation of state is one of them. The energy density, for instance, has been found to rise rapidly at some critical temperature. This is usually interpreted as deconfinement: liberation of many new degrees of freedom. The fluctuations of conserved charges such as baryon number or electric charge [127–130] are also an important signal of the quark-hadron phase transition. The quark (or baryon) number susceptibility, which measures the response of QCD to a change of the quark chemical potential, is one of such fluctuations [127, 131].The nature of the chiral transition of QCD depends on the number of quark flavors and the value of the quark mass. For pureSU(3) gauge theory with no quarks, it is first order. In the case of two massless and one massive quarks, the transition is the second-order at zero or small quark chemical potentials, and it becomes the first order as we increase the chemical potential. The point where the second order transition becomes the first order is called tricritical point. With physical quark masses of up, down, and strange, the second order at zero or low chemical potential becomes the crossover, and the tricritical point turns into the critical end point.


### 6.1. Confinement/Deconfinement Transition

We first discuss the deconfinement transition. In holographic QCD, the confinement to deconfinement phase transition is described by the Hawking-Page transition [132], a phase transition between the Schwarzschild-AdS black hole and thermal AdS backgrounds. This identification was made in [12]. One simple reasoning for this identification is from the observation that the Polyakov expectation value is zero on the thermal AdS geometry, while it is finite on the AdS black hole. See Appendix F for some more description of the Hawking-Page transition and the Polyakov expectation in thermal AdS and AdS black hole. In low-temperature confined phase, thermal AdS, which is nothing but the AdS metric in Euclidean space, dominates the partition function, while at high temperature, AdS-black hole geometry does. This was first discovered in the finite volume boundary case in [12]. In the bottom-up model, it is shown that the same phenomena happen also for infinite boundary volume if there is a finite scale associated with the fifth direction [59].Here we briefly summarize the Hawking-Page analysis of [59] done in the hard wall model. In the Euclidean gravitational action given by(6.2)Sgrav=-12κ2∫d5xg(R+12L2),
where κ2=8πG5 and L is the length scale of the AdS5, there are two solutions for the equations of motion derived from the gravitational action. The one is the sliced thermal AdS (tAdS):
(6.3)ds2=L2z2(dτ2+dz2+dx⃗32),
where the radial coordinate runs from the boundary of tAdS space z=0 to the cut-off zm. Here τ is for the compactified Euclidean time-direction with periodicity β′. The other solution is the AdS black hole (AdSBH) with the horizon zh:(6.4)ds2=L2z2(f(z)dτ2+dz2f(z)+dx⃗32),
where f(z)=1-(z/zh)4. The Hawking temperature of the black hole solution is T=1/(πzh), which is given by regularizing the metric near the horizon. At the boundary z=ϵ the periodicity of the time-direction in both backgrounds is the same and so the time periodicity of the tAdS is given by (6.5)β=πzhf(ϵ).
Now we calculate the action density V, which is defined by the action divided by the common volume factor of R3. The regularized action density of the tAdS is given by(6.6)V1(ϵ)=4L3κ2∫0β′dτ∫ϵzIRdzz5,
and that of the AdSBH is given by(6.7)V2(ϵ)=4L3κ2∫0πzhdτ∫ϵz¯dzz5,
where z¯=min(zm,zh). Then, the difference of the regularized actions is given by(6.8)ΔVg=lim⁡ϵ→0[V2(ϵ)-V1(ϵ)]={L3πzhκ212zh4,zm<zh,L3πzhκ2(1zm4-12zh4),zm>zh.
When ΔVg is positive (negative), tAdS (AdSBH) is stable. Thus, at ΔVg=0 there exists a Hawking-Page transition. In the first case zm<zh, there is no Hawking-Page transition and the tAdS is always stable. In the second case zm>zh, the Hawking-Page transition occurs at(6.9)Tc=21/4(πzm),
and at low temperature T<Tc (at high temperature T>Tc) the thermal AdS (the AdS black hole) geometry becomes a dominant background. When we fix the IR cutoff by the ρ meson mass, we obtain 1/zm=323 MeV and Tc=122 MeV. In the soft wall model, Tc=191MeV [59].This work has been extended in various directions. The authors of [133] revisited the thermodynamics of the hard wall and soft wall model. They used holographic renormalization to compute the finite actions of the relevant supergravity backgrounds and verify the presence of a Hawking-page type phase transition. They also showed that the entropy, in the gauge theory side, jumps from N0 to N2 at the transition point [133]. In [134], the extension was done by studying the thermodynamics of AdS black holes with spherical or negative constant curvature horizon, dual to a non-supersymmetric Yang-Mills theory on a sphere or hyperboloid respectively. They also studied charged AdS black holes [135] in the grand canonical ensemble, corresponding to a Yang-Mills theory at finite chemical potential, and found that there is always a gap for the infrared cutoff due to the existence of a minimal horizon for the charged AdS black holes with any horizon topology [134]. With an assumption that the gluon condensate melts out at finite temperature, a Hawking-Page type transition between the dilaton AdS geometry in (4.4) and the usual AdS black hole has studied in [136].The effect of the number of quark flavorsNf and baryon number density on the critical temperature was investigated by considering a bulk meson action together with the gravity action in [137]. It is shown that the critical temperature decreases with increasing Nf. As the number density was raised, the critical temperature begins to drop, but it saturates to a constant value even at very large density. This is mostly due to the absence of the back-reaction from number density [137]. The back-reaction due to the number density has included in [138–140]. In [141], deconfinement transition of AdS/QCD with 𝒪(α′3) corrections was investigated. In [142], thermodynamics of the asymptotically-logarithmically-AdS black-hole solutions of 5D dilaton gravity with a monotonic dilaton potential are analyzed in great detail, where it is shown that in a special case, where the asymptotic geometry in the string frame reduces to flat space with a linear dilaton, the phase transition could be second order. The renormalized Polyakov loop in the deconfined phase of a pure SU(3) gauge theory was computed in [143] based on a soft wall metric model. The result obtained in this work is in good agreement with the one from lattice QCD simulations.Due to this Hawking-Page transition, we are not to use the black hole in the confined phase, and so we are not to obtain the temperature dependence of any hadronic observables. This is consistent with largeNc QCD at leading order. For instance, it was shown in [144, 145] that the Wilson loops, both time-like and space-like, and the chiral condensate are independent of the temperature in confining phase to leading order in 1/Nc. This means that the chiral and deconfinement transitions are first order. The deconfinement and chiral phase transitions of an SU(N) gauge theory at large Nc were also discussed in [146]. However, in reality we observe temperature dependence of hadronic quantities, and therefore we have to include large Nc corrections in holographic QCD in a consistent way. A quick fix-up for this might be to use the temperature dependent chiral condensate as an input in a holographic QCD model and study how this temperature dependence conveys into other hadronic quantities [147].


