#### Decay of ultralight axion condensates

Received: June
Decay of ultralight axion condensates
Joshua Eby 0 1 2 3
Michael Ma 0 1 3
Peter Suranyi 0 1 3
L.C.R. Wijewardhana 0 1 3
0 P. O. Box 500, Batavia, IL 60510 , U.S.A
1 2600 Clifton Ave , Cincinnati, OH 45221 , U.S.A
2 Fermi National Accelerator Laboratory
3 Department of Physics, University of Cincinnati
Axion particles can form macroscopic condensates, whose size can be galactic in scale for models with very small axion masses m 10 22 eV, and which are sometimes referred to under the name of Fuzzy Dark Matter. Many analyses of these condensates are done in the non-interacting limit, due to the weakness of the self-interaction coupling of axions. We investigate here how certain results change upon inclusion of these interactions, nding a decreased maximum mass and a modi ed mass-radius relationship. Further, these condensates are, in general, unstable to decay through number-changing interactions. We analyze the stability of galaxy-sized condensates of axion-like particles, and sketch the parameter space of stable con gurations as a function of a binding energy parameter. We nd a strong lower bound on the size of Fuzzy Dark Matter condensates which are stable to decay, with lifetimes longer than the age of the universe.
Classical Theories of Gravity; Cosmology of Theories beyond the SM
1 Introduction
2
3
4
5
Decay through self interaction
The calculation of the wave function of the condensate
Stable spectrum of ultralight axion condensates
Conclusions
A The AN ! AN 3 + ap decay
B The continuous spectrum in the Ru
ni-Bonazzola equations
Axions are hypothetical particles introduced to solve the Strong CP problem in QCD [1{8].
Axions are self-adjoint bosons, with no conserved discrete quantum numbers to guarantee
particle number conservation. The axion potential can be written in terms of an angular
variable with a 2
shift symmetry. Axion-like scalar particles also appear in a variety
of models beyond QCD, especially in low-energy limits of string theories [9{11]. Those
axions have properties similar to QCD axions, but their mass scales and decay constants
are vastly di erent.
Axion-like particles are among the prime candidates for the composition of dark
matter [12{17]. Axions, being scalar bosons, can condense. Axion condensates have been
discussed by numerous groups, with condensate sizes ranging from galaxy or galaxy
cluster scale [18, 19], down to stellar size and smaller (termed \axion stars") [20{25], all the
way to radii of a few meters [26, 27]. Also of interest is the cosmological evolution of the
axion
eld, which has been worked on extensively in [28{32], but this will not be disussed
further here.
Uncondensed QCD axions are not stable, as they can decay to photons through a
process a ! 2 , but the decay rate is slow enough such that most axions created after
the Big Bang would survive many Hubble times [33]. Axion condensates, however, may
also decay slowly due to the self-interaction of axions [34, 35]. The self-interaction term of
their Lagrangian (for both the instanton [36, 37] and chiral [37, 38] cases) have only terms
containing an even number of axion elds. Thus, disregarding the rare decay into photons,
the axion number is conserved only modulo 2.
{ 1 {
In a recent paper [34] we have investigated the decay of weakly bound axion stars due to
the self-interaction of axions. The decay proceeds mostly through a sequence of processes,
where Ak is an axion star, which is a condensate containing k axions, and ap denotes an
axion in a scattering state with the magnitude of the momentum p. The process (1.1) is
the simplest of many possible decay modes responsible for the decay of axion stars.1 This
process is allowed by energy-momentum conservation, provided the binding energy of a
bound axion is small enough that a relativistic particle can be produced:
E < 2 m = 3,
where m is the mass of a free axion.
In [34] we used an axion
eld operator, which was the generalization of the eld
proposed by Ru ni and Bonazzola [39]. To facilitate the decay process, terms creating and
annihilating axions in the continuum of scattering states were included in the quantum
eld
of axions, in addition to terms creating and annihilating bound axions [39]. The Ru
niBonazzola method is based on taking the expectation value of the quantum Einstein and
Klein-Gordon equations in the condensate to derive equations of motion for the metric
components and the scalar eld. We solved the equations of motions numerically in the
weak gravity and weak binding ( E
m) limits, to nd the wave function of axions in
the condensate [24].
The bound axions are not in momentum eigenstates. They have de nite energies, but
their wave functions extend over the size of the axion star. Accordingly, the bound axions
have an extended momentum distribution as well. In the recent publications [35, 40],
the authors questioned the validity of the decay mechanism proposed in [34], arguing that
momentum is not conserved in (1.1), and that the decay rate through this process is exactly
zero by the Optical Theorem. These authors have further suggested that one can show
the rate to be zero by the classical equation of motion for the condensate. We will address
these issues and explain our response in appendix A.
In the present paper we will apply our method of discussing the decay of dilute axion
stars [34] to condensates of cosmological size; such models have been referred to previously
as Fuzzy Dark Matter (FDM) [41]. Condensates of galactic sizes have been considered by
a number of authors, and typically correspond to a scalar particle mass of m
10 21 eV [19, 41{50].2 When such condensates are formed from real scalars, a version of
the decay analysis of [34] applies, and we will investigate whether interesting bounds can
be placed on these models by taking decays into account. As we will explain in the next
section, we will utilize the axion potential with a cosine dependence on the
eld; other
proposals, for example a cosh potential [52, 53], have been investigated in the context of
ultralight scalars as well.
It is also an aim of this paper to emphasize the inclusion of axion self-interactions in
investigations of axion condensates. Although the self coupling
m2=f 2
10 95 n 1
for typical FDM models (f is the axion decay constant), the astronomical number of
unlikely to have any cosmological signi cance [34].
1The decay rates via processes AN ! AN 5 + ap or AN ! AN 4 + ap1 + ap2 are signi cantly lower and
2For a brief but recent review of ultralight scalar eld dark matter models, see [51] and references therein.
{ 2 {
HJEP01(28)6
axions in a condensate participating in these interactions could lead to large corrections
to certain important physical quanties. We investigate the macroscopic properties of these
condensates using the fully self-interacting analysis and emphasize the di erences from the
non-interacting limit.
In the next section, we give a more detailed explanation of how axion star decay
through the process (1.1) can be calculated. In section 3, we will outline the calculation
of the wavefunction, following largely [24]. In section 4, we apply the formulas for the
macroscopic properties and decay rates to condensates formed from ultralight axion-like
particles. We conclude in section 5.
