arXiv:0704.1657v3  [hep-th]  28 May 2008
arXiv:0704.1657
Bubbling Surface Operators
And S-Duality
Jaume Gomis1 and Shunji Matsuura2
Perimeter Institute for Theoretical Physics
Waterloo, Ontario N2L 2Y5, Canada1,2
Department of Physics
University of Tokyo, 7-3-1 Hongo, Tokyo2
Abstract
We construct smooth asymptotically AdS5×S5 solutions of Type IIB supergravity
corresponding to all the half-BPS surface operators in N = 4 SYM. All the parameters
labeling a half-BPS surface operator are identified in the corresponding bubbling geome-
try. We use the supergravity description of surface operators to study the action of the
SL(2, Z) duality group of N = 4 SYM on the parameters of the surface operator, and find
that it coincides with the recent proposal by Gukov and Witten in the framework of the
gauge theory approach to the geometrical Langlands with ramification. We also show that
whenever a bubbling geometry becomes singular that the path integral description of the
corresponding surface operator also becomes singular.
04/2007
jgomis@perimeterinstitute.ca
smatsuura@perimeterinstitute.ca
http://arxiv.org/abs/0704.1657v3
1. Introduction and Summary
Gauge invariant operators play a central role in the gauge theory holographically
describing quantum gravity with AdS boundary conditions [1][2][3], as correlation functions
of gauge invariant operators are the only observables in the boundary gauge theory. Finding
the bulk description of all gauge invariant operators is necessary in order to be able to
formulate an arbitrary bulk experiment in terms of gauge theory variables.
In this paper we provide the bulk description of a novel class of half-BPS operators
in N = 4 SYM which are supported on a surface Σ [4]. These nonlocal surface operators
OΣ are defined by quantizing N = 4 SYM in the presence of a certain codimension two
singularity for the classical fields of N = 4 SYM. The singularity characterizing such a
surface operator OΣ depends on 4M real parameters, whereM is the number of U(1)’s left
unbroken byOΣ. Surface operators are a higher dimensional generalization of Wilson and ’t
Hooft operators, which are supported on curves and induce a codimension three singularity
for the classical fields appearing in the Lagrangian. In this paper we extend the bulk
description of all half-BPS Wilson loop operators found in [5] (see also1 [8][9][10][11][12])
to all half-BPS surface operators.
We find the asymptotically AdS5×S5 solutions of Type IIB supergravity corresponding
to all half-BPS surface operators OΣ in N = 4 U(N) SYM. The topology and geometry
of the “bubbling” solution is completely determined in terms of some data, very much like
in the case studied by Lin, Lunin and Maldacena (LLM) in the context of half-BPS local
operators [13]2. In fact, we identify the system of equations determining the supergravity
solution corresponding to the half-BPS surface operators inN = 4 SYM with that obtained
by “analytic” continuation of the LLM equations [13][16].
The data determining the topology and geometry of a supergravity solution is charac-
terized by the position of a collection of M point particles in a three dimensional space X ,
whereX is a submanifold of the ten dimensional geometry. Different particle configurations
give rise to different asymptotically AdS5×S5 geometries.
1 The description of Wilson loops in the fundamental representation goes back to [6][7].
2 The bubbling geometry description of half-BPS Wilson loops was found in [10][11] while that
of half-BPS domain wall operators was found in [11][14]. For the bubbling Calabi-Yau geometries
for Wilson loops in Chern-Simons, see [15].
Fig. 1: a) The metric and five-form flux is determined once the position of the
particles in X – labeled by coordinates (~xl, yl) where y ≥ 0 – is given. The l-th
particle is associated with a point Pl ∈ X. b) The configuration corresponding to
the AdS5×S
5 vacuum.
Even though the choice of a particle distribution in X completely determines the
geometry and topology of the metric and the corresponding RR five-form field strength,
further choices have to be made to fully characterize a solution of Type IIB supergravity
on this geometry3.
Given a configuration ofM particles inX , the corresponding ten dimensional geometry
developsM non-trivial disks which end on the boundary4 of AdS5×S5 on a non-contractible
S1. Since Type IIB supergravity has two two-form gauge fields, one from the NS-NS sector
and one from the RR sector, a solution of the Type IIB supergravity equations of motion
is fully determined only once the holonomy of the two-forms around the various disks is
specified:
l = 1, . . . ,M. (1.1)
Therefore, an asymptotically AdS5×S5 solution depends on the position of theM particles
in X – given by (~xl, yl) – and on the holonomies of the two-forms (1.1).
A precise dictionary is given between all the 4M parameters that label a half-BPS
surface operator OΣ and all the parameters describing the corresponding supergravity
solution. We show that the supergravity solution describing a half-BPS surface operator is
regular and that whenever the supergravity solution develops a singularity the N = 4 SYM
path integral description of the corresponding surface operator also develops a singularity.
We study the action of the SL(2, Z) symmetry of Type IIB string theory on the
supergravity solutions representing the half-BPS surface operators in N = 4 SYM. By
3 This is on top of the obvious choice of dilaton and axion, which gets identified with the
complexified coupling constant in N = 4 SYM.
4 The conformal boundary in this case is AdS3×S
1, where surface operators in N = 4 SYM
can be studied by specifying non-trivial boundary conditions.
using the proposed dictionary between the parameters of a supergravity solution and the
parameters of the corresponding surface operator, we can show that the action of S-duality
induced on the parameters of a surface operator coincides with the recent proposal by
Gukov and Witten [4] in the framework of the gauge theory approach to the geometrical
Langlands [17]5 with ramification.
Whether surface operators can serve as novel order parameters in gauge theory remains
an important open question. It is our hope that the viewpoint on these operators provided
by the supergravity solutions in this paper may help shed light on this crucial question.
The plan of the rest of the paper is as follows. In section 2 we study the gauge theory
singularities corresponding to surface operators in N = 4 SYM, study the symmetries
preserved by a half-BPS surface operator and review the proposal in [4] for the action of
S-duality on the parameters that a half-BPS surface operator depends on. We also compute
the scaling weight of these operators and show that it is invariant under Montonen-Olive
duality. In section 3 we construct the solutions of Type IIB supergravity describing the
half-BPS surface operators. We identify all the parameters that a surface operator depends
on in the supergravity solution and show that the action of S-duality on surface operators
proposed in [4] follows from the action of SL(2, Z) on the classical solutions of supergravity.
The Appendices contain some details omitted in the main text.
2. Surface Operators in Gauge Theories
A surface operator OΣ is labeled by a surface Σ in R1,3 and by a conjugacy class U of
the gauge group G. The data that characterizes a surface operator OΣ, the surface Σ and
the conjugacy class U , can be identified with that of an external string used to probe the
theory. The surface Σ corresponds to the worldsheet of a string while the conjugacy class
U is associated to the Aharonov-Bohm phase acquired by a charged particle encircling the
string.
The singularity6 in the gauge field produced by a surface operator is that of a non-
abelian vortex. This singularity in the gauge field can be characterized by the phase
5 See e.g. [18] for a review of the geometric Langlands program.
6 Previous work involving codimension two singularities in gauge theory include [19][20][21].
acquired by a charged particle circumnavigating around the string. This gives rise to a
group element7 U ⊂ U(N)
U ≡ P exp i
A ⊂ U(N), (2.1)
which corresponds to the Aharonov-Bohm phase picked up by the wavefunction of the
charged particle. Since gauge transformations act by conjugation U → gUg−1, a surface
operator is labeled by a conjugacy class of the gauge group.
By performing a gauge transformation, the matrix U can be diagonalized. If we
demand that the gauge field configuration is scale invariant – so that OΣ has a well defined
scaling weight – then the gauge field produced by a surface operator can then be written
α1 ⊗ 1N1 0 . . . 0
0 α2 ⊗ 1N2 . . . 0
. . .
0 0 . . . αM ⊗ 1NM
dθ, (2.2)
where θ is the polar angle in the R2 ⊂ R1,3 plane normal to Σ and 1n is the n-dimensional
unit matrix. We note that the matrix U takes values on the maximal torus TN = RN/ZN
of the U(N) gauge group. Therefore the parameters αi take values on a circle of unit
radius.
The surface operator corresponding to (2.2) spontaneously breaks the U(N) gauge
symmetry along Σ down to the so called Levi group L, where a group of Levi type is char-
acterized by the subgroup of U(N) that commutes with (2.2). Therefore, L =
l=1 U(Nl),
where N =
l=1Nl.
Since the gauge group is broken down to the Levi group L =
l=1 U(Nl) along Σ,
there is a further choice [4] in the definition of OΣ consistent with the symmetries and
equations of motion. This corresponds to turning on a two dimensional θ-angle for the
unbroken U(1)’s along the string worldsheet Σ. The associated operator insertion into the
N = 4 SYM path integral is given by:
Tr Fl
. (2.3)
7 We now focus on G = U(N) as it is the relevant gauge group for describing string theory
with asymptotically AdS5×S
5 boundary conditions.
The parameters ηi takes values in the maximal torus of the S-dual or Langlands dual
gauge group LG [4]. Therefore, since LG = U(N) for G = U(N), we have that the matrix
of θ-angles of a surface operator OΣ characterized by the Levi group L =
l=1 U(Nl) is
given by the L-invariant matrix:
η1 ⊗ 1N1 0 . . . 0
0 η2 ⊗ 1N2 . . . 0
. . .
0 0 . . . ηM ⊗ 1NM
. (2.4)
The parameters ηi, being two dimensional θ-angles, also take values on a circle of unit
radius.
Therefore, a surface operator OΣ in pure gauge theory with Levi group L =
l=1 U(Nl) is labeled by 2M L-invariant parameters (αl, ηl) up to the action of SM ,
which acts by permuting the different eigenvalues in (2.2) and (2.4). The operator is then
defined by expanding the path integral with the insertion of the operator (2.3) around
the singularity (2.2), and by integrating over connections that are smooth near Σ. In
performing the path integral, we must divide [4] by the gauge transformations that take
values in L =
l=1 U(Nl) when restricted to Σ. This means that the operator becomes
singular whenever the unbroken gauge symmetry near Σ gets enhanced, corresponding to
when eigenvalues in (2.2) and (2.4) coincide.
Surface Operators in N = 4 SYM
In a gauge theory with extra classical fields like N = 4 SYM, the surface operator OΣ
may produce a singularity for the extra fields near the location of the surface operator.
The only requirement is that the singular field configuration solves the equations of motion
of the theory away8 from the surface Σ. The global symmetries imposed on the operator
OΣ determine which classical fields in the Lagrangian develop a singularity near Σ together
with the type of singularity.
A complementary viewpoint on surface operators is to add new degrees of freedom on
the surface Σ. Such an approach to surface operators in N = 4 SYM has been considered
8 For pure gauge theory, the field configuration in (2.2) does satisfy the Yang-Mills equation
of motion DmF
mn = 0 away from Σ. Moreover, adding the two dimensional θ-angles (2.3) does
not change the equations of motion.
in [22][23] where the new degrees of freedom arise from localized open strings on a brane
intersection.
The basic effect of OΣ is to generate an Aharonov-Bohm phase corresponding to a
group element U (2.1). If we let z be the complex coordinate in the R2 ⊂ R1,3 plane
normal to Σ, the singularity in the gauge field configuration is then given by
, (2.5)
where AI are constant matrices. Scale invariance of the singularity – which we are going
to impose – restricts AI = 0 for I ≥ 2.
The operator OΣ can also excite a complex scalar field Φ of N = 4 SYM near Σ
while preserving half of the Poincare supersymmetries of N = 4 SYM. Imposing that the
singularity is scale invariant9 yields
, (2.6)
where Φ1 is a constant matrix.
A surface operator OΣ is characterized by the choice of an unbroken gauge group
L ⊂ G along Σ. Correspondingly, the singularity of all the fields excited by OΣ must be
invariant under the unbroken gauge group L. For L =
l=1 U(Nl) ⊂ U(N) the singularity
in the gauge field is the non-abelian vortex configuration in (2.2) and the two dimensional
θ-angles are given by (2.4). L-invariance together with scale invariance requires that Φ
develops an L-invariant pole near Σ:
β1 + iγ1 ⊗ 1N1 0 . . . 0
0 β2 + iγ2 ⊗ 1N2 . . . 0
. . .