### 6.2. Chiral Transition

Now we turn to the chiral transition of QCD based on the chiral condensate. In the hard wall model, the chiral symmetry is broken, in a sense, by the IR boundary condition. In case we have a well-defined IR boundary condition at the wallz=zm, we could calculate the value of chiral condensate by solving the equation of motion for the bulk scalar X. In the case of the AdS black hole we could have a well defined IR boundary condition at the black hole horizon, which allows us to calculate the chiral condensate. For instance, in [148], it is shown that with the AdS black hole background the chiral condensate together with the current quark mass is zero in  both the hard wall and soft wall models. This is easy to see from the solution of X0 in the AdS black hole background [148, 149]: (6.10)X0(z)=z(mqF21(14,14,12,z4zh4)+σqz2F21(34,34,32,z4zh4)).
At z=zh, both terms in X0(z) diverge logarithmically, which requires to set both of them zero: mq=0, σ=0. This is different from real QCD, where current quark mass can be nonzero in the regime T>Tc.The finite temperature phase structure of the Sakai-Sugimoto model was analyzed in [150] to explore deconfinement and chiral symmetry restoration. Depending on a value of the model parameter, it is predicted that deconfinement and chiral symmetry restoration happens at the same temperature or the presence of a deconfined phase with broken chiral symmetry [150]. Phase structure of a stringy D3/D7 model has extensively studied in [151–153].