We will use natural units throughout, where ~ = c = 1.
2
Decay through self interaction
There is a variety of methods for the quantitative investigation of axion condensates, as
classical [20, 21, 33, 54, 55], quantum mechanical [25, 56{61], and
eld theoretic [24, 26,
27]. Our eld theoretic discussion of the decay of an axion condensates into relativistic
axions [34] was based on an extension of the Ru ni-Bonazzola operator [39], by the addition
of scattering state contributions.3 Thus, we proposed to extend the expansion of the boson
eld using the form [34]
where a(p~) is the annihilation operator for a particle which is the eigenstate of the
momentum operator with eigenvalue p~ (i.e. a plane wave). This annihilation operator and its
adjoint, the creation operator, satisfy the commutation relation
[alm(p); aly0m0 (p0)] = 2 !p (2 )3 (p
p0) ll0 mm0 :
3Appendix B we will discuss why a continuous spectrum of scattering state solutions can be added to
the boson
eld operator. Furthermore, we will also discuss why using free spherical wave scattering states
is quite su cient in our calculations.
{ 3 {
(r; t) = R(r) e i E0 t a0 + R(r) ei E0 t ay0 + f (r; t) + fy (r; t);
where E0 is the energy eigenvalue of a single bound axion, and where R(r) and a0 are
the wave function and annihilation operator of the axions in the condensate (respectively).
f (r; t) is the annihilation part of a complete system of free axion operators expanded in
scattering states,
f (r; t) =
1
where !p and alm(p) are the energy eigenvalue and the annihilation operator of the
scattering state axion, with quantum numbers l and m, respectively. The functions jl(x) and
Ylm(x^) are spherical Bessel functions and spherical harmonics, respectively.
The annihilation operator in the spherical wave basis is de ned by
alm(p) = il p Z
d p Ylm (p^) a(p~);
(2.1)
(2.2)
(2.3)
(2.4)
Note that (2.2) is exactly equal to the negative frequency part of a complete system of free
axion states given by
1
2 !p
Z d3p ei(p~ ~r !p t)a(p~);
which was used in [34] to investigate the decay of QCD axion stars.
We will see later that the bosons produced by the decay of a weakly bound boson
condensate are relativistic. Therefore, we have chosen to use free particle states to
approximate the scattering states. For the purposes of this calculation, this choice is admissible,
because the energy level of produced axions is su ciently high compared to the e ective
depth of the potential created by gravitation and self-interactions. This is explained in
greater detail in appendix B. In a future work we will take into account corrections to this
approximation.
Let us consider now (1.1) in the Born approximation. The axion self-interaction
po(2.5)
(2.6)
HJEP01(28)6
tential can be approximated by the so-called instanton potential [36, 37]
V ( ) = m2 f 2 1
cos
;
f
Z
where m and f are the axion mass and decay constant (respectively). In the Ru
ni
Bonazzola paradigm, one
nds the expectation value of eq. (2.6) transforms the cosine
into a Bessel function J0 [24].4 One also
process (1.1) is
nds that the transition matrix element for the
considering transitions of the form (1.1), where hN j is the initial N particle condensate
(the left hand side of (1.1)), and jN
3; (p)i is the direct product of the nal state N
3
particle condensate and a scattering state (p) of momentum magnitude p (the right hand
side of (1.1)).
We restrict this work to non-rotating axion condensates; the reason for this is twofold.
First, for the potential in eq. (2.6) and the parameters we use here, only the inner cores of
galaxies composed of axion particles can be described as a condensate; outside of this inner
region, the dark matter halo is described by a virialized gas of particles [19] and cannot
participate in the decay processes we describe here.5 Because the condensed core is small
compared to the full radius of the halo, it carries at most a tiny fraction of the angular
momentum of the galaxy, so as a rst approximation we believe restricting to ` = 0 angular
4More generally, the annihilation process for k bound axions generates an e ective potential proportional
to the Bessel function Jk, as explained in the appendix of [34].
can be much larger, and can even constitute the entire dark matter halo.
5For ultralight bosons with repulsive self-interaction, like those presented in e.g. [45, 48], the condensates
{ 4 {
momentum states is appropriate. The second reason is that a full treatment of rotating
axion condensates has not yet been done, though slowly rotating condensates were analyzed
in a particular limit by [14]. This is a topic we hope to return to in the near future.
Because we work in the limit of zero angular momentum, annihilation processes of the
form a a ! G, where G is a spin-2 graviton [62, 63], have a rate of zero. Such an interaction
would require the participating axions to have at least ~ of angular momentum each; and
even if we accounted for the nonzero rotation speed of the galaxy, by our estimation the
vast majority of axions in a galactic condensate would have far less angular momentum
than what would be necessary for this process to occur.
For static condensates, note that the integration over
r in eq. (2.7) vanishes for all
but s-wave axions.6 In that case, the scattering state axion is described by the zero angular
momentum contribution only, j (p)i = ay00(p)j0i. Then the wave function of the emitted
axion is
(r; t) = h0j f (r; t)j (p)i = p
4
e i !p t sin(p r)
r
:
The integration over t also xes the energy of the outgoing axion to !p = 3 E0. The matrix
element takes the form [34]
where the dimensionless integral is
tum representation wave function as
The symmetry of the integrand for the substitution r !
integration region to all real values of r, and to switch from sin(p r) to ei p r in the integrand.
In the second equality, we expanded the Bessel function J3 to leading order, an appropriate
r has been used to extend the
limit for dilute axion stars.
Now observe that for dilute axion stars the radius of the star R is very large. In
other words, X(r) has a large coordinate uncertainty, r
R
(m
) 1, where
=
1 [24, 34]. As a result, the range of p, as represented by the momentum
uncertainty p
m
m, is very small. Then due to the delta function in eq. (2.9),
enforcing energy conservation, the emitted axion has a momentum peaked at a very large
8 m; as a result, the matrix element (2.9) is very small for weak binding.
However, as the binding energy
E increases, M3 will take larger values and the decay
Now to bring out issues related to momentum conservation, we can de ne the
momen(q) =
1
d3r X(r) ei q~ ~r:
(2.11)
6Should we consider rotating axion stars, higher angular momentum scattering states would also
contribute.