0 0 . . . βM + iγM ⊗ 1NM
. (2.7)
Therefore, a half-BPS surface operator OΣ in N = 4 SYM with Levi group L =
l=1 U(Nl) is labeled by 4M L-invariant parameters (αl, βl, γl, ηl) up to the action of SM ,
which permutes the different eigenvalues in (2.2)(2.4)(2.7). The operator is defined by the
9 If we relax the restriction of scale invariance, one can then get other supersymmetric singu-
larities with higher order poles Φ =
and A (2.5). The surface operators associated with
these singularities may be relevant [4] for the gauge theory approach to the study of the geometric
Langlands program with wild ramification.
path integral of N = 4 SYM with the insertion of the operator (2.3) expanded around the
L-invariant singularities (2.2)(2.7) and by integrating over smooth fields near Σ. As in the
pure gauge theory case, we must mode out by gauge transformations that take values in
L ⊂ U(N) when restricted to Σ. The surface operator OΣ becomes singular whenever the
the parameters that label the surface operator (αl, βl, γl, ηl) for l = 1, . . . ,M are such that
they are invariant under a larger symmetry than L, the group of gauge transformations
we have to mode out when evaluating the path integral.
S-duality of Surface Operators
In N = 4 SYM the coupling constant combines with the four dimensional θ-angle into
a complex parameter taking values in the upper half-plane:
. (2.8)
The group of duality symmetries ofN = 4 SYM is an infinite discrete subgroup of SL(2, R),
which depends on the gauge group G. For N = 4 SYM with G = U(N) the relevant
symmetry group is SL(2, Z):
∈ SL(2, Z). (2.9)
Under S-duality τ → −1/τ and G gets mapped10 to the S-dual or Langlands dual gauge
group LG. For G = U(N) the S-dual group is LG = U(N), and SL(2, Z) is a symmetry of
the theory, which acts on the coupling of the theory by fractional linear transformations:
τ → aτ + b
cτ + d
. (2.10)
In [4], Gukov and Witten made a proposal of how S-duality acts on the parameters
(αl, βl, γl, ηl) labeling a half-BPS surface operator. The proposed action is given by [4]:
(βl, γl) → |cτ + d| (βl, γl)
(αl, ηl) → (αl, ηl)M−1.
(2.11)
10 For G not a simply-laced group, τ → −1/nτ , where n is the ratio of the length-squared of
the long and short roots of G.
With the aid of this proposal, it was shown in [4] that the gauge theory approach to
the geometric Langlands program pioneered in [17] naturally extends to the geometric
Langlands program with tame ramification.
Symmetries of half-BPS Surface Operators in N = 4 SYM
We now describe the unbroken symmetries of the half-BPS surface operators OΣ.
These symmetries play an important role in determining the gravitational dual description
of these operators, which we provide in the next section.
In the absence of any insertions, N = 4 SYM is invariant under the PSU(2, 2|4)
symmetry group. If we consider the surface Σ = R1,1 ⊂ R1,3, then Σ breaks the SO(2, 4)
conformal group to a subgroup. A surface operator OΣ supported on this surface inserts
into the gauge theory a static probe string. This surface is manifestly invariant under
rotations and translations in Σ and scale transformations. It is also invariant under the
action of inversion I : xµ → xµ/x2 and consequently11 invariant under special conformal
transformations in Σ. Therefore, the symmetries left unbroken by Σ = R1,1 generate
an SO(2, 2)× SO(2)23 subgroup of the SO(2, 4) conformal group, where SO(2)23 rotates
the plane transverse to Σ in R1,3. In Euclidean signature, the surface Σ =S2 preserves
an SO(1, 3) × SO(2)23 subgroup of the Euclidean conformal group. This surface can be
obtained from the surface Σ = R2 ∈ R4 by the action of a broken special conformal
generator and can also be used to construct a half-BPS surface operator OΣ in N = 4
Since the symmetry of a surface operator with Σ = R1,1 is SO(2, 2) × SO(2)23 one
can study such an operator either by considering the gauge theory in R1,3 or in AdS3×S1,
which can be obtained from R1,3 by a conformal transformation. Studying the gauge
theory in AdS3×S1 has the advantage of making the symmetries of the surface operator
manifest, as the conformal symmetries left unbroken by the surface act by isometries on
AdS3×S1. Surface operators in R1,3 are described by a codimension two singularity while
surface operators in AdS3×S1 are described by a boundary condition on the boundary of
AdS3. A surface operator with Σ = R
1,1 corresponds to a boundary condition on AdS3
in Poincare coordinates while a surface operator on Σ =S2 corresponds to a boundary
condition on global Euclidean AdS3.
11 We recall that a special conformal transformation Kµ is generated by IPµI, where Pµ is the
translation generator and I is an inversion.
The singularity in the classical fields produced by OΣ in (2.2)(2.7) is also invariant
under SO(2, 2). The N = 4 scalar field Φ carries charge under an SO(2)R subgroup of the
SO(6) R-symmetry and is therefore SO(4) invariant. The surface operator OΣ is therefore
invariant under SO(2, 2)×SO(2)a×SO(4), where SO(2)a is generated by the anti-diagonal
product12 of SO(2)23 × SO(2)R.
N = 4 SYM has sixteen Poincare supersymmetries and sixteen conformal super-
symmetries, generated by ten dimensional Majorana-Weyl spinors ǫ1 and ǫ2 of opposite
chirality. As shown in the Appendix A, the surface operator OΣ for Σ = R1,1 preserves
half of the Poincare and half of the conformal supesymmetries13 and is therefore half-BPS.
With the aid of these symmetries we study in the next section the gravitational de-
scription of half-BPS surface operators in N = 4 SYM.
Scaling Weight of half-BPS Surface Operators in N = 4 SYM
Conformal symmetry constraints the form of the OPE of the energy-energy tensor
Tmn with the operators in the theory. For a surface operator OΣ supported on Σ = R1,1,
SO(2, 2)× SO(2)23 invariance completely fixes the OPE of Tmn with OΣ:
< Tµν(x)OΣ >
< OΣ >
< Tij(x)OΣ >
< OΣ >
[4ninj − 3δij ] ; < Tµi(x)OΣ >= 0.
(2.12)
Here xm = (xµ, xi), where xµ are coordinates along Σ and ni = xi/r, and r is the radial
coordinate in the R2 transverse to R1,1. h is the scaling weight of OΣ, which generalizes
[24] the notion of conformal dimension of local conformal fields to surface operators.
In order to calculate the scaling dimension of a half-BPS surface operator OΣ in N = 4
SYM we evaluate the classical field configuration (2.2)(2.4)(2.7) characterizing a half-BPS
surface operator on the classical energy-momentum tensor of N = 4 SYM:
Tmn =
Tr[DmφDnφ−
gmn(Dφ)
2 − 1
(DmDn − gmnD2)φ2]
Tr[−FmlFnl +
gmnFlpFlp].
(2.13)
12 Since SO(2)a leaves Φ · z in (2.7) invariant.
13 For Σ =S2, the operator is also half-BPS, but it preserves a linear combination of Poincare
and special conformal supersymmetries.
A straightforward computation14 leads to:
h = − 2
l + γ
l ) = −
i + γ
i ). (2.14)
The action of an SL(2, Z) transformation (2.10) on the coupling constant of N = 4
SYM implies that:
Imτ → Imτ|cτ + d|2 . (2.15)
Combining this with the action (2.11) of Montonen-Olive duality on the parameters of the
surface operator, we find that the scaling weight (2.14) of a half-BPS surface operator OΣ
is invariant under S-duality:
h→ h. (2.16)
In this respect half-BPS surface operators behave like the half-BPS local operators of
N = 4 SYM, whose conformal dimension is invariant under SL(2, Z), and unlike the
half-BPS Wilson-’t Hooft operators whose scaling weight is not S-duality invariant [24].
3. Bubbling Surface Operators
In this section we find the dual gravitational description of the half-BPS surface op-
erators OΣ described in the previous section. The bulk description is given in terms of
asymptotically AdS5×S5 and singularity free solutions of the Type IIB supergravity equa-
tions of motion. The data from which the solution is uniquely determined encodes the
corresponding data about the surface operator OΣ.
The strategy to obtain these solutions is to make an ansatz for Type IIB supergravity
which is invariant under all the symmetries preserved by the half-BPS surface operators
OΣ. As discussed in the previous section, the bosonic symmetries preserved by a half-
BPS surface operator OΣ are SO(2, 2) × SO(4) × SO(2)a. Therefore the most general
ten dimensional metric invariant under these symmetries can be constructed by fibering
AdS3×S3×S1 over a three manifold X , where the symmetries act by isometries on the
fiber. The constraints imposed by unbroken supersymmetry on the ansatz are obtained by
demanding that the ansatz for the supergravity background possesses a sixteen component
14 Contact terms depending on α, β, γ and proportional to the derivative of the two-dimensional
δ-function appear when evaluating the on-shelll energy-momentum tensor. It would be interesting
to understand the physical content of these contact terms.
Killing spinor, which means that the background solves the Killing spinor equations of Type
IIB supergravity. A solution of the Killing spinor equations and the Bianchi identity for the
five-form field strength guarantee that the full set of equations of Type IIB supergravity
are satisfied and that a half-BPS solution has been obtained.
The problem of solving the Killing spinor equations of Type IIB supergravity with
an SO(2, 2)× SO(4)× SO(2)a symmetry can be obtained by analytic continuation of the
equations studied by LLM [13][16], which found the supergravity solutions describing the
half-BPS local operators of N = 4 SYM, which have an SO(4)×SO(4)×R symmetry. The
equations determining the metric and five-form flux can be read from [13][16], in which the
analytic continuation that we need to construct the gravitational description of half-BPS
surface operators OΣ was considered.
The ten dimensional metric and five-form flux is completely determined in terms of
data that needs to be specified on the three manifold X in the ten dimensional space. An
asymptotically AdS5×S5 metric is uniquely determined in terms of a function z(x1, x2, y),
where (x1, x2, y) ≡ (~x, y) are coordinates in X . The ten dimensional metric in the Einstein
frame is given by15
ds2 = y
2z + 1
2z − 1ds
2z − 1
2z + 1
dΩ3 +
4z2 − 1
(dχ+ V )2 +
4z2 − 1
(dy2 + dxidxi),
(3.1)
where ds2X = dy
2+dxidxi with y ≥ 0 and V is a one-form in X satisfying dV = 1/y ∗X dz.
AdS3 in Poincare coordinate corresponds to a surface operator on Σ = R
1,1 while AdS3
in global Euclidean coordinate corresponds to a surface operator on Σ =S2. The U(1)a
symmetry acts by shifts on χ while SO(2, 2) and SO(4) act by isometries on the coordinates
of AdS3 and S
3 respectively.
A non-trivial solution to the equations of motion is obtained by specifying a config-
uration of M point-like particles in X . The data from which the solution is determined
is the “charge” Ql of the particles together with their positions (~xl, yl) in X (see Figure
1). Given a “charge” distribution, the function z(x1, x2, y) solves the following differential
equation:
∂i∂iz(x1, x2, y) + y∂y
∂yz(x1, x2, y)
Qlδ(y − yl)δ(2)(~x− ~xl). (3.2)
15 The “analytic” continuation from the bubbling geometries dual to the half-BPS local opera-
tors is given by z → z, t → χ, y → −iy, ~x → i~x, dΩ3 → −ds
[13][16].
Introducing a “charge” at the point (~xl, yl) in X has the effect of shrinking
16 the
S1 with coordinate χ in (3.1) to zero size at that point. In order for this to occur in a
smooth fashion the magnitude of the “charge” has to be fixed [13][16] so that Ql = 2πyl.
Therefore, the independent data characterizing the metric and five-form of the solution is
the position of the M “charges”, given by (~xl, yl).
In summary, a smooth half-BPS SO(2, 2)× SO(4)× SO(2)a invariant asymptotically
AdS5×S5 metric (3.1) solving the Type IIB supergravity equations of motion is found by
solving (3.2) subject to the boundary condition z(x1, x2, 0) = 1/2 [13][16], so that the S
in (3.1) shrinks in a smooth way at y = 0. The function z(x1, x2, y) is given by
z(x1, x2, y) =
zl(x1, x2, y), (3.3)
where
zl(x1, x2, y) =
(~x− ~xl)2 + y2 + y2l
((~x− ~xl)2 + y2 + y2l )2 − 4y2l y2
, (3.4)
and V can be computed from z(x1, x2, y) from dV = 1/y ∗X dz. Both the metric and
five-form field strength are determined by an integer M and by (~xl, yl) for l = 1, . . . ,M .