### 6.3. Equation of State and Susceptibility

Apart from the chiral condensate, various thermodynamic quantities could serve as an indicator for a transition from hadron to quark-gluon phase. Energy density, entropy, pressure, and susceptibilities are such examples. We first consider energy density and pressure. Schematically, based on the ideal gas picture we discuss how the energy density and pressure tell hadronic matter to quark-gluon plasma. At low temperature thermodynamics of hadron gas will be dominated by pions which are almost massless, while in QGP quarks and gluons are the relevant degrees of freedom. Energy density and pressure of massless pions are(6.11)ϵ=γπ230T4,p=γπ290T4,
where the number of degrees of freedom γ is three. In the QGP, they are given by (6.12)ϵ=γπ230T4+B,p=γπ290T4-B,
where γ=37, and B is the bag constant. Apart from the bag constant, the degeneracy factor γ changes from 3 to 37, and therefore we can expect that the energy density and pressure will increase rapidly at the transition point. Since the dual of the boundary energy-momentum tensor Tμν is the metric, we can obtain the energy density and pressure of a boundary gauge theory from the near-boundary behavior of the gravity solution. To demonstrate how-to, we follow [154, 155]. We first rewrite the gravity solution in the Fefferman-Graham coordinate [156]: (6.13)ds2=1z2(gμνdxμdxν-dz2).
Next, we expand the metric gμν at the boundary z→0: (6.14)gμν=gμν(0)+z2gμν(2)+gμν(4)+⋯.
Now we consider flat 4D metric such that gμν(0)=ημν. Then gμν(2)=0 and the vacuum expectation value of the energy momentum tensor is given by (6.15)〈Tμν〉=const⋅gμν(4).
For example, we consider an AdS black hole in the Fefferman-Graham coordinate: (6.16)ds2=1z2((1-z4/zh4)(1+z4/zh4)dt2-(1+z4zh4)dx⃗2-dz2).
Here the temperature is defined by T=2/(πzh). Then, we read off (6.17)〈Tμν〉≈diag⁡(3zh4,1zh4,1zh4,1zh4),
which satisfies ϵ=3p. There have been many works on the equations of state for a holographic matter at finite temperature [142, 157–160]. In [161], the energy density, pressure, and entropy of a deconfined pure Yang-Mills matter were evaluated in the improved holographic QCD model [162, 163]. The energy density and pressure vanish at low temperature, and at the critical temperature, Tc~235 MeV, they jump up to a finite value, showing the first-order phase transition. It is interesting to note that in [164] some high-precision lattice QCD simulations were performed with increasing Nc at finite temperature, and the results were compared with those from holographic QCD studies.Various susceptibilities are also useful quantities to characterize phases of QCD. For instance, the quark number susceptibility has been calculated in holographic QCD in a series of works [148, 165]. The quark number susceptibility was originally proposed as a probe of the QCD chiral phase transition at zero chemical potential [127, 131]:(6.18)χq=∂nq∂μq.
In terms of the retarded Green function GμνR(ω,k), the quark number susceptibility can be written as [166] (6.19)χq(T,μ)=lim⁡k→0⁡Re(GttR(ω=0,k)).
In [165], it is claimed that quark number susceptibility will show a sudden jump at Tc in high-density regime, and so QCD phase transition in low-temperature and high-density regime will be always first order. Thermodynamics of a charged dilatonic black hole, which is asymptotically RN-AdS black hole in the UV and AdS2×R3 in the IR, including the quark number susceptibility were extensively studied in [167]. The critical end point of the QCD phase diagram was studied in [168] by considering the critical exponents of the specific heat, number density, quark number susceptibility, and the relation between the number density and chemical potential at finite chemical potential and temperature. It is shown that the critical end point is located at T=143 MeV and μ=783 MeV in the QCD phase diagram [168].