{ 5 {
Then we can rewrite eq. (2.9) as
depends crucially on the large q tail of momentum distribution
(q). However, for
calculational purposes, it is still advantageous to use (2.9) rather than (2.12). One can rely on the
numerical solution of the equations of motion, as explained below, using a simple approach
to estimate approximate behavior of (q) at large q [34].
Note that the process (1.1) is not the only possible channel through which decay can
proceed; however, it is by far the dominant process. First, we have shown previously [34]
that processes of the form AN ! AN (2 j+1) + ap; are suppressed by higher powers of
for each higher j > 1. In the weak-binding limit, where
1, these corrections are
completely negligible. On the other hand, this argument breaks down for dense axion
stars [27, 64], where
= O(1); we will return to this case in a future publication.
Second, there are processes of the form AN ! AN k +
ap; where
> 1 axions are
emitted at once. Unlike the process (1.1), the emission of
> 1 axions from a condensate
can proceed on-shell, implying that the corresponding decay rate has a weak dependence
on
. Nonetheless, as shown in [34], these processes are suppressed by the very small factor
m2=f 2 for each additional axion in the nal state. Since in FDM m2=f 2
10 95 n 1, we
can safely neglect these corrections as well. We conclude that (1.1) is by far the dominant
contribution to the decay of axion condensates.
3
The calculation of the wave function of the condensate
We review here the calculation of the condensate wavefunction X(r) [24]. The matrix
elements of the rr and tt components of the Einstein equation, along with the
KleinGordon equation, form a closed set of equations for the metric and the axion eld, X(r):
A0
A
B0
B
X00 =
=
=
1
A
r
r
A
1
2
r
+ 2 r A
+ 2 r A
+
B0
2 B
E02 X2
E02 X2
B
B
X0
A
+
+
X02
A
X02
A
E02 X
B
+ m2X2
m2X2 +
X4 ;
X4 ;
m2X +
X3 ;
m2
16
m2
16
m2
8
where the metric is
ds2 =
B(r) dt2 + A(r) dr2 + r2 d 2
;
with
= f 2=MP 2 and MP = 1 =
G = 1:22
1019 GeV (the Planck mass). As above,
we have taken only the leading contribution to the Bessel function which represents the
self-interaction potential; doing so preserves the leading attractive X(r)4 interaction term.
Assuming that
1, a condition satis ed in applications where gravity is weak
(Newtonian limit), we can write A = 1 +
a and B = 1 +
b, where a; b = O(1).
Furthermore using the large radius approximation and the de nition of the scale parameter
=
Z
p
~q3) Y
3
k=1
A0
2 A
p
{ 6 {
(3.1)
(3.2)
(3.3)
(3.4)
HJEP01(28)6
E02=m, we can introduce dimensionless radial coordinate as x = m r . In that
case the axion
eld also scales with its engineering dimension, such that X(r) =
Y (x),
leading to the following system of equations for a, b, and Y in leading order of
Near the singular point x = i , this has a solution of the form
Y (x) =
8
x2 + 2
2
3
1
18 3 (x2 + 2) + O([x2 + 2]2):
7This de nition of
appears to di er by a factor of 8 compared with [34], because in that work we
wrote
in terms of the reduced Planck mass. In fact, the two de nitions of
are equivalent.
8The maximum masses for attractive interactions were discussed by Stoof [68] in the context of condensed
matter BECs, and in the context of boson stars by Chavanis and Del ni [56, 57].
9In fact, this expression contains the next to leading order term proportional to Y 0(x).
where7
= 8
= 2. Since b(x) is proportional to the Newtonian gravitational potential,
G is the e ective coupling constant of the eld Y (x) to gravity. Further details, and a
more comprehensive justi cation of this double expansion of the equations, can be found
in [24].
Solutions of the equations of motion (3.5) correspond to ground-state con gurations
of axions, which can be stable or metastable. In [24], we solved these equations and found
a spectrum of solutions which were parameterized by
(or, equivalently, by
). When
applied to QCD axion parameters m = 10 5 eV and f = 6 1011 GeV, we found a maximum
gravitationally stable mass of Mc
1019 kg.8 By rescaling these solutions to values of m
and f appropriate for FDM, we can analyze the properties of galactic-size condensates in
a way that includes the self-interaction term in the potential. The physical interpretation
of these condensates is that they form the cores of FDM halos; they are surrounded by a
virialized distribution of axions which extend to the outer edge of the dark matter halo.
To analyze the decay of these condensates through processes like (1.1), we investigate
the singularity structure of solutions of (3.5). Now, (3.5) is a system of equations with
two singular points, x = 0 and x = 1. Using boundary conditions we require that the
solutions are regular at x = 0 and decrease exponentially at x ! 1. The solutions are
even functions of x, so they also approach zero at x !
1. Then they are real analytic
functions at
1 < x < 1 and can be continued into the complex plane of x. As they fast
vanish at in nity, the contour of integration can be moved up along the imaginary axis in
the rescaled version of the integrals in eqs. (2.9) and (2.10), until we reach a singularity in
the complex plane. The contribution of that singularity dominates the decay rate integrals
It is easy to show that the Klein-Gordon equation (3.5), in which the leading order
at large momentum.
singular terms are retained, is9
{ 7 {
(3.5)
(3.6)
(3.7)
The parameter
is an integration constant, having a one-to-one relationship with the
rescaled central density of the axion
eld Y (0)2, and in turn, with mass and the radius of
the condensate. In fact, high order Taylor series expansion of equations around x = 0 show
that the singularity closest to the origin is indeed of the form (3.7), connecting the value
of Y (0) with
[34]. In principle, gravitational interactions have an e ect on the solutions
Y (x). In practice, however, the term in the equations of motion (3.5) which couple Y (x)
to gravity give a subleading contribution to the singularity. We can therefore solve the
equations in the limit
1, i.e. where gravity decouples. In that limit, the nontrivial
solution has Y (0) = 12:268, which implies a xed value
Finally, we can rewrite the Fourier transform of eq. (2.11) as
the singular term of (3.7) term dominates the
integral. To calculate the decay rate, we follow the procedure of [34]: we take the leading
order solution for Y (x) near the singularity x = i , given by eq. (3.7), and evaluate I3(p)
in the matrix element of eq. (2.9). The result is
I3(p0) '
The decay rate for the process (1.1) is then
dN ' m
d
dM '
=
3 yM
2
Further details regarding the evaluation of eq. (3.11) can be found in [34]; the result for
the process (1.1) is
where yM ' 75:4 is determined by the relationship between M and
in the large
region [24]. The lifetime is a monotonically decreasing function of
in the relevant range;
in the case of QCD axions, we found in [34] that above a value
condensates become very unstable to decay to relativistic axions, their lifetimes becoming
shorter than the age of the universe. We will examine the consequences of this fact in the
' :05
:06, axion
context of ultralight axions in the next section.