Topology of Bubbling Solutions And Two-form Holonomies
The asymptotically AdS5×S5 solutions constructed from (3.3)(3.4) are topologically
quite rich. In particular, a solution withM point “charges” hasM topologically non-trivial
S5’s. We can associate to each point Pl ∈ X a corresponding five-sphere S5l . S5l can be
constructed by fibering the S1×S3 in the geometry (3.1) over a straight line between the
point (~xl, 0) and the point (~xl, yl) in X . The topology of this manifold is indeed an S
as an S5 can be represented17 by an S1×S3 fibration over an interval where the S1 and S3
shrink to zero size at opposite ends of the interval, which is what happens in our geometry
where the S3 shrinks at (~xl, 0) while the S
1 shrinks at the other endpoint (~xl, yl).
16 Near y = yl the form of the relevant part of the metric is that of the Taub-NUT space. Fixing
the value of the “charge” at y = yl by imposing regularity of the metric coincides with the usual
regularity constraint on the periodicity of the circle in Taub-NUT space.
17 This can be seen explicitly by writing dΩ5 = cos
2 θdΩ3 + dθ
2 + sin2 θdφ2.
Fig. 2: A topologically non-trivial S5 can be constructed by fibering S1×S3 over
an interval connecting the y = 0 plane and the location of the “charge” at the
point Pl ∈ X with (~xl, yl) coordinates.
Following [13][16] we can now integrate the five-form flux over the topologically non-
trivial S5’s (see Appendix B):
F5 = y
l . (3.5)
Since flux has to be quantized, the position in the y-axis of the l-th particle in X is also
quantized
y2l = 4πNll
p Nl ∈ Z, (3.6)
where lp is the ten dimensional Planck length. For an asymptotically AdS5×S5 geometry
with radius of curvature R4 = 4πNl4p, which is dual to N = 4 U(N) SYM, we have that
the total amount of five-form flux must be N :
Nl. (3.7)
The asymptotically AdS5×S5 solutions constructed from (3.3)(3.4) also contain non-
trivial surfaces. In particular, a solution with M point “charges” has M non-trivial disks
Dl. Just as in the case of the S
5’s, we can associate to each point Pl ∈ X a disk Dl.
Inspection of the asymptotic form of the metric (3.1) given in (3.3)(3.4) reveals that
the metric is conformal to AdS3×S1. This geometry on the boundary of AdS5×S5, which
is where the dual N = 4 U(N) SYM lives, is the natural background geometry on which to
study conformally invariant surface operators in N = 4 SYM. As explained in section 2, an
SO(2, 2)× SO(2)23 invariant surface operator can be defined by specifying a codimension
two singularity in R1,3 or by specifying appropriate boundary conditions for the classical
fields in the gauge theory at the boundary of AdS3×S1. In the latter formulation, the
worldsheet of the surface operator Σ is the boundary of AdS3.
Therefore, in the boundary of AdS5×S5 we have a non-contractible S1. If we fiber the
S1 parametrized by χ in (3.1) over a straight line connecting a point (~xl, yl) in X – where
the S1 shrinks to zero size – to a point in X corresponding to the boundary of AdS5×S5
– given by ~x, y → ∞ – we obtain a surface Dl. This surface is topologically a disk18 and
there are M of them for a “charge” distribution of M particles in X .
Fig. 3: A disk D can be constructed by fibering S1 over an interval connecting
the “charge” at the point Pl ∈ X with (~xl, yl) coordinates and the boundary of
AdS5×S
Due to the existence of the disks Di, the supergravity solution given by the metric
and five-form flux alone is not unique. Type IIB supergravity has a two-form gauge field
from the NS-NS sector and another one from the RR sector. In order to fully specify a
solution of Type IIB supergravity in the bubbling geometry (3.1) we must complement the
metric and the five-form with the integral of the two-forms around the disks19
αl = −
l = 1, . . . ,M, (3.8)
where we have used notation conducive to the later comparison with the parameters char-
acterizing a half-BPS surface operator OΣ. Since both BNS and BR are invariant under
large gauge transformations, the parameters (αl, ηl) take values on a circle of unit radius.
Apart from the M disks Dl, the bubbling geometry constructed from (3.3)(3.4) also
has topologically non-trivial S2’s. One can construct an S2 by fibering the S1 in (3.1) over
18 Such disks also appear in the study of the high temperature regime of N = 4 SYM, where
the bulk geometry [25] is the AdS Schwarzschild black hole, which also has a non-contractible S1
in the boundary which is contractible in the full geometry.
19 The overall signs in the identification are fixed by demanding consistent action of S-duality
of N = 4 SYM with that of Type IIB supergravity.
a straight line connecting the points Pl and Pm in X . Since the S
1 shrinks to zero size in
a smooth manner at the endpoints we obtain an S2. Therefore, to every pair of “charges”
in X , characterized by different points Pl and Pm in X , we can construct a corresponding
S2, which we label by S2l,m. The integral of BNS and BR over S
l,m do not give rise to new
parameters, since [S2l,m] = [Dl] − [Dm] in homology, and the periods can be determined
from (3.8).
Fig. 4: An S2 can be constructed by fibering S1 over an interval connecting the
“charge” at the point Pl ∈ X with a different “charge” at point Pm ∈ X.
Bubbling Geometries as Surface Operators
As we discussed in section 2, a surface operator OΣ is characterized by an unbroken
gauge group L ∈ U(N) along together with 4M L-invariant parameters (αl, βl, γl, ηl). On
the other hand, the Type IIB supergravity solutions we have described depend on the
positions (~xl, yl) of M “charged” particles in X and the two-form holonomies:
. (3.9)
We now establish an explicit dictionary between the parameters in gauge theory and the
parameters in supergravity.
For illustration purposes, it is convenient to start by considering the half-BPS surface
operator OΣ with the largest Levi group L, which is L = U(N) for G = U(N). U(N)
invariance requires that the singularity in the fields produced by OΣ take values in the
center of U(N). Therefore, the gauge field and scalar field produced by OΣ is given by
A = α01Ndθ
(β0 + iγ0)1N ,
(3.10)
where 1N is the identity matrix. We can also turn a two-dimensional θ-angle (2.3) for the
overall U(1), so that:
η = η01N . (3.11)
We now identify this operator with the supergravity solution obtained by having a
single point “charge” source in X (see Figure 1b). If we let the position of the “charge”
be (~x0, y0) then
z(x1, x2, y) =
(~x− ~x0)2 + y2 + y20
((~x− ~x0)2 + y2 + y20)2 − 4y20y2
VI = −ǫIJ
(xJ − xJ0 )((~x− ~x0)2 + y2 − y20)
2(~x− ~x0)2
((~x− ~x0)2 + y2 + y20)2 − 4y20y2
(3.12)
where V = VIdx
I . The metric (3.1) obtained using (3.12) is the metric of AdS5×S5. This
can be seen by the following change of variables [16]
x1 − x10 + i(x2 − x20) = rei(ψ+φ)
r = y0 sinhu sin θ
y = y0 coshu cos θ
(ψ − φ),
(3.13)
which yields the AdS5×S5 metric with AdS5 foliated by AdS3×S1 slices:
ds2 = y0
(cosh2 uds2AdS3 + du
2 + sinh2 udψ2) + (cos2 θdΩ3 + dθ
2 + sin2 θdφ2)
. (3.14)
We note that the U(1)a symmetry of the metric (3.1) – which acts by shifts on χ – identifies
via (3.13) an SO(2)R subgroup of the the SO(6) symmetry of the S
5, acting by shifts on
φ, with an SO(2)23 subgroup of the SO(2, 4) isometry group of AdS5, acting by opposite
shifts on ψ. This is precisely the same combination of generators discussed in section 2
that is preserved by a half-BPS surface operator OΣ in N = 4 SYM.
The radius of curvature of AdS5×S5 in (3.14) is given by R4 = y20 . Therefore using
that R4 = 4πNl4p, where N is the rank of the N = 4 YM theory, we have that
4πl4p
, (3.15)
and the position of the “charge” in y gets identified with the rank of the unbroken gauge
group and is therefore quantized.
The residue of the pole in Φ (3.10) gets identified with the position of the “charge” in
the ~x-plane. It follows from (3.1) that the coordinates ~x and y have dimensions of length2.
Therefore, we identify the residue of the pole of Φ with the position of the “charge” in the
~x-plane in X via:
(β0, γ0) =
2πl2s
. (3.16)
Unlike the position in y, the position in ~x is not quantized.
The remaining parameters of the surface operator OΣ with U(N) Levi group – given
by (α0, η0) – get identified with the holonomy of the two-forms of Type IIB supergravity
over D
α0 = −
, (3.17)
where D is the disk ending on the AdS5×S5 boundary on the S1. This identification
properly accounts for the correct periodicity of these parameters, which take values on a
circle of unit radius.
The path integral which defines a half-BPS surface operator OΣ when L = U(N)
is never singular as the gauge symmetry cannot be further enhanced by changing the
parameters (α0, β0, γ0, η0) of the surface operator. Correspondingly, the dual supergravity
solution with one “charge” also never acquires a singularity by changing the parameters
of the solution.
Let’s now consider the most general half-BPS surface operator OΣ. First we need
to characterize the operator by its Levi group, which for a U(N) gauge group takes the
l=1 U(Nl) with N =
l=1Nl = N . The operator then depends on 4M L-invariant
parameters (αl, βl, γl, ηl) for l = 1. . . . ,M up to the action of SM , which acts by permuting
the parameters.
The corresponding supergravity solution associated to such an operator is given by
the metric (3.1). The number of unbroken gauge group factors – given by the integer M –
corresponds to the number of point “charges” in (3.2). For M > 1, the metric that follows
from (3.3)(3.4) is AdS5×S5 only asymptotically and not globally.
The rank of the various gauge group factors in the Levi group
l=1 U(Nl) – given by
the integers Nl – correspond to the position of the “charges” along y ∈ X , given by the
coordinates yl. The precise identification follows from (3.5)(3.6):
4πl4p
l = 1, . . . ,M. (3.18)
Nl also corresponds to the amount of five-form flux over S
l , the S
5 associated with the
l-th point charge:
4π4l4p
F5 l = 1, . . . ,M. (3.19)
This identification quantizes the y coordinate in X into lp size bits. Thus length is quan-
tized as opposed to area, which is what happens for the geometry dual to the half-BPS
local operators [13][16], where it can be interpreted as the quantization of phase space in
the boundary gauge theory.
A half-BPS surface operator OΣ develops a pole for the scalar field Φ (2.7). The
pole is characterized by its residue, which is given by 2M real parameters (βl, γl). These
parameters are identified with the position of the M “charges” in the ~x-plane in X via:
(βl, γl) =
2πl2s
. (3.20)
All these parameters take values on the real line.
The remaining parameters characterizing a half-BPS surface operator OΣ are the
periodic variables (αl, ηl), which determine the holonomy produced by OΣ and the corre-
sponding two-dimensional θ-angles. These parameters get identified with the holonomy of
the two-forms of Type IIB supergravity over the M non-trivial disks Dl that the geometry
generates in the presence of M “charges” in X :
αl = −
(3.21)
The identification respects the periodicity of (αl, ηl), which in supergravity arises from the
invariance of BNS and BR under large gauge transformations
We have given a complete dictionary between all the parameters that a half-BPS
surface operator in N = 4 SYM depends on and all the parameters in the corresponding
bubbling geometry. We note that a surface operator OΣ depends on a set of parameters
up to the action of the permutation group SM on the parameters, which is part of the
U(N) gauge symmetry. The corresponding statement in supergravity is that the solution
dual to a surface operator is invariant under the action of SM , which acts by permuting
the “charges” in X .
20 Since the gauge invariant variables are e
The supergravity solution is regular as long as the “charges” do not collide. A singu-
larity arises whenever two point “charges” in X coincide (see Figure 1):
(~xl, yl) → (~xm, ym) for l 6= m. (3.22)
Whenever this occurs, there is a reduction in the number of independent disks since (see
Figure 3):
Dl → Dm for l 6= m, (3.23)
and therefore
for l 6= m.
(3.24)
In this limit of parameter space the non-trivial S2 connecting the points Pl and Pm in X
shrinks to zero size as [S2l,m] = [Dl]− [Dm] → 0, and the geometry becomes singular.