### 6.4. Dense Baryonic Matter

Understanding the properties of dense QCD is of key importance for laboratory physics such as heavy ion collision and for our understanding of the physics of stable/unstable nuclei and of various astrophysical objects such as neutron stars.To expose an essential physics of dense nuclear matter, we take the Walecka model [169, 170], which describes nuclear matter properties rather well, as an example. The simplest version of the model contains the nucleon ψ, omega meson ω, and an isospin singlet, Lorentz scalar meson σ whose minimal Lagrangian is (6.20)L=ψ¯(i∂+gσσ-gωω)ψ+12(∂μσ∂μσ-mσ2σ2)-14FμνFμν+12mω2ωμωμ.
Within the mean field approximation, the properties of nuclear matter are mostly determined by the scalar mean field σ¯=(gσ/mσ2)ns and the mean field of the time component of the ω field ω¯0=(gω/mω2)n, where n is the baryon number density and ns is the scalar density. For instance, the pressure of the nuclear matter described by the Walecka is(6.21)P=14π2[23EF*pF3-mN*2EF*pF+mN*4ln⁡(EF*+pFmN*)]+12gω2mω2n2-12gσ2mσ2ns2,
where (6.22)EF*=pF2+mN*2,mN*=mN-gσ2mσ2ns.
Further, many successful predictions based on the Walecka model and its generalized versions, Quantum Hadrodynamics, require large scalar and vector fields in nuclei. This implies that to gain a successful description of nuclear matter or nuclei, having both scalar and vector mean fields in the model seems crucial. The importance of the interplay between the scalar and vector fields can be also seen in the static nonrelativistic potential between two nucleons. The nucleon-nucleon potential from single σ-exchange and single ω-exchange is given by (6.23)V(r)=gω24π1re-mωr-gσ24π1re-mσr.
Note that single σ-exchange can be replaced by two pion exchange. If gω>gσ and mω>mσ, then the potential in (6.23) captures some essential features of the two nucleon potential to form stable nuclear matter: repulsive at short distance and attraction at intermediate and long distance. We remark here that the scalar field in the Walecka model may not be the scalar associated with a linear realization of usual chiral symmetry breaking in QCD; see, for instance, [171].The hard wall model or soft wall model in its original form does not do much in dense matter. This is primarily due to its simple structure and chiral symmetry. Suppose that we turn on the time component of aU(1) bulk vector field dual to a boundary number operator, Vt(z)=μ+ρz2. To incorporate this U(1) bulk field into the hard wall model, we consider U(2) chiral symmetry. The covariant derivative with U(1) vector and axial-vector is given by DμX=∂μX-iALμX+iXARμ and it becomes DμX=∂μX-iX(ALμ-ARμ). Therefore, the U(1) bulk field Vμ=ALμ+ARμ does not couple to the scalar X, meaning that the physical properties of X are not affected by the chemical potential or number density. Note, however, that the vacuum energy of the hard wall or soft wall model should depend on the chemical potential and number density by the AdS/CFT. One simple way to study the physics of dense matter in the hard or soft wall model is to work with higher-dimensional terms in the action. For instance, the role of dimension six terms in the hard wall model was studied in free space [172, 173]. If we turn on the number density through the U(1) bulk field, we have a term like X02FV2, where FV is the field strength of the bulk U(1) gauge field [174]. Then we may see interplay between number density and chiral condensate encoded in X0. In [175], based on the hard wall model with the Chern-Simons term it is shown that there exists a Chern-Simons coupling between vector and axial-vector mesons at finite baryon density. This mixes transverse ρ and a1 mesons and leads to the condensation of the vector and axial-vector mesons. The role of the scalar density or the scalar field in the hard wall model was explored in [176]. In [139], a back-reaction due to the density is studied in the hard wall model.Physics of dense matter in Sakai-Sugimoto model has been developed with/without the source term for baryon charge [177–181]. For instance, in [180] localized and smeared source terms are introduced and a Fermi sea has been observed, though there are no explicit fermionic modes in the model. A deficit with the Sakai-Sugimoto model for nuclear matter might be the absence of the scalar field which is quite important together with U(1) vector field. The phase structure of the D3/D7 model at finite density is studied in [182, 183]. The nucleon-nucleon potential is playing very important role in understanding the properties of nuclear matter. For example, one of the conventional methods to study nuclear matter is to work with the independent-pair approximation, Brueckner’s theory, where two-nucleon potentials are essential inputs. Holographic nuclear forces were studied in [184–187].