{ 8 {
10-6
in the shaded region are unstable to decay to relativistic axions with a very short lifetime. Here
we have used the model parameters m = 10 22 eV, and f in the range between 1014 and 1018 GeV;
increasing the particle mass m merely shifts these curves down proportionally to 1=m.
4
Stable spectrum of ultralight axion condensates
Very light axion elds can have de Broglie wavelengths as large as entire dark matter halos,
possibly implying a connection between these two scales. Ultralight bosons appear often in
theories of physics beyond the Standard Model, including those requiring compacti cation
of extra dimensions. Such models, termed \Fuzzy Dark Matter" (FDM) [41], have been
written about extensively [19, 41{50].
While there are signi cant constraints on these
models,10 they remain a viable alternative to WIMP or QCD axion models of dark matter.
We will consider an ultralight axion in this context, using the potential of eq. (2.6).
The mass of the ultralight axion in question will be taken to be m
the right approximate scale for the size of dark matter halos [19],
10 22 eV, which gives
2
dB =
1
m v
= 1:92 kpc
10 22 eV
m
10 km/sec
v
;
(4.1)
where v is the velocity in the halo. This choice is also consistent with the known epoch of
matter-radiation equality. Further, a decay constant of f
5
1016 GeV naturally leads to
the correct relic density, and can thus account for the observed dark matter abundance [19];
however, to remain as general as possible, we allow f to deviate from this value by a few
10While this work was being nalized, a paper appeared suggesting a strong tension between the preferred
mass scale for FDM, m
10 21 eV, and data from Lyman- forest simulations [65]. We will not
comment here about whether such constraints could rule out FDM as a viable paradigm.
{ 9 {
HJEP01(28)6
10-6
m
the shaded region are unstable to decay to relativistic axions with a very short lifetime. Here we
have used the model parameters m = 10 22 eV, and f in the range between 1014 and 1018 GeV;
increasing the particle mass m merely shifts these curves down proportionally to 1=m.
orders of magnitude. At the upper limit of what we consider, f = 1018 GeV is still below the
Planck scale, implying that the parameter
approximation holds reasonably well over our entire range.
= f 2=MP2
:007
1; thus, the weak-gravity
In [24], we found numerically the solutions to the system (3.5) over a wide range of .
In the FDM picture, these solutions correspond to the cores of FDM halos discussed (most
recently) in [19]. We found that there exists a maximum mass at
:34, above which
axion condensates are gravitationally unstable. On the stable branch of masses
> :34,
the mass M and radius11 R99 of the condensate are t by the functions [24]
M ( )
p
gure 1 that at xed m, the value of f determines the position of the
maximum mass, and thus the turnaround of the function M ( ). A similar structure can be
observed in gure 2 for the radius, where the position of the maximum mass corresponds to
a slight dip in the otherwise straight line representing R99( ). Trading
for
in eq. (4.2),
we see that the lines
M ( )
1:75
R99( )
5:75
m
bound the full set of solutions from above.
located, represents the \size" of the condensate.
11We use the common convention that R99, the radius inside which :99 of the mass of the condensate is
(4.2)
(4.3)
For the speci c choice of FDM parameters m = 10 22 eV and f = 5
nd the maximum mass Mc
interacting limit of 8
1011M
6
1010M ; this is lower than the value found in the
non[19] by about an order of magnitude, due to the inclusion
of attractive self-interactions. Our estimate of the maximum mass also agrees well with
the recent analysis of [66], which also includes the leading attractive self-interaction. More
generally, the mass and radius of FDM halo cores over a wide range of the scale parameter
and at di erent values of f are illustrated in
gures 1 and 2.
We can also analyze the relationship between M and R99, which were investigated
for both attractive and repulsive self-interactions in [56, 57]. In a recent paper [19], the
authors present a bound on the product
where R1=2 is the radius inside which :5 of the mass of the condensate is contained; the
inequality is saturated for stationary, ground state con gurations, i.e. for the condensates
considered here. In our calculation, on the stable branch of solutions (where
> :34), we
nd the product
using eq. (4.2). To nd the relationship between R99 and R1=2, we calculated their ratio
numerically and found that R99=R1=2
over many orders of magnitude. This implies
3:65 holds within 1%, in a range of 1=
extending
M R1=2 = 2:76
(interacting axions):
This product is below the lower bound (4.4) presented in [19] due to our inclusion of
self-interactions.
It is worth noting also that in the limit f
MP , the mass-radius relationship for axion
condensates approaches that of a black hole. This is easy to see using eq. (4.2):
G M
R99
=
1
p
Near
= O(1) (the position of the maximum mass) and f
implying that R99
RS, the Schwarzschild radius.
MP , we nd G M=R99
1,
We must also ensure that the weak-binding approximation, on which our analysis [34]
and the classical one of [19] depends, is also valid.12 We observe in
gure 2 that cores of
radius R . 1 pc have
& :3, and become relatively strongly bound. Such cores would not
be well-described by our weak-binding analysis.
An estimate of the decay rate, obtained from the expression derived in [34], is given
in eq. (3.12); it is a one-to-one function of the binding energy parameter
in the region
of interest. Because the condensate mass M is determined by the value of
, it is easy
12A stability analysis for very strongly-bound condensates, with
= O(1), is a task we plan to undertake
in the near future.
(4.4)
(4.5)
(4.6)
(4.7)
to connect
to M as well. Following the analysis of [34], we nd that axion condensates
with m
10 22 eV which have
& 0:1 have lifetimes shorter than the age of the universe.
This region is represented by the shaded regions in
gures 1 and 2. For f . 1018 GeV, the
transition to decay instability occurs on the gravitationally unstable branch of solutions;
however at f & 1018 GeV, the bounds from decay are as strong or stronger than those
coming from gravitational stability. This can be an important constraint on bound structures
originating in theories of Planck scale axions.