By using the dictionary developed in this paper, such a singular geometry corresponds
to a limit when two of each of the set of parameters (αl, βl, γl, ηl) defining a half-BPS surface
operator OΣ become equal:
αl → αm, βl → βm, γl → γm, ηl → ηm for l 6= m. (3.25)
In this limit the unbroken gauge group preserved by the surface operator OΣ is enhanced
to L′ from the original Levi group
l=1 U(Nl), where L ⊂ L′. As explained in section 2
the path integral from which OΣ is defined becomes singular.
In summary, we have found the description of all half-BPS surface operators OΣ in
N = 4 SYM in terms of solutions of Type IIB supergravity. The asymptotically AdS5×S5
solutions are regular and when they develop a singularity then the corresponding operator
also becomes singular.
S-Duality of Surface Operators from Type IIB Supergravity
The group of dualities of N = 4 SYM acts non-trivially [4] on surface operators OΣ
(see discussion in section 2). For G = U(N) the duality group is SL(2, Z) and its proposed
action on the parameters on which OΣ depends on is [4]:
(βl, γl) → |cτ + d| (βl, γl)
(αl, ηl) → (αl, ηl)M−1,
(3.26)
where M is an SL(2, Z) matrix
. (3.27)
We now reproduce21 this transformation law by studying the action of the SL(2, Z)
subgroup of the SL(2, R) classical symmetry of Type IIB supergravity, which is in fact the
appropriate symmetry group of Type IIB string theory. For that we need to analyze the
action of S-duality on our bubbling geometries.
SL(2, Z) acts on the complex scalar τ = C0 + ie
−φ of Type IIB supergravity in the
familiar fashion
τ → aτ + b
cτ + d
, (3.28)
where as usual τ gets identified with the complexified coupling constant of N = 4 SYM
(2.8). SL(2, Z) also rotates the two-form gauge fields22 of Type IIB supergravity
, (3.29)
while leaving the metric in the Einstein frame and the five-form flux invariant.
Given that the metric in (3.1) is in the Einstein frame, SL(2, Z) acts trivially on the
coordinates (~x, y). Nevertheless, since
ls = g
s lp with gs = e
φ (3.30)
the string scale transforms under SL(2, Z) as follows:
l2s →
|cτ + d| . (3.31)
Therefore, under S-duality:
2πl2s
→ |cτ + d| ~xl
2πl2s
. (3.32)
Given our dictionary in (3.20), we find that the surface operator parameters (βl, γl) trans-
form as in (3.26), agreeing with the proposal in [4].
21 If we apply the same idea to the LLM geometries dual to half-BPS local operators in [13],
we conclude that the half-BPS local operators are invariant under S-duality.
22 See e.g. [26][27].
The identification of the rest of the parameters is (3.21):
αl = −
(3.33)
Using the action of SL(2, Z) on the two-forms (3.29) and the identification (3.33), it
follows from a straightforward manipulation that the surface operator paramaters (αl, ηl)
transform as in (3.26), agreeing with the proposal in [4].
Acknowledgements
We would like to thank Xiao Liu for very useful discussions. Research at Perimeter In-
stitute for Theoretical Physics is supported in part by the Government of Canada through
NSERC and by the Province of Ontario through MRI. We also acknowledge further sup-
port from an NSERC Discovery Grant. SM acknowledges support from JSPS Research
Fellowships for Young Scientists.
Appendix A. Supersymmetry of Surface Operator in N=4 SYM
In this Appendix we study the Poincare and conformal supersymmetries preserved
by a surface operator in N=4 SYM supported on R1,1. These symmetries are generated
by ten dimensional Majorana-Weyl spinors ǫ1 and ǫ2 of opposite chirality. We determine
the supersymmetries left unbroken by a surface operator by studying the supersymmetry
variation of the gaugino in the presence of the surface operator singularity in (2.2)(2.7).
The metric is given by:
ds2 = −(dx0)2 + (dx1)2 + (dx2)2 + (dx3)2 = −(dx0)2 + (dx1)2 + 2dzdz̄. (A.1)
where z = 1√
(x2 + ix3) = |z|eiθ, while the singularity in the fields is
Φω = Φω̄ =
(Φ8 + iΦ9) =
β + iγ√
A = αdθ, F = dA = 2παδD,
(A.2)
where [α, β] = [α, γ] = [β, γ] = 0 and δD = d(dθ) is a two form delta function. The relevant
Γ-matrices are
(Γ2 − iΓ3)
Γz̄ =
(Γ2 + iΓ3)
{Γz,Γz̄} = 2.
(A.3)
A Poincare supersymmetry transformation is given by
δλ = (
µν +∇µΦiΓµi +
[Φi,Φj]Γ
ij)ǫ1 (A.4)
where µ runs from 0 to 3 and i runs from 4 to 9, while a superconformal supersymmetry
transformation is given by
δλ = [(
µν +∇µΦiΓµi +
[Φi,Φj]Γ
ij)xσΓσ − 2ΦiΓi]ǫ2 (A.5)
From (A.4), it follows that the unbroken Poincare supersymmetries are given by:
Γω̄zǫ1 = 0 ⇔ Γ2389ǫ1 = −ǫ1. (A.6)
The unbroken superconformal supersymmetries are given by:
−β + iγ
Γzω̄ − β − iγ
0 + x1Γ
1 + zΓz̄ + z̄Γz)− 2ΦωΓω − 2Φω̄Γω̄
ǫ2 = 0.
(A.7)
From the terms proportional to x0 and x1, we find that the unbroken superconformal
supersymmetries are given by:
Γzω̄ǫ2 = 0 ⇔ Γ2389ǫ2 = −ǫ2. (A.8)
The rest of the conditions
[Γz̄Γω̄Γz]ǫ2 = 0, (A.9)
are automatically satisfied once (A.8) is imposed.
We conclude that the singularity (2.2)(2.7) is half-BPS and that the preserved su-
persymmetry is generated by ǫ1 and ǫ2 subject o the constraints Γ2389ǫ1 = −ǫ1 and
Γ2389ǫ2 = −ǫ2.
By acting with a broken special conformal transformation on Σ = R2 ⊂ R4 to get a
surface operator supported on Σ =S2, one can show following [28] that such an operator
also preserves half of the thirty-two supersymmetries, but are now generated by a linear
combination of the Poincare and special conformal supersymmetries.
Appendix B. Five form flux
In this Appendix, we calculate (3.5) explicitly to evaluate the flux over a non-trivial
The five-form flux is [13][16]
{d[y2 2z + 1
2z − 1(dχ+ V )]− y
3 ∗3 d(
z + 1
)} ∧ dV olAdS3
{d[y2 2z − 1
2z + 1
(dχ+ V )]− y3 ∗3 d(
z − 1
)} ∧ dΩ3
(B.1)
The five-cycle S5l in the bubbling geometry is spanned by coordinates Ω3, χ and y.
Then the integration is:
4π4l4p
F5 = −
16π4l4p
2z − 1
2z + 1
dχ] ∧ dΩ3 =
4πl4p
= N. (B.2)
References
[1] J. M. Maldacena, “The large N limit of superconformal field theories and supergrav-
ity,” Adv. Theor. Math. Phys. 2, 231 (1998) [Int. J. Theor. Phys. 38, 1113 (1999)]
[arXiv:hep-th/9711200].
[2] S. S. Gubser, I. R. Klebanov and A. M. Polyakov, “Gauge theory correlators from
non-critical string theory,” Phys. Lett. B 428, 105 (1998) [arXiv:hep-th/9802109].
[3] E. Witten, “Anti-de Sitter space and holography,” Adv. Theor. Math. Phys. 2, 253
(1998) [arXiv:hep-th/9802150].
[4] S. Gukov and E. Witten, “Gauge theory, ramification, and the geometric langlands
program,” arXiv:hep-th/0612073.
[5] J. Gomis and F. Passerini, “Holographic Wilson loops,” JHEP 0608, 074 (2006)
[arXiv:hep-th/0604007].
[6] S. J. Rey and J. T. Yee, “Macroscopic strings as heavy quarks in large N gauge
theory and anti-de Sitter supergravity,” Eur. Phys. J. C 22, 379 (2001) [arXiv:hep-
th/9803001].
[7] J. M. Maldacena, “Wilson loops in large N field theories,” Phys. Rev. Lett. 80, 4859
(1998) [arXiv:hep-th/9803002].
[8] N. Drukker and B. Fiol, JHEP 0502, 010 (2005) [arXiv:hep-th/0501109].
[9] S. Yamaguchi, “Bubbling geometries for half BPSWilson lines,”arXiv:hep-th/0601089.
[10] S. Yamaguchi, “Wilson loops of anti-symmetric representation and D5-branes,” JHEP
0605, 037 (2006) [arXiv:hep-th/0603208].
[11] O. Lunin, “On gravitational description of Wilson lines,” JHEP 0606, 026 (2006)
[arXiv:hep-th/0604133].
[12] J. Gomis and F. Passerini, “Wilson loops as D3-branes,” JHEP 0701, 097 (2007)
[arXiv:hep-th/0612022].
[13] H. Lin, O. Lunin and J. M. Maldacena, “Bubbling AdS space and 1/2 BPS geome-
tries,” JHEP 0410, 025 (2004) [arXiv:hep-th/0409174].
[14] J. Gomis and C. Romelsberger, “Bubbling defect CFT’s,” JHEP 0608, 050 (2006)
[arXiv:hep-th/0604155].
[15] J. Gomis and T. Okuda, “Wilson loops, geometric transitions and bubbling Calabi-
Yau’s,” JHEP 0702, 083 (2007) [arXiv:hep-th/0612190].
[16] H. Lin and J. M. Maldacena, “Fivebranes from gauge theory,” Phys. Rev. D 74,
084014 (2006) [arXiv:hep-th/0509235].
[17] A. Kapustin and E. Witten, “Electric-magnetic duality and the geometric Langlands
program,” arXiv:hep-th/0604151.
[18] E. Frenkel, “Lectures on the Langlands program and conformal field theory,”
arXiv:hep-th/0512172.
http://arxiv.org/abs/hep-th/9711200
http://arxiv.org/abs/hep-th/9802109
http://arxiv.org/abs/hep-th/9802150
http://arxiv.org/abs/hep-th/0612073
http://arxiv.org/abs/hep-th/0604007
http://arxiv.org/abs/hep-th/9803001
http://arxiv.org/abs/hep-th/9803001
http://arxiv.org/abs/hep-th/9803002
http://arxiv.org/abs/hep-th/0501109
http://arxiv.org/abs/hep-th/0601089
http://arxiv.org/abs/hep-th/0603208
http://arxiv.org/abs/hep-th/0604133
http://arxiv.org/abs/hep-th/0612022
http://arxiv.org/abs/hep-th/0409174
http://arxiv.org/abs/hep-th/0604155
http://arxiv.org/abs/hep-th/0612190
http://arxiv.org/abs/hep-th/0509235
http://arxiv.org/abs/hep-th/0604151
http://arxiv.org/abs/hep-th/0512172
[19] R. M. Rohm, “Some Current Problems In Particle Physics Beyond The Standard
Model,”
[20] J. Preskill and L. M. Krauss, “Local Discrete Symmetry And Quantum Mechanical
Hair,” Nucl. Phys. B 341, 50 (1990).
[21] M. G. Alford, K. M. Lee, J. March-Russell and J. Preskill, Nucl. Phys. B 384, 251
(1992) [arXiv:hep-th/9112038].
[22] A. Kapustin and S. Sethi, “The Higgs branch of impurity theories,” Adv. Theor. Math.
Phys. 2, 571 (1998) [arXiv:hep-th/9804027].
[23] N. R. Constable, J. Erdmenger, Z. Guralnik and I. Kirsch, “Intersecting D3-branes
and holography,” Phys. Rev. D 68, 106007 (2003) [arXiv:hep-th/0211222].
[24] A. Kapustin, “Wilson-’t Hooft operators in four-dimensional gauge theories and Phys.
Rev. D 74, 025005 (2006) [arXiv:hep-th/0501015].
[25] E. Witten, “Anti-de Sitter space, thermal phase transition, and confinement in gauge
Adv. Theor. Math. Phys. 2, 505 (1998) [arXiv:hep-th/9803131].
[26] J. Polchinski, “String Theory”, Cambridge University Press, Chapter 12 in Volume 2.
[27] K. Becker, M. Becker and J. Schwarz, ”String Theory and M-theory”, Cambridge
University Press, Chapter 8.