## 6.1. Confinement/Deconfinement Transition

We first discuss the deconfinement transition. In holographic QCD, the confinement to deconfinement phase transition is described by the Hawking-Page transition [132], a phase transition between the Schwarzschild-AdS black hole and thermal AdS backgrounds. This identification was made in [12]. One simple reasoning for this identification is from the observation that the Polyakov expectation value is zero on the thermal AdS geometry, while it is finite on the AdS black hole. See Appendix F for some more description of the Hawking-Page transition and the Polyakov expectation in thermal AdS and AdS black hole. In low-temperature confined phase, thermal AdS, which is nothing but the AdS metric in Euclidean space, dominates the partition function, while at high temperature, AdS-black hole geometry does. This was first discovered in the finite volume boundary case in [12]. In the bottom-up model, it is shown that the same phenomena happen also for infinite boundary volume if there is a finite scale associated with the fifth direction [59].Here we briefly summarize the Hawking-Page analysis of [59] done in the hard wall model. In the Euclidean gravitational action given by(6.2)Sgrav=-12κ2∫d5xg(R+12L2),
where κ2=8πG5 and L is the length scale of the AdS5, there are two solutions for the equations of motion derived from the gravitational action. The one is the sliced thermal AdS (tAdS):
(6.3)ds2=L2z2(dτ2+dz2+dx⃗32),
where the radial coordinate runs from the boundary of tAdS space z=0 to the cut-off zm. Here τ is for the compactified Euclidean time-direction with periodicity β′. The other solution is the AdS black hole (AdSBH) with the horizon zh:(6.4)ds2=L2z2(f(z)dτ2+dz2f(z)+dx⃗32),
where f(z)=1-(z/zh)4. The Hawking temperature of the black hole solution is T=1/(πzh), which is given by regularizing the metric near the horizon. At the boundary z=ϵ the periodicity of the time-direction in both backgrounds is the same and so the time periodicity of the tAdS is given by (6.5)β=πzhf(ϵ).
Now we calculate the action density V, which is defined by the action divided by the common volume factor of R3. The regularized action density of the tAdS is given by(6.6)V1(ϵ)=4L3κ2∫0β′dτ∫ϵzIRdzz5,
and that of the AdSBH is given by(6.7)V2(ϵ)=4L3κ2∫0πzhdτ∫ϵz¯dzz5,
where z¯=min(zm,zh). Then, the difference of the regularized actions is given by(6.8)ΔVg=lim⁡ϵ→0[V2(ϵ)-V1(ϵ)]={L3πzhκ212zh4,zm<zh,L3πzhκ2(1zm4-12zh4),zm>zh.
When ΔVg is positive (negative), tAdS (AdSBH) is stable. Thus, at ΔVg=0 there exists a Hawking-Page transition. In the first case zm<zh, there is no Hawking-Page transition and the tAdS is always stable. In the second case zm>zh, the Hawking-Page transition occurs at(6.9)Tc=21/4(πzm),
and at low temperature T<Tc (at high temperature T>Tc) the thermal AdS (the AdS black hole) geometry becomes a dominant background. When we fix the IR cutoff by the ρ meson mass, we obtain 1/zm=323 MeV and Tc=122 MeV. In the soft wall model, Tc=191MeV [59].This work has been extended in various directions. The authors of [133] revisited the thermodynamics of the hard wall and soft wall model. They used holographic renormalization to compute the finite actions of the relevant supergravity backgrounds and verify the presence of a Hawking-page type phase transition. They also showed that the entropy, in the gauge theory side, jumps from N0 to N2 at the transition point [133]. In [134], the extension was done by studying the thermodynamics of AdS black holes with spherical or negative constant curvature horizon, dual to a non-supersymmetric Yang-Mills theory on a sphere or hyperboloid respectively. They also studied charged AdS black holes [135] in the grand canonical ensemble, corresponding to a Yang-Mills theory at finite chemical potential, and found that there is always a gap for the infrared cutoff due to the existence of a minimal horizon for the charged AdS black holes with any horizon topology [134]. With an assumption that the gluon condensate melts out at finite temperature, a Hawking-Page type transition between the dilaton AdS geometry in (4.4) and the usual AdS black hole has studied in [136].The effect of the number of quark flavorsNf and baryon number density on the critical temperature was investigated by considering a bulk meson action together with the gravity action in [137]. It is shown that the critical temperature decreases with increasing Nf. As the number density was raised, the critical temperature begins to drop, but it saturates to a constant value even at very large density. This is mostly due to the absence of the back-reaction from number density [137]. The back-reaction due to the number density has included in [138–140]. In [141], deconfinement transition of AdS/QCD with 𝒪(α′3) corrections was investigated. In [142], thermodynamics of the asymptotically-logarithmically-AdS black-hole solutions of 5D dilaton gravity with a monotonic dilaton potential are analyzed in great detail, where it is shown that in a special case, where the asymptotic geometry in the string frame reduces to flat space with a linear dilaton, the phase transition could be second order. The renormalized Polyakov loop in the deconfined phase of a pure SU(3) gauge theory was computed in [143] based on a soft wall metric model. The result obtained in this work is in good agreement with the one from lattice QCD simulations.Due to this Hawking-Page transition, we are not to use the black hole in the confined phase, and so we are not to obtain the temperature dependence of any hadronic observables. This is consistent with largeNc QCD at leading order. For instance, it was shown in [144, 145] that the Wilson loops, both time-like and space-like, and the chiral condensate are independent of the temperature in confining phase to leading order in 1/Nc. This means that the chiral and deconfinement transitions are first order. The deconfinement and chiral phase transitions of an SU(N) gauge theory at large Nc were also discussed in [146]. However, in reality we observe temperature dependence of hadronic quantities, and therefore we have to include large Nc corrections in holographic QCD in a consistent way. A quick fix-up for this might be to use the temperature dependent chiral condensate as an input in a holographic QCD model and study how this temperature dependence conveys into other hadronic quantities [147].