In gure 2, it is particularly striking that almost regardless of the value of f ,
condensates with R . 2 pc lie in the unstable, shaded region. This implies a fundamental limiting
radius of Rmin
2 pc for FDM dark matter cores composed of axions. Observe also that,
gure 1, it is easy to read o the maximum mass of FDM condensates for each value
of the decay constant f . For any axion theory with f
MP , no stable condensate exists
with a mass larger than about Mmax
1012M .
5
Conclusions
In a previous publication [24] we established scaling relations for the mass and radius of
weakly bound condensates of interacting axions, as functions of the mass of the axion,
its decay constant, and the particle energy (or alternatively the central density). We also
found the maximum mass and size of the bound states as functions of those parameters. In
this paper we have applied those results to condensates of axions forming FDM, providing
corrections to similar calculations which neglect the self-interaction of axions [19, 41, 44,
45, 47, 49, 50].
In another publication [34] we developed methods to calculate the lifetime of axion
condensates due to their self-interaction, through the four-axion interaction term in which
three bound axions produce a single free relativistic axion. Here we have applied those
results to estimate the lifetime of condensates formed from FDM. We have found that,
provided the decay constant of FDM axions satis es f . 0:05 MP , all condensates having
binding energy smaller than that of those of maximal mass have lifetimes greater than the
age of the universe making them viable candidates for forming central regions of galactic
halos. We have also explained in details the decay mechanism described in [34] and further
clari ed the justi cation of its validity.
The methods we have described here, based on previous work in [24, 34], rely on a
double expansion to leading order in
and
. This is appropriate for so-called dilute
axion stars, which are weakly bound. However, it has been pointed out that an e ective
short-distance repulsive interaction in the axion potential also gives rise to dense axion
stars [27, 64, 73], which are at least energetically stable. We plan to extend our methods
to this regime to analyze the properties of these states in the near future.
Collapsing boson condensates have been investigated by a number of groups [64, 67{
73]. Recently, we found that supercritical QCD axion condensates, having masses larger
than the maximal allowed stable mass, collapse towards the global minimum of the e ective
axion potential [64, 73]. Similar arguments indicate that FDM axion condensates which
exceed the maximum mass Mc will also collapse in this way. Such supercritical condensates
can form during galactic collisions, in a manner similar to the mechanism outlined in [74];
such events would lead to collapse, causing the condensate to emit a large number of
relativistic particles. Consequences of such events will be studied in a future publication.
Acknowledgments
We thank M. Amin, P. Argyres, E. Braaten, J. Brod, P. Fox, R. Harnik, A. Kagan, A.
Mohapatra, K. Schutz, G. Semeno , M. Takimoto, H. Zhang, and J. Zupan for fruitful
discussions. The work of J.E. was supported by the U.S. Department of Energy, O
ce
of Science, O
ce of Workforce Development for Teachers and Scientists, O
ce of Science
Graduate Student Research (SCGSR) program. The SCGSR program is administered by
the Oak Ridge Institute for Science and Education for the DOE under contract number
DE-SC0014664. J.E. also thanks the Fermilab Theory Group and the Weizmann
Institute Department of Physics for their hospitality. L.C.R.W. thanks Mainz Institute for
Theoretical Physics for their hospitality, and the participants of the Quantum Vacuum
and Gravitation program, especially M. Bartelmann, A. Mazumdar, and T. Prokopec,
for discussions.
A
The AN ! AN 3 + ap decay
The results of our paper, \The Lifetime of Axion Stars" [34], have been called into question
in recent publications [35, 40]. Before discussing this issue in detail, we review the premises
of our work. We proposed a way to discuss the decay of axion stars through the repeated
elementary decay mode
where Ak denotes an axion star, which is a condensate of k axions and ap denotes a
(relativistic) free axion labeled by its momentum p. This calculation was performed using
an extended axion eld operator, eq. (2.1), which included both bound and scattering state
contributions.
First of all, we need to establish the fact that there is no conservation law that would
forbid (A.1). Axions are real bosons, and consequently the axion number is not conserved.
Axions, being coupled to the electromagnetic eld as
E B, can decay to photons through
the slow process a ! 2 , though this decay process does not signi cantly a ect the lifetime
of axion stars [33]. Disregarding axion decay to photons, the axion number is conserved
modulo 2, as the self-interaction terms of the axion potential contains only even powers of
the axion
eld, through a dependence of cos( = f ) in eq. (2.6). Momentum and energy
conservation would allow the decay process to proceed even if the condensates were in
momentum eigenstates: a decay process is always allowed if the sum of the masses of the
decay products is smaller than the mass of the decaying object, unless the conservation of
discrete quantum numbers prevents it.
At any rate, the condensates represented in (A.1) are not in momentum eigenstates.
They are quantum objects, which have extended wave functions localized on a large radius
R. Consequently, though they have mean momentum of zero, their momentum
distributions are smeared. In fact, the momentum distributions extend from zero to in nity, albeit
with fast decreasing amplitudes.
We have to emphasize that a Gross-Pitaevskii approach, being the non-relativistic limit
of the relativistic quantum
eld theory, cannot be used to discuss particle number violating
processes. In the non-relativistic limit the axion number is conserved and process (A.1) is
forbidden. Note however that a method to adapt the Gross-Pitaevskii approach to these
processes was formulated by the authors of [75]. We also point out Gell-Mann's totalitarian
principle [76], borrowed from T. H. White's \The Once and Future King", which when it
is applied to physics is: \Everything not forbidden is compulsory." This principle implies
that the process (A.1) should exist. In the present context, this principle also implies
that taking the nonrelativistic limit discards important contributions that make such a
transition possible.
Now the questions raised in [35, 40] have two sides: (a) microscopic, and (b)
macroscopic. The microscopic argument (a) pertains to the question of momentum conservation
in the elementary process of three bound axions at rest turn into a moving axion. The
macroscopic argument (b) concerns the whole condensate, namely how momentum is
conserved in the overall process (A.1), i.e. how the momentum carried away by the produced
scattering state axion is transferred to the outgoing axion star, AN 3
.