[28] M. Bianchi, M. B. Green and S. Kovacs, “Instanton corrections to circular Wilson
loops in N = 4 supersymmetric JHEP 0204, 040 (2002) [arXiv:hep-th/0202003].
http://arxiv.org/abs/hep-th/9112038
http://arxiv.org/abs/hep-th/9804027
http://arxiv.org/abs/hep-th/0211222
http://arxiv.org/abs/hep-th/0501015
http://arxiv.org/abs/hep-th/9803131
http://arxiv.org/abs/hep-th/0202003

  We construct smooth asymptotically AdS_5xS^5 solutions of Type IIB
supergravity corresponding to all the half-BPS surface operators in N=4 SYM.
All the parameters labeling a half-BPS surface operator are identified in the
corresponding bubbling geometry. We use the supergravity description of surface
operators to study the action of the SL(2,Z) duality group of N=4 SYM on the
parameters of the surface operator, and find that it coincides with the recent
proposal by Gukov and Witten in the framework of the gauge theory approach to
the geometrical Langlands with ramification. We also show that whenever a
bubbling geometry becomes singular that the path integral description of the
corresponding surface operator also becomes singular.

Introduction and Summary
Gauge invariant operators play a central role in the gauge theory holographically
describing quantum gravity with AdS boundary conditions [1][2][3], as correlation functions
of gauge invariant operators are the only observables in the boundary gauge theory. Finding
the bulk description of all gauge invariant operators is necessary in order to be able to
formulate an arbitrary bulk experiment in terms of gauge theory variables.
In this paper we provide the bulk description of a novel class of half-BPS operators
in N = 4 SYM which are supported on a surface Σ [4]. These nonlocal surface operators
OΣ are defined by quantizing N = 4 SYM in the presence of a certain codimension two
singularity for the classical fields of N = 4 SYM. The singularity characterizing such a
surface operator OΣ depends on 4M real parameters, whereM is the number of U(1)’s left
unbroken byOΣ. Surface operators are a higher dimensional generalization of Wilson and ’t
Hooft operators, which are supported on curves and induce a codimension three singularity
for the classical fields appearing in the Lagrangian. In this paper we extend the bulk
description of all half-BPS Wilson loop operators found in [5] (see also1 [8][9][10][11][12])
to all half-BPS surface operators.
We find the asymptotically AdS5×S5 solutions of Type IIB supergravity corresponding
to all half-BPS surface operators OΣ in N = 4 U(N) SYM. The topology and geometry
of the “bubbling” solution is completely determined in terms of some data, very much like
in the case studied by Lin, Lunin and Maldacena (LLM) in the context of half-BPS local
operators [13]2. In fact, we identify the system of equations determining the supergravity
solution corresponding to the half-BPS surface operators inN = 4 SYM with that obtained
by “analytic” continuation of the LLM equations [13][16].
The data determining the topology and geometry of a supergravity solution is charac-
terized by the position of a collection of M point particles in a three dimensional space X ,
whereX is a submanifold of the ten dimensional geometry. Different particle configurations
give rise to different asymptotically AdS5×S5 geometries.
1 The description of Wilson loops in the fundamental representation goes back to [6][7].
2 The bubbling geometry description of half-BPS Wilson loops was found in [10][11] while that
of half-BPS domain wall operators was found in [11][14]. For the bubbling Calabi-Yau geometries
for Wilson loops in Chern-Simons, see [15].
Fig. 1: a) The metric and five-form flux is determined once the position of the
particles in X – labeled by coordinates (~xl, yl) where y ≥ 0 – is given. The l-th
particle is associated with a point Pl ∈ X. b) The configuration corresponding to
the AdS5×S
5 vacuum.
Even though the choice of a particle distribution in X completely determines the
geometry and topology of the metric and the corresponding RR five-form field strength,
further choices have to be made to fully characterize a solution of Type IIB supergravity
on this geometry3.
Given a configuration ofM particles inX , the corresponding ten dimensional geometry
developsM non-trivial disks which end on the boundary4 of AdS5×S5 on a non-contractible
S1. Since Type IIB supergravity has two two-form gauge fields, one from the NS-NS sector
and one from the RR sector, a solution of the Type IIB supergravity equations of motion
is fully determined only once the holonomy of the two-forms around the various disks is
specified:
l = 1, . . . ,M. (1.1)
Therefore, an asymptotically AdS5×S5 solution depends on the position of theM particles
in X – given by (~xl, yl) – and on the holonomies of the two-forms (1.1).
A precise dictionary is given between all the 4M parameters that label a half-BPS
surface operator OΣ and all the parameters describing the corresponding supergravity
solution. We show that the supergravity solution describing a half-BPS surface operator is
regular and that whenever the supergravity solution develops a singularity the N = 4 SYM
path integral description of the corresponding surface operator also develops a singularity.
We study the action of the SL(2, Z) symmetry of Type IIB string theory on the
supergravity solutions representing the half-BPS surface operators in N = 4 SYM. By
3 This is on top of the obvious choice of dilaton and axion, which gets identified with the
complexified coupling constant in N = 4 SYM.
4 The conformal boundary in this case is AdS3×S
1, where surface operators in N = 4 SYM
can be studied by specifying non-trivial boundary conditions.
using the proposed dictionary between the parameters of a supergravity solution and the
parameters of the corresponding surface operator, we can show that the action of S-duality
induced on the parameters of a surface operator coincides with the recent proposal by
Gukov and Witten [4] in the framework of the gauge theory approach to the geometrical
Langlands [17]5 with ramification.
Whether surface operators can serve as novel order parameters in gauge theory remains
an important open question. It is our hope that the viewpoint on these operators provided
by the supergravity solutions in this paper may help shed light on this crucial question.
The plan of the rest of the paper is as follows. In section 2 we study the gauge theory
singularities corresponding to surface operators in N = 4 SYM, study the symmetries
preserved by a half-BPS surface operator and review the proposal in [4] for the action of
S-duality on the parameters that a half-BPS surface operator depends on. We also compute
the scaling weight of these operators and show that it is invariant under Montonen-Olive
duality. In section 3 we construct the solutions of Type IIB supergravity describing the
half-BPS surface operators. We identify all the parameters that a surface operator depends
on in the supergravity solution and show that the action of S-duality on surface operators
proposed in [4] follows from the action of SL(2, Z) on the classical solutions of supergravity.
The Appendices contain some details omitted in the main text.
2. Surface Operators in Gauge Theories
A surface operator OΣ is labeled by a surface Σ in R1,3 and by a conjugacy class U of
the gauge group G. The data that characterizes a surface operator OΣ, the surface Σ and
the conjugacy class U , can be identified with that of an external string used to probe the
theory. The surface Σ corresponds to the worldsheet of a string while the conjugacy class
U is associated to the Aharonov-Bohm phase acquired by a charged particle encircling the
string.
The singularity6 in the gauge field produced by a surface operator is that of a non-
abelian vortex. This singularity in the gauge field can be characterized by the phase
5 See e.g. [18] for a review of the geometric Langlands program.
6 Previous work involving codimension two singularities in gauge theory include [19][20][21].
acquired by a charged particle circumnavigating around the string. This gives rise to a
group element7 U ⊂ U(N)
U ≡ P exp i
A ⊂ U(N), (2.1)
which corresponds to the Aharonov-Bohm phase picked up by the wavefunction of the
charged particle. Since gauge transformations act by conjugation U → gUg−1, a surface
operator is labeled by a conjugacy class of the gauge group.
By performing a gauge transformation, the matrix U can be diagonalized. If we
demand that the gauge field configuration is scale invariant – so that OΣ has a well defined
scaling weight – then the gauge field produced by a surface operator can then be written
α1 ⊗ 1N1 0 . . . 0
0 α2 ⊗ 1N2 . . . 0
. . .
0 0 . . . αM ⊗ 1NM
dθ, (2.2)
where θ is the polar angle in the R2 ⊂ R1,3 plane normal to Σ and 1n is the n-dimensional
unit matrix. We note that the matrix U takes values on the maximal torus TN = RN/ZN
of the U(N) gauge group. Therefore the parameters αi take values on a circle of unit
radius.
The surface operator corresponding to (2.2) spontaneously breaks the U(N) gauge
symmetry along Σ down to the so called Levi group L, where a group of Levi type is char-
acterized by the subgroup of U(N) that commutes with (2.2). Therefore, L =
l=1 U(Nl),
where N =
l=1Nl.
Since the gauge group is broken down to the Levi group L =
l=1 U(Nl) along Σ,
there is a further choice [4] in the definition of OΣ consistent with the symmetries and
equations of motion. This corresponds to turning on a two dimensional θ-angle for the
unbroken U(1)’s along the string worldsheet Σ. The associated operator insertion into the
N = 4 SYM path integral is given by:
Tr Fl
. (2.3)
7 We now focus on G = U(N) as it is the relevant gauge group for describing string theory
with asymptotically AdS5×S
5 boundary conditions.
The parameters ηi takes values in the maximal torus of the S-dual or Langlands dual
gauge group LG [4]. Therefore, since LG = U(N) for G = U(N), we have that the matrix
of θ-angles of a surface operator OΣ characterized by the Levi group L =
l=1 U(Nl) is
given by the L-invariant matrix:
η1 ⊗ 1N1 0 . . . 0
0 η2 ⊗ 1N2 . . . 0
. . .
0 0 . . . ηM ⊗ 1NM
. (2.4)
The parameters ηi, being two dimensional θ-angles, also take values on a circle of unit
radius.
Therefore, a surface operator OΣ in pure gauge theory with Levi group L =
l=1 U(Nl) is labeled by 2M L-invariant parameters (αl, ηl) up to the action of SM ,
which acts by permuting the different eigenvalues in (2.2) and (2.4). The operator is then
defined by expanding the path integral with the insertion of the operator (2.3) around
the singularity (2.2), and by integrating over connections that are smooth near Σ. In
performing the path integral, we must divide [4] by the gauge transformations that take
values in L =
l=1 U(Nl) when restricted to Σ. This means that the operator becomes
singular whenever the unbroken gauge symmetry near Σ gets enhanced, corresponding to
when eigenvalues in (2.2) and (2.4) coincide.
Surface Operators in N = 4 SYM
In a gauge theory with extra classical fields like N = 4 SYM, the surface operator OΣ
may produce a singularity for the extra fields near the location of the surface operator.
The only requirement is that the singular field configuration solves the equations of motion
of the theory away8 from the surface Σ. The global symmetries imposed on the operator
OΣ determine which classical fields in the Lagrangian develop a singularity near Σ together
with the type of singularity.
A complementary viewpoint on surface operators is to add new degrees of freedom on
the surface Σ. Such an approach to surface operators in N = 4 SYM has been considered
8 For pure gauge theory, the field configuration in (2.2) does satisfy the Yang-Mills equation
of motion DmF
mn = 0 away from Σ. Moreover, adding the two dimensional θ-angles (2.3) does
not change the equations of motion.
in [22][23] where the new degrees of freedom arise from localized open strings on a brane
intersection.
The basic effect of OΣ is to generate an Aharonov-Bohm phase corresponding to a
group element U (2.1). If we let z be the complex coordinate in the R2 ⊂ R1,3 plane
normal to Σ, the singularity in the gauge field configuration is then given by
, (2.5)
where AI are constant matrices. Scale invariance of the singularity – which we are going
to impose – restricts AI = 0 for I ≥ 2.
The operator OΣ can also excite a complex scalar field Φ of N = 4 SYM near Σ
while preserving half of the Poincare supersymmetries of N = 4 SYM. Imposing that the
singularity is scale invariant9 yields
, (2.6)
where Φ1 is a constant matrix.
A surface operator OΣ is characterized by the choice of an unbroken gauge group
L ⊂ G along Σ. Correspondingly, the singularity of all the fields excited by OΣ must be
invariant under the unbroken gauge group L. For L =
l=1 U(Nl) ⊂ U(N) the singularity
in the gauge field is the non-abelian vortex configuration in (2.2) and the two dimensional
θ-angles are given by (2.4). L-invariance together with scale invariance requires that Φ
develops an L-invariant pole near Σ:
β1 + iγ1 ⊗ 1N1 0 . . . 0
0 β2 + iγ2 ⊗ 1N2 . . . 0
. . .