## 6.2. Chiral Transition

Now we turn to the chiral transition of QCD based on the chiral condensate. In the hard wall model, the chiral symmetry is broken, in a sense, by the IR boundary condition. In case we have a well-defined IR boundary condition at the wallz=zm, we could calculate the value of chiral condensate by solving the equation of motion for the bulk scalar X. In the case of the AdS black hole we could have a well defined IR boundary condition at the black hole horizon, which allows us to calculate the chiral condensate. For instance, in [148], it is shown that with the AdS black hole background the chiral condensate together with the current quark mass is zero in  both the hard wall and soft wall models. This is easy to see from the solution of X0 in the AdS black hole background [148, 149]: (6.10)X0(z)=z(mqF21(14,14,12,z4zh4)+σqz2F21(34,34,32,z4zh4)).
At z=zh, both terms in X0(z) diverge logarithmically, which requires to set both of them zero: mq=0, σ=0. This is different from real QCD, where current quark mass can be nonzero in the regime T>Tc.The finite temperature phase structure of the Sakai-Sugimoto model was analyzed in [150] to explore deconfinement and chiral symmetry restoration. Depending on a value of the model parameter, it is predicted that deconfinement and chiral symmetry restoration happens at the same temperature or the presence of a deconfined phase with broken chiral symmetry [150]. Phase structure of a stringy D3/D7 model has extensively studied in [151–153].