The argument of [35, 40] pertaining to the microscopic side (a) is that in the
overwhelmingly dominating Born approximation, the process around which the decay process (1.1) is
built is
3 ac ! af
(A.2)
axion is p
R 1
momentum to create a free axion, p = p
of p '
p
where ac represents a bound axion and af represents the nal state axion, is inadmissible
due to energy-momentum conservation. The total energy of the three bound axions is
Etot ' 3 m, which is certainly su cient to produce a free relativistic axion. However, if the
three axions were in momentum representation, having zero momentum in the rest system
of the axion star, then the axion in the nal state would not have the required momentum
To resolve this problem, observe that the three axions contributing to the decay process
are not at rest. The notion of \rest" is tied to particles in momentum representation,
and for the axions in the axion star, this is not so. They have momentum space wave
functions, which, though peaked near zero, extend to large momenta (albeit with very
small probability). As we discussed above, the uncertainty of the momentum of each
. The probability that three of the bound axions have su cient total
m2, decreases rapidly with the size of a
condensate, though it is never exactly zero. Consequently, local momentum conservation
always allows the decay of condensate. The question of whether this decay process a ects
cosmology, due to the survival or non-survival of the condensate, is a question of numerical
calculations and depends on the parameters of the axion theory, as well as the size and
mass of the condensate.
Another argument presented in [35, 40] to the local momentum conservation is based
on the optical theorem. The argument invokes diagram
3 ac ! af ! 3 ac
(A.3)
HJEP01(28)6
claiming that the propagator of the axion af in the intermediate state does not have an
imaginary part, because presumably its denominator, Et2ot
p
2
m2 ' 8 m2 6= 0. The
diagram having no imaginary part, the decay rate must vanish. However, this argument
is based again on the premise that the momentum of the condensed axions in the initial
and
nal states of the process (A.3) is zero. This is not a valid assumption, since those
particles are not in momentum eigenstates and with a tiny probability they can produce
su cient momentum to allow the axion in the intermediate state to go on mass-shell. A
comprehensive discussion of how the imaginary part appears in a process like (A.3) for
particles in a condensate can be found in [75].
The macroscopic argument (b) claims that even if the above discussed elementary
process is possible, there is no mechanism by which the three axions participating in the
decay process transmit the momentum to the axion star as a whole; so (the argument goes)
contrary to the requirement of momentum conservation in (A.1) the recoil of the axion star
AN 3 in the nal state of (A.1) is not possible.
Consider, however, that the condensate itself is not in momentum representation either,
unlike condensates in a container, for which the boundary conditions are set by the wall
of the container. Still, one could evoke the valid argument that just like the expectation
value of the coordinate, the expectation value of the momentum of the condensate must
be zero. Consequently, the only constraint we can impose on the decay process is the
conservation of the expectation value of the momentum. Now, the condensate of the nal
state of (A.1) is not in momentum representation either, though its average momentum
is zero as well. Consider now the created scattering state axion. It is produced as a zero
angular momentum spherical wave, going with the same probability into every direction.
Though the magnitude of the momentum is sharply peaked at a particular value, the
spherical wave also has a vanishing average momentum. Thus, the average momentum is
conserved in (A.1).
One should not confuse the emission of the axion with its detection. Suppose we detect
an emitted axion, the decay product of (A.1). By performing a measurement we alter
the system. Just like in the case of the collapse of a simple wave packet by performing
a measurement, the conservation of the average momentum is valid only if we include
the measuring device, which absorbs the appropriate amount of momentum, to make the
average momentum of the complete system zero.
We turn now to the last critique presented in [35, 40], namely that the classical equation
of motion for the axion star precludes a linear coupling to a scattering state axion. In [35],
the authors write about our previous work [34] as follows:
They expanded the scalar axion eld
around the classical eld 0:
where ~ is the quantum
proportional to
uctuation
eld. The Hamiltonian includes a term
03 ~ from the axion interaction potential. They claimed that
this term produces transitions of 3 condensate axions into one relativistic axion
of energy 3 ma. However, the sum of all terms in the Hamiltonian that are
linear in ~ is zero by the classical equations of motion for 0. Thus the matrix
element for producing a single relativistic axion is 0.
= 0 + ~,
On the contrary, we have emphasized (and explain further in appendix B) that a
complete set of states satisfying the Klein-Gordon equation for the axion
eld contains
both bound and scattering state components,
=
b +
s
. Our expansion does not
distinguish a classical and a quantum component, and as such, one cannot identify
b as
0, a purely classical eld. By direct calculation we see that the matrix element for (A.1)
is nonzero and is proportional to
shN j b3jN
We can see this more quantitatively by analyzing the equation of motion. The standard
procedure to calculate the equation of motion is to take the matrix element
of the Klein-Gordon operator KG[ ]. This is precisely what we have used in eq. (3.3) and
is equivalent to the Gross-Pitaevskii equation used by numerous other authors, e.g. [25, 27,
56]. In an \exact" eld theory, any matrix element of the form h jKG[ ]j 0i = 0, including
hN jKG[ ]jN
1i = 0
hN jKG[ ]jN
However, because this exact theory is not known, one is restricted to using some ansatz
for the eld
. We have chosen the Ru ni-Bonazzola ansatz, expanded to include the
scattering state solutions, which was given in eq. (2.1). With this choice, the wavefunction
R(r) (which is equivalent to the Gross-Pitaevskii wavefunction), does not satisfy eq. (A.5),
and in this same parameterization, the matrix element for the process (A.1) is nonzero
as well.
We are working to extend the Ru ni-Bonazzola paradigm so that eq. (A.5), as well
as higher-order expressions in the operator Klein-Gordon equation, can be simultaneously
satis ed. In this extension, we nd that the rate for the process (A.1) is still nonzero, and
is equal at leading order to the result we obtain here. We will present these and related
ndings in a future publication.
B
The continuous spectrum in the Ru ni-Bonazzola equations
The most general bound solution, discussed in [39], for a real scalar eld in spherically
symmetric metric of eq. (3.4) and satisfying the non-interacting Klein-Gordon equation is
b =
X cnlmRnl(r) Ylm( ; ) e i Enl t + c:c:
nlm
where cnlm are arbitrary amplitudes of the nth bounds state solution with angular
momentum l. The wave functions Rnl(r) satisfy wave equations
Rn00l =
2
r
+
B0
2 B
A0
2 A
Rn0l
A
Enl2
B
m2
l(l + 1)
r2
Rnl:
(B.2)
(A.4)
anlm and cnlm ! aynlm; these operators satisfy
To be able to describe condensates of bosons, Ru ni and Bonazzola [39] introduced second
quantization by promoting coe cients cnlm to creation and annihilation operators cnlm !