0 0 . . . βM + iγM ⊗ 1NM
. (2.7)
Therefore, a half-BPS surface operator OΣ in N = 4 SYM with Levi group L =
l=1 U(Nl) is labeled by 4M L-invariant parameters (αl, βl, γl, ηl) up to the action of SM ,
which permutes the different eigenvalues in (2.2)(2.4)(2.7). The operator is defined by the
9 If we relax the restriction of scale invariance, one can then get other supersymmetric singu-
larities with higher order poles Φ =
and A (2.5). The surface operators associated with
these singularities may be relevant [4] for the gauge theory approach to the study of the geometric
Langlands program with wild ramification.
path integral of N = 4 SYM with the insertion of the operator (2.3) expanded around the
L-invariant singularities (2.2)(2.7) and by integrating over smooth fields near Σ. As in the
pure gauge theory case, we must mode out by gauge transformations that take values in
L ⊂ U(N) when restricted to Σ. The surface operator OΣ becomes singular whenever the
the parameters that label the surface operator (αl, βl, γl, ηl) for l = 1, . . . ,M are such that
they are invariant under a larger symmetry than L, the group of gauge transformations
we have to mode out when evaluating the path integral.
S-duality of Surface Operators
In N = 4 SYM the coupling constant combines with the four dimensional θ-angle into
a complex parameter taking values in the upper half-plane:
. (2.8)
The group of duality symmetries ofN = 4 SYM is an infinite discrete subgroup of SL(2, R),
which depends on the gauge group G. For N = 4 SYM with G = U(N) the relevant
symmetry group is SL(2, Z):
∈ SL(2, Z). (2.9)
Under S-duality τ → −1/τ and G gets mapped10 to the S-dual or Langlands dual gauge
group LG. For G = U(N) the S-dual group is LG = U(N), and SL(2, Z) is a symmetry of
the theory, which acts on the coupling of the theory by fractional linear transformations:
τ → aτ + b
cτ + d
. (2.10)
In [4], Gukov and Witten made a proposal of how S-duality acts on the parameters
(αl, βl, γl, ηl) labeling a half-BPS surface operator. The proposed action is given by [4]:
(βl, γl) → |cτ + d| (βl, γl)
(αl, ηl) → (αl, ηl)M−1.
(2.11)
10 For G not a simply-laced group, τ → −1/nτ , where n is the ratio of the length-squared of
the long and short roots of G.
With the aid of this proposal, it was shown in [4] that the gauge theory approach to
the geometric Langlands program pioneered in [17] naturally extends to the geometric
Langlands program with tame ramification.
Symmetries of half-BPS Surface Operators in N = 4 SYM
We now describe the unbroken symmetries of the half-BPS surface operators OΣ.
These symmetries play an important role in determining the gravitational dual description
of these operators, which we provide in the next section.
In the absence of any insertions, N = 4 SYM is invariant under the PSU(2, 2|4)
symmetry group. If we consider the surface Σ = R1,1 ⊂ R1,3, then Σ breaks the SO(2, 4)
conformal group to a subgroup. A surface operator OΣ supported on this surface inserts
into the gauge theory a static probe string. This surface is manifestly invariant under
rotations and translations in Σ and scale transformations. It is also invariant under the
action of inversion I : xµ → xµ/x2 and consequently11 invariant under special conformal
transformations in Σ. Therefore, the symmetries left unbroken by Σ = R1,1 generate
an SO(2, 2)× SO(2)23 subgroup of the SO(2, 4) conformal group, where SO(2)23 rotates
the plane transverse to Σ in R1,3. In Euclidean signature, the surface Σ =S2 preserves
an SO(1, 3) × SO(2)23 subgroup of the Euclidean conformal group. This surface can be
obtained from the surface Σ = R2 ∈ R4 by the action of a broken special conformal
generator and can also be used to construct a half-BPS surface operator OΣ in N = 4
Since the symmetry of a surface operator with Σ = R1,1 is SO(2, 2) × SO(2)23 one
can study such an operator either by considering the gauge theory in R1,3 or in AdS3×S1,
which can be obtained from R1,3 by a conformal transformation. Studying the gauge
theory in AdS3×S1 has the advantage of making the symmetries of the surface operator
manifest, as the conformal symmetries left unbroken by the surface act by isometries on
AdS3×S1. Surface operators in R1,3 are described by a codimension two singularity while
surface operators in AdS3×S1 are described by a boundary condition on the boundary of
AdS3. A surface operator with Σ = R
1,1 corresponds to a boundary condition on AdS3
in Poincare coordinates while a surface operator on Σ =S2 corresponds to a boundary
condition on global Euclidean AdS3.
11 We recall that a special conformal transformation Kµ is generated by IPµI, where Pµ is the
translation generator and I is an inversion.
The singularity in the classical fields produced by OΣ in (2.2)(2.7) is also invariant
under SO(2, 2). The N = 4 scalar field Φ carries charge under an SO(2)R subgroup of the
SO(6) R-symmetry and is therefore SO(4) invariant. The surface operator OΣ is therefore
invariant under SO(2, 2)×SO(2)a×SO(4), where SO(2)a is generated by the anti-diagonal
product12 of SO(2)23 × SO(2)R.
N = 4 SYM has sixteen Poincare supersymmetries and sixteen conformal super-
symmetries, generated by ten dimensional Majorana-Weyl spinors ǫ1 and ǫ2 of opposite
chirality. As shown in the Appendix A, the surface operator OΣ for Σ = R1,1 preserves
half of the Poincare and half of the conformal supesymmetries13 and is therefore half-BPS.
With the aid of these symmetries we study in the next section the gravitational de-
scription of half-BPS surface operators in N = 4 SYM.
Scaling Weight of half-BPS Surface Operators in N = 4 SYM
Conformal symmetry constraints the form of the OPE of the energy-energy tensor
Tmn with the operators in the theory. For a surface operator OΣ supported on Σ = R1,1,
SO(2, 2)× SO(2)23 invariance completely fixes the OPE of Tmn with OΣ:
< Tµν(x)OΣ >
< OΣ >
< Tij(x)OΣ >
< OΣ >
[4ninj − 3δij ] ; < Tµi(x)OΣ >= 0.
(2.12)
Here xm = (xµ, xi), where xµ are coordinates along Σ and ni = xi/r, and r is the radial
coordinate in the R2 transverse to R1,1. h is the scaling weight of OΣ, which generalizes
[24] the notion of conformal dimension of local conformal fields to surface operators.
In order to calculate the scaling dimension of a half-BPS surface operator OΣ in N = 4
SYM we evaluate the classical field configuration (2.2)(2.4)(2.7) characterizing a half-BPS
surface operator on the classical energy-momentum tensor of N = 4 SYM:
Tmn =
Tr[DmφDnφ−
gmn(Dφ)
2 − 1
(DmDn − gmnD2)φ2]
Tr[−FmlFnl +
gmnFlpFlp].
(2.13)
12 Since SO(2)a leaves Φ · z in (2.7) invariant.
13 For Σ =S2, the operator is also half-BPS, but it preserves a linear combination of Poincare
and special conformal supersymmetries.
A straightforward computation14 leads to:
h = − 2
l + γ
l ) = −
i + γ
i ). (2.14)
The action of an SL(2, Z) transformation (2.10) on the coupling constant of N = 4
SYM implies that:
Imτ → Imτ|cτ + d|2 . (2.15)
Combining this with the action (2.11) of Montonen-Olive duality on the parameters of the
surface operator, we find that the scaling weight (2.14) of a half-BPS surface operator OΣ
is invariant under S-duality:
h→ h. (2.16)
In this respect half-BPS surface operators behave like the half-BPS local operators of
N = 4 SYM, whose conformal dimension is invariant under SL(2, Z), and unlike the
half-BPS Wilson-’t Hooft operators whose scaling weight is not S-duality invariant [24].
3. Bubbling Surface Operators
In this section we find the dual gravitational description of the half-BPS surface op-
erators OΣ described in the previous section. The bulk description is given in terms of
asymptotically AdS5×S5 and singularity free solutions of the Type IIB supergravity equa-
tions of motion. The data from which the solution is uniquely determined encodes the
corresponding data about the surface operator OΣ.
The strategy to obtain these solutions is to make an ansatz for Type IIB supergravity
which is invariant under all the symmetries preserved by the half-BPS surface operators
OΣ. As discussed in the previous section, the bosonic symmetries preserved by a half-
BPS surface operator OΣ are SO(2, 2) × SO(4) × SO(2)a. Therefore the most general
ten dimensional metric invariant under these symmetries can be constructed by fibering
AdS3×S3×S1 over a three manifold X , where the symmetries act by isometries on the
fiber. The constraints imposed by unbroken supersymmetry on the ansatz are obtained by
demanding that the ansatz for the supergravity background possesses a sixteen component
14 Contact terms depending on α, β, γ and proportional to the derivative of the two-dimensional
δ-function appear when evaluating the on-shelll energy-momentum tensor. It would be interesting
to understand the physical content of these contact terms.
Killing spinor, which means that the background solves the Killing spinor equations of Type
IIB supergravity. A solution of the Killing spinor equations and the Bianchi identity for the
five-form field strength guarantee that the full set of equations of Type IIB supergravity
are satisfied and that a half-BPS solution has been obtained.
The problem of solving the Killing spinor equations of Type IIB supergravity with
an SO(2, 2)× SO(4)× SO(2)a symmetry can be obtained by analytic continuation of the
equations studied by LLM [13][16], which found the supergravity solutions describing the
half-BPS local operators of N = 4 SYM, which have an SO(4)×SO(4)×R symmetry. The
equations determining the metric and five-form flux can be read from [13][16], in which the
analytic continuation that we need to construct the gravitational description of half-BPS
surface operators OΣ was considered.
The ten dimensional metric and five-form flux is completely determined in terms of
data that needs to be specified on the three manifold X in the ten dimensional space. An
asymptotically AdS5×S5 metric is uniquely determined in terms of a function z(x1, x2, y),
where (x1, x2, y) ≡ (~x, y) are coordinates in X . The ten dimensional metric in the Einstein
frame is given by15
ds2 = y
2z + 1
2z − 1ds
2z − 1
2z + 1
dΩ3 +
4z2 − 1
(dχ+ V )2 +
4z2 − 1
(dy2 + dxidxi),
(3.1)
where ds2X = dy
2+dxidxi with y ≥ 0 and V is a one-form in X satisfying dV = 1/y ∗X dz.
AdS3 in Poincare coordinate corresponds to a surface operator on Σ = R
1,1 while AdS3
in global Euclidean coordinate corresponds to a surface operator on Σ =S2. The U(1)a
symmetry acts by shifts on χ while SO(2, 2) and SO(4) act by isometries on the coordinates
of AdS3 and S
3 respectively.
A non-trivial solution to the equations of motion is obtained by specifying a config-
uration of M point-like particles in X . The data from which the solution is determined
is the “charge” Ql of the particles together with their positions (~xl, yl) in X (see Figure
1). Given a “charge” distribution, the function z(x1, x2, y) solves the following differential
equation:
∂i∂iz(x1, x2, y) + y∂y
∂yz(x1, x2, y)
Qlδ(y − yl)δ(2)(~x− ~xl). (3.2)
15 The “analytic” continuation from the bubbling geometries dual to the half-BPS local opera-
tors is given by z → z, t → χ, y → −iy, ~x → i~x, dΩ3 → −ds
[13][16].
Introducing a “charge” at the point (~xl, yl) in X has the effect of shrinking
16 the
S1 with coordinate χ in (3.1) to zero size at that point. In order for this to occur in a
smooth fashion the magnitude of the “charge” has to be fixed [13][16] so that Ql = 2πyl.
Therefore, the independent data characterizing the metric and five-form of the solution is
the position of the M “charges”, given by (~xl, yl).
In summary, a smooth half-BPS SO(2, 2)× SO(4)× SO(2)a invariant asymptotically
AdS5×S5 metric (3.1) solving the Type IIB supergravity equations of motion is found by
solving (3.2) subject to the boundary condition z(x1, x2, 0) = 1/2 [13][16], so that the S
in (3.1) shrinks in a smooth way at y = 0. The function z(x1, x2, y) is given by
z(x1, x2, y) =
zl(x1, x2, y), (3.3)
where
zl(x1, x2, y) =
(~x− ~xl)2 + y2 + y2l
((~x− ~xl)2 + y2 + y2l )2 − 4y2l y2
, (3.4)
and V can be computed from z(x1, x2, y) from dV = 1/y ∗X dz. Both the metric and
five-form field strength are determined by an integer M and by (~xl, yl) for l = 1, . . . ,M .