## 6.3. Equation of State and Susceptibility

Apart from the chiral condensate, various thermodynamic quantities could serve as an indicator for a transition from hadron to quark-gluon phase. Energy density, entropy, pressure, and susceptibilities are such examples. We first consider energy density and pressure. Schematically, based on the ideal gas picture we discuss how the energy density and pressure tell hadronic matter to quark-gluon plasma. At low temperature thermodynamics of hadron gas will be dominated by pions which are almost massless, while in QGP quarks and gluons are the relevant degrees of freedom. Energy density and pressure of massless pions are(6.11)ϵ=γπ230T4,p=γπ290T4,
where the number of degrees of freedom γ is three. In the QGP, they are given by (6.12)ϵ=γπ230T4+B,p=γπ290T4-B,
where γ=37, and B is the bag constant. Apart from the bag constant, the degeneracy factor γ changes from 3 to 37, and therefore we can expect that the energy density and pressure will increase rapidly at the transition point. Since the dual of the boundary energy-momentum tensor Tμν is the metric, we can obtain the energy density and pressure of a boundary gauge theory from the near-boundary behavior of the gravity solution. To demonstrate how-to, we follow [154, 155]. We first rewrite the gravity solution in the Fefferman-Graham coordinate [156]: (6.13)ds2=1z2(gμνdxμdxν-dz2).
Next, we expand the metric gμν at the boundary z→0: (6.14)gμν=gμν(0)+z2gμν(2)+gμν(4)+⋯.
Now we consider flat 4D metric such that gμν(0)=ημν. Then gμν(2)=0 and the vacuum expectation value of the energy momentum tensor is given by (6.15)〈Tμν〉=const⋅gμν(4).
For example, we consider an AdS black hole in the Fefferman-Graham coordinate: (6.16)ds2=1z2((1-z4/zh4)(1+z4/zh4)dt2-(1+z4zh4)dx⃗2-dz2).
Here the temperature is defined by T=2/(πzh). Then, we read off (6.17)〈Tμν〉≈diag⁡(3zh4,1zh4,1zh4,1zh4),
which satisfies ϵ=3p. There have been many works on the equations of state for a holographic matter at finite temperature [142, 157–160]. In [161], the energy density, pressure, and entropy of a deconfined pure Yang-Mills matter were evaluated in the improved holographic QCD model [162, 163]. The energy density and pressure vanish at low temperature, and at the critical temperature, Tc~235 MeV, they jump up to a finite value, showing the first-order phase transition. It is interesting to note that in [164] some high-precision lattice QCD simulations were performed with increasing Nc at finite temperature, and the results were compared with those from holographic QCD studies.Various susceptibilities are also useful quantities to characterize phases of QCD. For instance, the quark number susceptibility has been calculated in holographic QCD in a series of works [148, 165]. The quark number susceptibility was originally proposed as a probe of the QCD chiral phase transition at zero chemical potential [127, 131]:(6.18)χq=∂nq∂μq.
In terms of the retarded Green function GμνR(ω,k), the quark number susceptibility can be written as [166] (6.19)χq(T,μ)=lim⁡k→0⁡Re(GttR(ω=0,k)).
In [165], it is claimed that quark number susceptibility will show a sudden jump at Tc in high-density regime, and so QCD phase transition in low-temperature and high-density regime will be always first order. Thermodynamics of a charged dilatonic black hole, which is asymptotically RN-AdS black hole in the UV and AdS2×R3 in the IR, including the quark number susceptibility were extensively studied in [167]. The critical end point of the QCD phase diagram was studied in [168] by considering the critical exponents of the specific heat, number density, quark number susceptibility, and the relation between the number density and chemical potential at finite chemical potential and temperature. It is shown that the critical end point is located at T=143 MeV and μ=783 MeV in the QCD phase diagram [168].


## 6.4. Dense Baryonic Matter

Understanding the properties of dense QCD is of key importance for laboratory physics such as heavy ion collision and for our understanding of the physics of stable/unstable nuclei and of various astrophysical objects such as neutron stars.To expose an essential physics of dense nuclear matter, we take the Walecka model [169, 170], which describes nuclear matter properties rather well, as an example. The simplest version of the model contains the nucleon ψ, omega meson ω, and an isospin singlet, Lorentz scalar meson σ whose minimal Lagrangian is (6.20)L=ψ¯(i∂+gσσ-gωω)ψ+12(∂μσ∂μσ-mσ2σ2)-14FμνFμν+12mω2ωμωμ.
Within the mean field approximation, the properties of nuclear matter are mostly determined by the scalar mean field σ¯=(gσ/mσ2)ns and the mean field of the time component of the ω field ω¯0=(gω/mω2)n, where n is the baryon number density and ns is the scalar density. For instance, the pressure of the nuclear matter described by the Walecka is(6.21)P=14π2[23EF*pF3-mN*2EF*pF+mN*4ln⁡(EF*+pFmN*)]+12gω2mω2n2-12gσ2mσ2ns2,
where (6.22)EF*=pF2+mN*2,mN*=mN-gσ2mσ2ns.
Further, many successful predictions based on the Walecka model and its generalized versions, Quantum Hadrodynamics, require large scalar and vector fields in nuclei. This implies that to gain a successful description of nuclear matter or nuclei, having both scalar and vector mean fields in the model seems crucial. The importance of the interplay between the scalar and vector fields can be also seen in the static nonrelativistic potential between two nucleons. The nucleon-nucleon potential from single σ-exchange and single ω-exchange is given by (6.23)V(r)=gω24π1re-mωr-gσ24π1re-mσr.
Note that single σ-exchange can be replaced by two pion exchange. If gω>gσ and mω>mσ, then the potential in (6.23) captures some essential features of the two nucleon potential to form stable nuclear matter: repulsive at short distance and attraction at intermediate and long distance. We remark here that the scalar field in the Walecka model may not be the scalar associated with a linear realization of usual chiral symmetry breaking in QCD; see, for instance, [171].The hard wall model or soft wall model in its original form does not do much in dense matter. This is primarily due to its simple structure and chiral symmetry. Suppose that we turn on the time component of aU(1) bulk vector field dual to a boundary number operator, Vt(z)=μ+ρz2. To incorporate this U(1) bulk field into the hard wall model, we consider U(2) chiral symmetry. The covariant derivative with U(1) vector and axial-vector is given by DμX=∂μX-iALμX+iXARμ and it becomes DμX=∂μX-iX(ALμ-ARμ). Therefore, the U(1) bulk field Vμ=ALμ+ARμ does not couple to the scalar X, meaning that the physical properties of X are not affected by the chemical potential or number density. Note, however, that the vacuum energy of the hard wall or soft wall model should depend on the chemical potential and number density by the AdS/CFT. One simple way to study the physics of dense matter in the hard or soft wall model is to work with higher-dimensional terms in the action. For instance, the role of dimension six terms in the hard wall model was studied in free space [172, 173]. If we turn on the number density through the U(1) bulk field, we have a term like X02FV2, where FV is the field strength of the bulk U(1) gauge field [174]. Then we may see interplay between number density and chiral condensate encoded in X0. In [175], based on the hard wall model with the Chern-Simons term it is shown that there exists a Chern-Simons coupling between vector and axial-vector mesons at finite baryon density. This mixes transverse ρ and a1 mesons and leads to the condensation of the vector and axial-vector mesons. The role of the scalar density or the scalar field in the hard wall model was explored in [176]. In [139], a back-reaction due to the density is studied in the hard wall model.Physics of dense matter in Sakai-Sugimoto model has been developed with/without the source term for baryon charge [177–181]. For instance, in [180] localized and smeared source terms are introduced and a Fermi sea has been observed, though there are no explicit fermionic modes in the model. A deficit with the Sakai-Sugimoto model for nuclear matter might be the absence of the scalar field which is quite important together with U(1) vector field. The phase structure of the D3/D7 model at finite density is studied in [182, 183]. The nucleon-nucleon potential is playing very important role in understanding the properties of nuclear matter. For example, one of the conventional methods to study nuclear matter is to work with the independent-pair approximation, Brueckner’s theory, where two-nucleon potentials are essential inputs. Holographic nuclear forces were studied in [184–187].