[anlm; ayn0l0m0 ] = nn0 ll0 mm0 :
Now notice that b is not the most general solution of the wave equation. Because the
attractive gravitational potential, which makes bound states solutions possible, vanishes at
large r, (B.2) has scattering state solutions as well. After quantization these states can be
written as
where k =
s =
The complete set of solutions to the Klein-Gordon equation is
b +
s
.
The Ru ni-Bonazzola method was used by Barranco and Bernal [26] and by us [24] to
describe axion stars. It is su cient to use of quantized eld
b in leading order to describe
static axion stars. However, the term
s becomes signi cant if we notice that in rst order
of the expansion of the axion potential using
results in an operator
k0):
=
f
b :
m
16
X(r)3
(B.7)
(B.8)
(B.9)
Li =
This has been described in detail in [34] and also applied in the present paper.
The second subject we discuss in this appendix is the reason why free spherical waves,
fl(k) ! jl(k r), can be used to calculate the decay rate of axion stars. This is the
consequence of the fact that dilute axion stars, those we consider in this work, are weakly bound.
In fact, only axion stars whose particle energy satis es jE0
mj . 0:002 m, corresponding
to a value of
:05, could survive from the big bang until the present epoch.
To further justify our use of free spherical waves in the
nal state, consider (3.3)
and (3.5). Using Y (x) = X(r)=
and x = m r , where
Klein-Gordon equation in (3.5) can be rewritten as the nonlinear Schrodinger equation
2 '
2 (E0
m)=m, the
(E0
m) X(r) =
X00(r) +
X0(r) +
(E0
m) b(r) X(r) +
where we de ned an e ective potential
1
2 m
1
2 m
2
r
2
r
X00(r) +
X0(r) + Ve (r) X(r);
Ve (r)
(E0
m) b(r) +
= (E0
m)
b(r)
m
16
1
8
X(r)2
Y (x)2 :
Now considering that b(r) < 0, b0(r) > 0 and X0(r) < 0 over the whole range of r, as shown
by the numerical calculations [24], the e ective potential Ve satis es
3 (E0
m) . Ve (r)
0
The quantity 3 (E0
m) &
0:006 m at
. :05. This shows that the depth of the
e ective potential is of O(10 3 m) or smaller, implying that produced relativistic axions,
which have energy E ' 3 m, can well be regarded as free. The situation would be di erent
in a discussion of dense axion star states, where the binding energy is much more signi cant;
we will return to this topic in a future work.
Open Access.
This article is distributed under the terms of the Creative Commons
Attribution License (CC-BY 4.0), which permits any use, distribution and reproduction in
any medium, provided the original author(s) and source are credited.
HJEP01(28)6
Lett. 38 (1977) 1440 [INSPIRE].
instantons, Phys. Rev. D 16 (1977) 1791 [INSPIRE].
[2] R.D. Peccei and H.R. Quinn, Constraints imposed by CP conservation in the presence of
103 [INSPIRE].
Lett. 40 (1978) 279 [INSPIRE].
harmless axion, Phys. Lett. 104B (1981) 199 [INSPIRE].
[5] M. Dine, W. Fischler and M. Srednicki, A simple solution to the strong CP problem with a
[6] A.R. Zhitnitsky, On possible suppression of the axion hadron interactions, (in Russian),
Sov. J. Nucl. Phys. 31 (1980) 260 [Yad. Fiz. 31 (1980) 497] [INSPIRE].
Phys. Lett. 165B (1985) 71 [INSPIRE].
[11] T. Banks, M. Dine, P.J. Fox and E. Gorbatov, On the possibility of large axion decay
constants, JCAP 06 (2003) 001 [hep-th/0303252] [INSPIRE].
[12] J. Preskill, M.B. Wise and F. Wilczek, Cosmology of the invisible axion, Phys. Lett. 120B
[13] L.F. Abbott and P. Sikivie, A cosmological bound on the invisible axion, Phys. Lett. 120B
(1983) 127 [INSPIRE].
(1983) 133 [INSPIRE].
123509 [arXiv:1603.04249] [INSPIRE].
Rev. D 27 (1983) 995 [INSPIRE].
[14] S. Davidson and T. Schwetz, Rotating drops of axion dark matter, Phys. Rev. D 93 (2016)
[15] M. Dine and W. Fischler, The not so harmless axion, Phys. Lett. 120B (1983) 137
[16] R. Holman, G. Lazarides and Q. Sha , Axions and the dark matter of the universe, Phys.
[INSPIRE].
103 (2009) 111301 [arXiv:0901.1106] [INSPIRE].
3051 [hep-ph/9303313] [INSPIRE].
025001 [hep-ph/9901396] [INSPIRE].
HJEP01(28)6
043525 [arXiv:1001.1769] [INSPIRE].
121801 [arXiv:1512.00108] [INSPIRE].
[22] J. Barranco, A.C. Monteverde and D. Delepine, Can the dark matter halo be a collisionless
ensemble of axion stars?, Phys. Rev. D 87 (2013) 103011 [arXiv:1212.2254] [INSPIRE].
[arXiv:1411.3900] [INSPIRE].
JHEP 03 (2015) 080 [Erratum ibid. 11 (2016) 134] [arXiv:1412.3430] [INSPIRE].
distribution of coherent oscillations of a primordial axion
Nucl. 59 (1996) 1005 [Yad. Fiz. 59 (1996) 1050] [INSPIRE].
[32] M.Y. Khlopov, A.S. Sakharov and D.D. Sokolo , The nonlinear modulation of the density
distribution in standard axionic CDM and its cosmological impact, Nucl. Phys. Proc. Suppl.
B 72 (1999) 105.
[INSPIRE].
[33] I.I. Tkachev, On the possibility of Bose star formation, Phys. Lett. B 261 (1991) 289
[34] J. Eby, P. Suranyi and L.C.R. Wijewardhana, The lifetime of axion stars, Mod. Phys. Lett.
A 31 (2016) 1650090 [arXiv:1512.01709] [INSPIRE].