Topology of Bubbling Solutions And Two-form Holonomies
The asymptotically AdS5×S5 solutions constructed from (3.3)(3.4) are topologically
quite rich. In particular, a solution withM point “charges” hasM topologically non-trivial
S5’s. We can associate to each point Pl ∈ X a corresponding five-sphere S5l . S5l can be
constructed by fibering the S1×S3 in the geometry (3.1) over a straight line between the
point (~xl, 0) and the point (~xl, yl) in X . The topology of this manifold is indeed an S
as an S5 can be represented17 by an S1×S3 fibration over an interval where the S1 and S3
shrink to zero size at opposite ends of the interval, which is what happens in our geometry
where the S3 shrinks at (~xl, 0) while the S
1 shrinks at the other endpoint (~xl, yl).
16 Near y = yl the form of the relevant part of the metric is that of the Taub-NUT space. Fixing
the value of the “charge” at y = yl by imposing regularity of the metric coincides with the usual
regularity constraint on the periodicity of the circle in Taub-NUT space.
17 This can be seen explicitly by writing dΩ5 = cos
2 θdΩ3 + dθ
2 + sin2 θdφ2.
Fig. 2: A topologically non-trivial S5 can be constructed by fibering S1×S3 over
an interval connecting the y = 0 plane and the location of the “charge” at the
point Pl ∈ X with (~xl, yl) coordinates.
Following [13][16] we can now integrate the five-form flux over the topologically non-
trivial S5’s (see Appendix B):
F5 = y
l . (3.5)
Since flux has to be quantized, the position in the y-axis of the l-th particle in X is also
quantized
y2l = 4πNll
p Nl ∈ Z, (3.6)
where lp is the ten dimensional Planck length. For an asymptotically AdS5×S5 geometry
with radius of curvature R4 = 4πNl4p, which is dual to N = 4 U(N) SYM, we have that
the total amount of five-form flux must be N :
Nl. (3.7)
The asymptotically AdS5×S5 solutions constructed from (3.3)(3.4) also contain non-
trivial surfaces. In particular, a solution with M point “charges” has M non-trivial disks
Dl. Just as in the case of the S
5’s, we can associate to each point Pl ∈ X a disk Dl.
Inspection of the asymptotic form of the metric (3.1) given in (3.3)(3.4) reveals that
the metric is conformal to AdS3×S1. This geometry on the boundary of AdS5×S5, which
is where the dual N = 4 U(N) SYM lives, is the natural background geometry on which to
study conformally invariant surface operators in N = 4 SYM. As explained in section 2, an
SO(2, 2)× SO(2)23 invariant surface operator can be defined by specifying a codimension
two singularity in R1,3 or by specifying appropriate boundary conditions for the classical
fields in the gauge theory at the boundary of AdS3×S1. In the latter formulation, the
worldsheet of the surface operator Σ is the boundary of AdS3.
Therefore, in the boundary of AdS5×S5 we have a non-contractible S1. If we fiber the
S1 parametrized by χ in (3.1) over a straight line connecting a point (~xl, yl) in X – where
the S1 shrinks to zero size – to a point in X corresponding to the boundary of AdS5×S5
– given by ~x, y → ∞ – we obtain a surface Dl. This surface is topologically a disk18 and
there are M of them for a “charge” distribution of M particles in X .
Fig. 3: A disk D can be constructed by fibering S1 over an interval connecting
the “charge” at the point Pl ∈ X with (~xl, yl) coordinates and the boundary of
AdS5×S
Due to the existence of the disks Di, the supergravity solution given by the metric
and five-form flux alone is not unique. Type IIB supergravity has a two-form gauge field
from the NS-NS sector and another one from the RR sector. In order to fully specify a
solution of Type IIB supergravity in the bubbling geometry (3.1) we must complement the
metric and the five-form with the integral of the two-forms around the disks19
αl = −
l = 1, . . . ,M, (3.8)
where we have used notation conducive to the later comparison with the parameters char-
acterizing a half-BPS surface operator OΣ. Since both BNS and BR are invariant under
large gauge transformations, the parameters (αl, ηl) take values on a circle of unit radius.
Apart from the M disks Dl, the bubbling geometry constructed from (3.3)(3.4) also
has topologically non-trivial S2’s. One can construct an S2 by fibering the S1 in (3.1) over
18 Such disks also appear in the study of the high temperature regime of N = 4 SYM, where
the bulk geometry [25] is the AdS Schwarzschild black hole, which also has a non-contractible S1
in the boundary which is contractible in the full geometry.
19 The overall signs in the identification are fixed by demanding consistent action of S-duality
of N = 4 SYM with that of Type IIB supergravity.
a straight line connecting the points Pl and Pm in X . Since the S
1 shrinks to zero size in
a smooth manner at the endpoints we obtain an S2. Therefore, to every pair of “charges”
in X , characterized by different points Pl and Pm in X , we can construct a corresponding
S2, which we label by S2l,m. The integral of BNS and BR over S
l,m do not give rise to new
parameters, since [S2l,m] = [Dl] − [Dm] in homology, and the periods can be determined
from (3.8).
Fig. 4: An S2 can be constructed by fibering S1 over an interval connecting the
“charge” at the point Pl ∈ X with a different “charge” at point Pm ∈ X.
Bubbling Geometries as Surface Operators
As we discussed in section 2, a surface operator OΣ is characterized by an unbroken
gauge group L ∈ U(N) along together with 4M L-invariant parameters (αl, βl, γl, ηl). On
the other hand, the Type IIB supergravity solutions we have described depend on the
positions (~xl, yl) of M “charged” particles in X and the two-form holonomies:
. (3.9)
We now establish an explicit dictionary between the parameters in gauge theory and the
parameters in supergravity.
For illustration purposes, it is convenient to start by considering the half-BPS surface
operator OΣ with the largest Levi group L, which is L = U(N) for G = U(N). U(N)
invariance requires that the singularity in the fields produced by OΣ take values in the
center of U(N). Therefore, the gauge field and scalar field produced by OΣ is given by
A = α01Ndθ
(β0 + iγ0)1N ,
(3.10)
where 1N is the identity matrix. We can also turn a two-dimensional θ-angle (2.3) for the
overall U(1), so that:
η = η01N . (3.11)
We now identify this operator with the supergravity solution obtained by having a
single point “charge” source in X (see Figure 1b). If we let the position of the “charge”
be (~x0, y0) then
z(x1, x2, y) =
(~x− ~x0)2 + y2 + y20
((~x− ~x0)2 + y2 + y20)2 − 4y20y2
VI = −ǫIJ
(xJ − xJ0 )((~x− ~x0)2 + y2 − y20)
2(~x− ~x0)2
((~x− ~x0)2 + y2 + y20)2 − 4y20y2
(3.12)
where V = VIdx
I . The metric (3.1) obtained using (3.12) is the metric of AdS5×S5. This
can be seen by the following change of variables [16]
x1 − x10 + i(x2 − x20) = rei(ψ+φ)
r = y0 sinhu sin θ
y = y0 coshu cos θ
(ψ − φ),
(3.13)
which yields the AdS5×S5 metric with AdS5 foliated by AdS3×S1 slices:
ds2 = y0
(cosh2 uds2AdS3 + du
2 + sinh2 udψ2) + (cos2 θdΩ3 + dθ
2 + sin2 θdφ2)
. (3.14)
We note that the U(1)a symmetry of the metric (3.1) – which acts by shifts on χ – identifies
via (3.13) an SO(2)R subgroup of the the SO(6) symmetry of the S
5, acting by shifts on
φ, with an SO(2)23 subgroup of the SO(2, 4) isometry group of AdS5, acting by opposite
shifts on ψ. This is precisely the same combination of generators discussed in section 2
that is preserved by a half-BPS surface operator OΣ in N = 4 SYM.
The radius of curvature of AdS5×S5 in (3.14) is given by R4 = y20 . Therefore using
that R4 = 4πNl4p, where N is the rank of the N = 4 YM theory, we have that
4πl4p
, (3.15)
and the position of the “charge” in y gets identified with the rank of the unbroken gauge
group and is therefore quantized.
The residue of the pole in Φ (3.10) gets identified with the position of the “charge” in
the ~x-plane. It follows from (3.1) that the coordinates ~x and y have dimensions of length2.
Therefore, we identify the residue of the pole of Φ with the position of the “charge” in the
~x-plane in X via:
(β0, γ0) =
2πl2s
. (3.16)
Unlike the position in y, the position in ~x is not quantized.
The remaining parameters of the surface operator OΣ with U(N) Levi group – given
by (α0, η0) – get identified with the holonomy of the two-forms of Type IIB supergravity
over D
α0 = −
, (3.17)
where D is the disk ending on the AdS5×S5 boundary on the S1. This identification
properly accounts for the correct periodicity of these parameters, which take values on a
circle of unit radius.
The path integral which defines a half-BPS surface operator OΣ when L = U(N)
is never singular as the gauge symmetry cannot be further enhanced by changing the
parameters (α0, β0, γ0, η0) of the surface operator. Correspondingly, the dual supergravity
solution with one “charge” also never acquires a singularity by changing the parameters
of the solution.
Let’s now consider the most general half-BPS surface operator OΣ. First we need
to characterize the operator by its Levi group, which for a U(N) gauge group takes the
l=1 U(Nl) with N =
l=1Nl = N . The operator then depends on 4M L-invariant
parameters (αl, βl, γl, ηl) for l = 1. . . . ,M up to the action of SM , which acts by permuting
the parameters.
The corresponding supergravity solution associated to such an operator is given by
the metric (3.1). The number of unbroken gauge group factors – given by the integer M –
corresponds to the number of point “charges” in (3.2). For M > 1, the metric that follows
from (3.3)(3.4) is AdS5×S5 only asymptotically and not globally.
The rank of the various gauge group factors in the Levi group
l=1 U(Nl) – given by
the integers Nl – correspond to the position of the “charges” along y ∈ X , given by the
coordinates yl. The precise identification follows from (3.5)(3.6):
4πl4p
l = 1, . . . ,M. (3.18)
Nl also corresponds to the amount of five-form flux over S
l , the S
5 associated with the
l-th point charge:
4π4l4p
F5 l = 1, . . . ,M. (3.19)
This identification quantizes the y coordinate in X into lp size bits. Thus length is quan-
tized as opposed to area, which is what happens for the geometry dual to the half-BPS
local operators [13][16], where it can be interpreted as the quantization of phase space in
the boundary gauge theory.
A half-BPS surface operator OΣ develops a pole for the scalar field Φ (2.7). The
pole is characterized by its residue, which is given by 2M real parameters (βl, γl). These
parameters are identified with the position of the M “charges” in the ~x-plane in X via:
(βl, γl) =
2πl2s
. (3.20)
All these parameters take values on the real line.
The remaining parameters characterizing a half-BPS surface operator OΣ are the
periodic variables (αl, ηl), which determine the holonomy produced by OΣ and the corre-
sponding two-dimensional θ-angles. These parameters get identified with the holonomy of
the two-forms of Type IIB supergravity over the M non-trivial disks Dl that the geometry
generates in the presence of M “charges” in X :
αl = −
(3.21)
The identification respects the periodicity of (αl, ηl), which in supergravity arises from the
invariance of BNS and BR under large gauge transformations
We have given a complete dictionary between all the parameters that a half-BPS
surface operator in N = 4 SYM depends on and all the parameters in the corresponding
bubbling geometry. We note that a surface operator OΣ depends on a set of parameters
up to the action of the permutation group SM on the parameters, which is part of the
U(N) gauge symmetry. The corresponding statement in supergravity is that the solution
dual to a surface operator is invariant under the action of SM , which acts by permuting
the “charges” in X .
20 Since the gauge invariant variables are e
The supergravity solution is regular as long as the “charges” do not collide. A singu-
larity arises whenever two point “charges” in X coincide (see Figure 1):
(~xl, yl) → (~xm, ym) for l 6= m. (3.22)
Whenever this occurs, there is a reduction in the number of independent disks since (see
Figure 3):
Dl → Dm for l 6= m, (3.23)
and therefore
for l 6= m.
(3.24)
In this limit of parameter space the non-trivial S2 connecting the points Pl and Pm in X
shrinks to zero size as [S2l,m] = [Dl]− [Dm] → 0, and the geometry becomes singular.
By using the dictionary developed in this paper, such a singular geometry corresponds
to a limit when two of each of the set of parameters (αl, βl, γl, ηl) defining a half-BPS surface
operator OΣ become equal:
αl → αm, βl → βm, γl → γm, ηl → ηm for l 6= m. (3.25)
In this limit the unbroken gauge group preserved by the surface operator OΣ is enhanced
to L′ from the original Levi group
l=1 U(Nl), where L ⊂ L′. As explained in section 2
the path integral from which OΣ is defined becomes singular.