## 7. Closing Remarks

The holographic QCD model has proven to be a successful and promising analytic tool to study nonperturbative nature of low energy QCD. However, its success should always come with “qualitative” since it is capturing only largeNc leading physics. To have any transitions from “qualitative” to “quantitative”, we have to invent a way to calculate subleading corrections in a consistent manner. A bit biased, but the most serious defect of the approach based on the gauge/gravity duality might be that it offers inherently macroscopic descriptions of a physical system. For instance, we may understand the QCD confinemnt/deconfinement transition through the Hawking-Page transition, qualitatively. Even though we accept generously the word “qualitatively”, we are not to be satisfied completely since we do not know how gluons and quarks bound together to form a color singlet hadron or how hadrons dissolve themselves into quark and gluon degrees of freedom. In this sense, the holographic QCD cannot be stand-alone. Therefore, the holographic QCD should go together with conventional QCD-based models or theories to guide them qualitatively and to gain microscopic pictures revealed by the conventional approaches.Finally, we collect some interesting works done in bottom-up models that are not yet properly discussed in this review. Due to our limited knowledge, we could not list all of the interesting works and most results from top-down models will not be quoted. To excuse this defect we refer to recent review articles on holographic QCD [28, 188–197].Deep inelastic scattering has been studied in gauge/gravity duality [198–209]. Light and heavy mesons were studied in the soft-wall holographic approach [210].Unusual bound states of quarks are also interesting subjects to work in holographic QCD. In [211], the multiquark potential was calculated and tetra-quarks were discussed in AdS/QCD. Based on holographic quark-antiquark potential in the static limit, the masses of the states X(3872) or Y(3940) were predicted and also tetra-quark masses with open charm and strangeness were computed in [212]. A hybrid exotic meson, π1(1400), was discussed in [213]. The spectrum of baryons with two heavy quarks was predicted in [214].Low-energy theorems of QCD and spectral density of the Dirac operator were studied in the soft wall model [215]. A holographic model of hadronization was suggested in [216].The equation of state for a cold quark matter was calculated in the soft wall metric model with aU(1) gauge field. The result is in agreement with phenomenology [217].



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