[35] E. Braaten, A. Mohapatra and H. Zhang, Emission of photons and relativistic axions from
axion stars, Phys. Rev. D 96 (2017) 031901 [arXiv:1609.05182] [INSPIRE].
(1980) 253 [INSPIRE].
JHEP 01 (2016) 034 [arXiv:1511.02867] [INSPIRE].
(2000) 1158 [astro-ph/0003365] [INSPIRE].
(1983) 1243 [INSPIRE].
[39] R. Ru ni and S. Bonazzola, Systems of selfgravitating particles in general relativity and the
concept of an equation of state, Phys. Rev. 187 (1969) 1767 [INSPIRE].
[40] E. Braaten, H.-W. Hammer and G.P. Lepage, Open e ective eld theories from highly
inelastic reactions, PoS(ICHEP2016)350 [arXiv:1612.08047] [INSPIRE].
Rev. D 50 (1994) 3650 [hep-ph/9205208] [INSPIRE].
[INSPIRE].
[INSPIRE].
[INSPIRE].
[47] L. Amendola and R. Barbieri, Dark matter from an ultra-light pseudo-Goldsone-boson,
Phys. Lett. B 642 (2006) 192 [hep-ph/0509257] [INSPIRE].
[48] B. Li, T. Rindler-Daller and P.R. Shapiro, Cosmological constraints on
Bose-Einstein-condensed scalar eld dark matter, Phys. Rev. D 89 (2014) 083536
[arXiv:1310.6061] [INSPIRE].
[49] H.-Y. Schive, T. Chiueh and T. Broadhurst, Cosmic structure as the quantum interference
of a coherent dark wave, Nature Phys. 10 (2014) 496 [arXiv:1406.6586] [INSPIRE].
[51] J.-W. Lee, Brief history of ultra-light scalar dark matter models, arXiv:1704.05057
[arXiv:1704.07286] [INSPIRE].
[hep-ph/9507385] [INSPIRE].
[52] V. Sahni and L.-M. Wang, A new cosmological model of quintessence and dark matter,
Phys. Rev. D 62 (2000) 103517 [astro-ph/9910097] [INSPIRE].
[53] M. Alcubierre et al., Galactic collapse of scalar eld dark matter, Class. Quant. Grav. 19
(2002) 5017 [gr-qc/0110102] [INSPIRE].
[54] A. Maselli, P. Pnigouras, N.G. Nielsen, C. Kouvaris and K.D. Kokkotas, Dark stars:
gravitational and electromagnetic observables, Phys. Rev. D 96 (2017) 023005
[55] J.-w. Lee and I.-g. Koh, Galactic halos as boson stars, Phys. Rev. D 53 (1996) 2236
043531 [arXiv:1103.2050] [INSPIRE].
Bose-Einstein condensates with short-range interactions: II. Numerical results, Phys. Rev.
D 84 (2011) 043532 [arXiv:1103.2054] [INSPIRE].
HJEP01(28)6
selfgravitating scalar eld, Phys. Rev. D 69 (2004) 124033 [gr-qc/0404014] [INSPIRE].
[60] F.S. Guzman and L.A. Urena-Lopez, Gravitational cooling of self-gravitating
Bose-Condensates, Astrophys. J. 645 (2006) 814 [astro-ph/0603613] [INSPIRE].
[61] A. Bernal and F.S. Guzman, Scalar eld dark matter: non-spherical collapse and late time
behavior, Phys. Rev. D 74 (2006) 063504 [astro-ph/0608523] [INSPIRE].
[62] A. Arvanitaki, M. Baryakhtar and X. Huang, Discovering the QCD axion with black holes
and gravitational waves, Phys. Rev. D 91 (2015) 084011 [arXiv:1411.2263] [INSPIRE].
[63] A. Arvanitaki, M. Baryakhtar, S. Dimopoulos, S. Dubovsky and R. Lasenby, Black hole
mergers and the QCD axion at advanced LIGO, Phys. Rev. D 95 (2017) 043001
[arXiv:1604.03958] [INSPIRE].
[64] J. Eby, M. Leembruggen, P. Suranyi and L.C.R. Wijewardhana, Collapse of axion stars,
JHEP 12 (2016) 066 [arXiv:1608.06911] [INSPIRE].
[65] V. Irsic, M. Viel, M.G. Haehnelt, J.S. Bolton and G.D. Becker, First constraints on fuzzy
dark matter from Lyman- forest data and hydrodynamical simulations, Phys. Rev. Lett.
119 (2017) 031302 [arXiv:1703.04683] [INSPIRE].
[66] A. Diez-Tejedor and D.J.E. Marsh, Cosmological production of ultralight dark matter
axions, arXiv:1702.02116 [INSPIRE].
(1997) 1353 [cond-mat/9601150].
[67] H.T.C. Stoof, Macroscopic quantum tunneling of a bose condensate, J. Stat. Phys. 87
[68] C.A. Sackett, H.T.C. Stoof and R.G. Hulet, Growth and collapse of a Bose condensate with
attractive interactions, Phys. Rev. Lett. 80 (1998) 2031 [cond-mat/9801014].
[69] T. Harko, Gravitational collapse of Bose-Einstein condensate dark matter halos, Phys. Rev.
D 89 (2014) 084040 [arXiv:1403.3358] [INSPIRE].
[70] P.-H. Chavanis, Collapse of a self-gravitating Bose-Einstein condensate with attractive
self-interaction, Phys. Rev. D 94 (2016) 083007 [arXiv:1604.05904] [INSPIRE].
[71] D.G. Levkov, A.G. Panin and I.I. Tkachev, Relativistic axions from collapsing Bose stars,
Phys. Rev. Lett. 118 (2017) 011301 [arXiv:1609.03611] [INSPIRE].
[72] T. Helfer, D.J.E. Marsh, K. Clough, M. Fairbairn, E.A. Lim and R. Becerril, Black hole
formation from axion stars, JCAP 03 (2017) 055 [arXiv:1609.04724] [INSPIRE].
[73] J. Eby, M. Leembruggen, P. Suranyi and L.C.R. Wijewardhana, QCD axion star collapse
with the chiral potential, JHEP 06 (2017) 014 [arXiv:1702.05504] [INSPIRE].
(2017) 122 [arXiv:1612.07750] [INSPIRE].
Cim. 4 (1956) 848 [INSPIRE].