In summary, we have found the description of all half-BPS surface operators OΣ in
N = 4 SYM in terms of solutions of Type IIB supergravity. The asymptotically AdS5×S5
solutions are regular and when they develop a singularity then the corresponding operator
also becomes singular.
S-Duality of Surface Operators from Type IIB Supergravity
The group of dualities of N = 4 SYM acts non-trivially [4] on surface operators OΣ
(see discussion in section 2). For G = U(N) the duality group is SL(2, Z) and its proposed
action on the parameters on which OΣ depends on is [4]:
(βl, γl) → |cτ + d| (βl, γl)
(αl, ηl) → (αl, ηl)M−1,
(3.26)
where M is an SL(2, Z) matrix
. (3.27)
We now reproduce21 this transformation law by studying the action of the SL(2, Z)
subgroup of the SL(2, R) classical symmetry of Type IIB supergravity, which is in fact the
appropriate symmetry group of Type IIB string theory. For that we need to analyze the
action of S-duality on our bubbling geometries.
SL(2, Z) acts on the complex scalar τ = C0 + ie
−φ of Type IIB supergravity in the
familiar fashion
τ → aτ + b
cτ + d
, (3.28)
where as usual τ gets identified with the complexified coupling constant of N = 4 SYM
(2.8). SL(2, Z) also rotates the two-form gauge fields22 of Type IIB supergravity
, (3.29)
while leaving the metric in the Einstein frame and the five-form flux invariant.
Given that the metric in (3.1) is in the Einstein frame, SL(2, Z) acts trivially on the
coordinates (~x, y). Nevertheless, since
ls = g
s lp with gs = e
φ (3.30)
the string scale transforms under SL(2, Z) as follows:
l2s →
|cτ + d| . (3.31)
Therefore, under S-duality:
2πl2s
→ |cτ + d| ~xl
2πl2s
. (3.32)
Given our dictionary in (3.20), we find that the surface operator parameters (βl, γl) trans-
form as in (3.26), agreeing with the proposal in [4].
21 If we apply the same idea to the LLM geometries dual to half-BPS local operators in [13],
we conclude that the half-BPS local operators are invariant under S-duality.
22 See e.g. [26][27].
The identification of the rest of the parameters is (3.21):
αl = −
(3.33)
Using the action of SL(2, Z) on the two-forms (3.29) and the identification (3.33), it
follows from a straightforward manipulation that the surface operator paramaters (αl, ηl)
transform as in (3.26), agreeing with the proposal in [4].
Acknowledgements
We would like to thank Xiao Liu for very useful discussions. Research at Perimeter In-
stitute for Theoretical Physics is supported in part by the Government of Canada through
NSERC and by the Province of Ontario through MRI. We also acknowledge further sup-
port from an NSERC Discovery Grant. SM acknowledges support from JSPS Research
Fellowships for Young Scientists.
Appendix A. Supersymmetry of Surface Operator in N=4 SYM
In this Appendix we study the Poincare and conformal supersymmetries preserved
by a surface operator in N=4 SYM supported on R1,1. These symmetries are generated
by ten dimensional Majorana-Weyl spinors ǫ1 and ǫ2 of opposite chirality. We determine
the supersymmetries left unbroken by a surface operator by studying the supersymmetry
variation of the gaugino in the presence of the surface operator singularity in (2.2)(2.7).
The metric is given by:
ds2 = −(dx0)2 + (dx1)2 + (dx2)2 + (dx3)2 = −(dx0)2 + (dx1)2 + 2dzdz̄. (A.1)
where z = 1√
(x2 + ix3) = |z|eiθ, while the singularity in the fields is
Φω = Φω̄ =
(Φ8 + iΦ9) =
β + iγ√
A = αdθ, F = dA = 2παδD,
(A.2)
where [α, β] = [α, γ] = [β, γ] = 0 and δD = d(dθ) is a two form delta function. The relevant
Γ-matrices are
(Γ2 − iΓ3)
Γz̄ =
(Γ2 + iΓ3)
{Γz,Γz̄} = 2.
(A.3)
A Poincare supersymmetry transformation is given by
δλ = (
µν +∇µΦiΓµi +
[Φi,Φj]Γ
ij)ǫ1 (A.4)
where µ runs from 0 to 3 and i runs from 4 to 9, while a superconformal supersymmetry
transformation is given by
δλ = [(
µν +∇µΦiΓµi +
[Φi,Φj]Γ
ij)xσΓσ − 2ΦiΓi]ǫ2 (A.5)
From (A.4), it follows that the unbroken Poincare supersymmetries are given by:
Γω̄zǫ1 = 0 ⇔ Γ2389ǫ1 = −ǫ1. (A.6)
The unbroken superconformal supersymmetries are given by:
−β + iγ
Γzω̄ − β − iγ
0 + x1Γ
1 + zΓz̄ + z̄Γz)− 2ΦωΓω − 2Φω̄Γω̄
ǫ2 = 0.
(A.7)
From the terms proportional to x0 and x1, we find that the unbroken superconformal
supersymmetries are given by:
Γzω̄ǫ2 = 0 ⇔ Γ2389ǫ2 = −ǫ2. (A.8)
The rest of the conditions
[Γz̄Γω̄Γz]ǫ2 = 0, (A.9)
are automatically satisfied once (A.8) is imposed.
We conclude that the singularity (2.2)(2.7) is half-BPS and that the preserved su-
persymmetry is generated by ǫ1 and ǫ2 subject o the constraints Γ2389ǫ1 = −ǫ1 and
Γ2389ǫ2 = −ǫ2.
By acting with a broken special conformal transformation on Σ = R2 ⊂ R4 to get a
surface operator supported on Σ =S2, one can show following [28] that such an operator
also preserves half of the thirty-two supersymmetries, but are now generated by a linear
combination of the Poincare and special conformal supersymmetries.
Appendix B. Five form flux
In this Appendix, we calculate (3.5) explicitly to evaluate the flux over a non-trivial
The five-form flux is [13][16]
{d[y2 2z + 1
2z − 1(dχ+ V )]− y
3 ∗3 d(
z + 1
)} ∧ dV olAdS3
{d[y2 2z − 1
2z + 1
(dχ+ V )]− y3 ∗3 d(
z − 1
)} ∧ dΩ3
(B.1)
The five-cycle S5l in the bubbling geometry is spanned by coordinates Ω3, χ and y.
Then the integration is:
4π4l4p
F5 = −
16π4l4p
2z − 1
2z + 1
dχ] ∧ dΩ3 =
4πl4p
= N. (B.2)
References
[1] J. M. Maldacena, “The large N limit of superconformal field theories and supergrav-
ity,” Adv. Theor. Math. Phys. 2, 231 (1998) [Int. J. Theor. Phys. 38, 1113 (1999)]
[arXiv:hep-th/9711200].
[2] S. S. Gubser, I. R. Klebanov and A. M. Polyakov, “Gauge theory correlators from
non-critical string theory,” Phys. Lett. B 428, 105 (1998) [arXiv:hep-th/9802109].
[3] E. Witten, “Anti-de Sitter space and holography,” Adv. Theor. Math. Phys. 2, 253
(1998) [arXiv:hep-th/9802150].
[4] S. Gukov and E. Witten, “Gauge theory, ramification, and the geometric langlands
program,” arXiv:hep-th/0612073.
[5] J. Gomis and F. Passerini, “Holographic Wilson loops,” JHEP 0608, 074 (2006)
[arXiv:hep-th/0604007].
[6] S. J. Rey and J. T. Yee, “Macroscopic strings as heavy quarks in large N gauge
theory and anti-de Sitter supergravity,” Eur. Phys. J. C 22, 379 (2001) [arXiv:hep-
th/9803001].
[7] J. M. Maldacena, “Wilson loops in large N field theories,” Phys. Rev. Lett. 80, 4859
(1998) [arXiv:hep-th/9803002].
[8] N. Drukker and B. Fiol, JHEP 0502, 010 (2005) [arXiv:hep-th/0501109].
[9] S. Yamaguchi, “Bubbling geometries for half BPSWilson lines,”arXiv:hep-th/0601089.
[10] S. Yamaguchi, “Wilson loops of anti-symmetric representation and D5-branes,” JHEP
0605, 037 (2006) [arXiv:hep-th/0603208].
[11] O. Lunin, “On gravitational description of Wilson lines,” JHEP 0606, 026 (2006)
[arXiv:hep-th/0604133].
[12] J. Gomis and F. Passerini, “Wilson loops as D3-branes,” JHEP 0701, 097 (2007)
[arXiv:hep-th/0612022].
[13] H. Lin, O. Lunin and J. M. Maldacena, “Bubbling AdS space and 1/2 BPS geome-
tries,” JHEP 0410, 025 (2004) [arXiv:hep-th/0409174].
[14] J. Gomis and C. Romelsberger, “Bubbling defect CFT’s,” JHEP 0608, 050 (2006)
[arXiv:hep-th/0604155].
[15] J. Gomis and T. Okuda, “Wilson loops, geometric transitions and bubbling Calabi-
Yau’s,” JHEP 0702, 083 (2007) [arXiv:hep-th/0612190].
[16] H. Lin and J. M. Maldacena, “Fivebranes from gauge theory,” Phys. Rev. D 74,
084014 (2006) [arXiv:hep-th/0509235].
[17] A. Kapustin and E. Witten, “Electric-magnetic duality and the geometric Langlands
program,” arXiv:hep-th/0604151.
[18] E. Frenkel, “Lectures on the Langlands program and conformal field theory,”
arXiv:hep-th/0512172.
http://arxiv.org/abs/hep-th/9711200
http://arxiv.org/abs/hep-th/9802109
http://arxiv.org/abs/hep-th/9802150
http://arxiv.org/abs/hep-th/0612073
http://arxiv.org/abs/hep-th/0604007
http://arxiv.org/abs/hep-th/9803001
http://arxiv.org/abs/hep-th/9803001
http://arxiv.org/abs/hep-th/9803002
http://arxiv.org/abs/hep-th/0501109
http://arxiv.org/abs/hep-th/0601089
http://arxiv.org/abs/hep-th/0603208
http://arxiv.org/abs/hep-th/0604133
http://arxiv.org/abs/hep-th/0612022
http://arxiv.org/abs/hep-th/0409174
http://arxiv.org/abs/hep-th/0604155
http://arxiv.org/abs/hep-th/0612190
http://arxiv.org/abs/hep-th/0509235
http://arxiv.org/abs/hep-th/0604151
http://arxiv.org/abs/hep-th/0512172
[19] R. M. Rohm, “Some Current Problems In Particle Physics Beyond The Standard
Model,”
[20] J. Preskill and L. M. Krauss, “Local Discrete Symmetry And Quantum Mechanical
Hair,” Nucl. Phys. B 341, 50 (1990).
[21] M. G. Alford, K. M. Lee, J. March-Russell and J. Preskill, Nucl. Phys. B 384, 251
(1992) [arXiv:hep-th/9112038].
[22] A. Kapustin and S. Sethi, “The Higgs branch of impurity theories,” Adv. Theor. Math.
Phys. 2, 571 (1998) [arXiv:hep-th/9804027].
[23] N. R. Constable, J. Erdmenger, Z. Guralnik and I. Kirsch, “Intersecting D3-branes
and holography,” Phys. Rev. D 68, 106007 (2003) [arXiv:hep-th/0211222].
[24] A. Kapustin, “Wilson-’t Hooft operators in four-dimensional gauge theories and Phys.
Rev. D 74, 025005 (2006) [arXiv:hep-th/0501015].
[25] E. Witten, “Anti-de Sitter space, thermal phase transition, and confinement in gauge
Adv. Theor. Math. Phys. 2, 505 (1998) [arXiv:hep-th/9803131].
[26] J. Polchinski, “String Theory”, Cambridge University Press, Chapter 12 in Volume 2.
[27] K. Becker, M. Becker and J. Schwarz, ”String Theory and M-theory”, Cambridge
University Press, Chapter 8.
[28] M. Bianchi, M. B. Green and S. Kovacs, “Instanton corrections to circular Wilson
loops in N = 4 supersymmetric JHEP 0204, 040 (2002) [arXiv:hep-th/0202003].
http://arxiv.org/abs/hep-th/9112038
http://arxiv.org/abs/hep-th/9804027
http://arxiv.org/abs/hep-th/0211222
http://arxiv.org/abs/hep-th/0501015
http://arxiv.org/abs/hep-th/9803131
http://arxiv.org/abs/hep-th/0202003
