rapidly becomes independent of any initial velocity. The
second-order differential equation of motion for ζ
k
giv es rise
to two different time-dependent pieces in the solution. These
are commonly designated the “gro wing” and “decaying”
modes, although in the case of standard slow-roll the “gro w-
ing” mode actually approaches a constant for k ≪ aH while
the decaying mode rapidly becomes a negligible component of
the solution. Then ζ
k
can be considered constant (and
_
ζ
k
≈ 0)
roughly from the time k ≲ aH ≡ −1=η
⋆
until the end of the
inflationary phase, η ¼ 0. The observed curvature pertu-
rbation, ζ
k
, and its conjugate momentum, π
−
k
, sati-
sfy ½lim
kη→0
−
ζ
k
ðηÞ; lim
kη→0
−
π
−
k
ðηÞ ¼ 0, making quantum
mechanical effects extremely diff icult to observe even in the
absence of any sources of decoherence. Since ϵ is nearly
constant, the canonical field χ satisfie s Eq. (2) with z
0
=z ¼
a
0
=a to a very good approximation.
A rather different behavior can be found if the potential
for the scalar field has an exactly flat region, but the energy
density is driven to evolve by giving the scalar field an
initial velocity [27]. Such a phase, often called a “non-
attractor phase” would only persist a short time, since the
initial velocity is damped away by Hubble friction. But,
while it lasts, it provides a background metric that is nearly
de Sitter, but with ϵ ∼ a
−6
ðηÞ far from constant and δ ¼ −6
not a small parameter. Assuming c
s
is a constant,
z
0
=z ¼ −2a
0
=a. Crucially for our purposes, a change in
dynamics affects the solution for ζ
k
, which now has one
contribution that grows as η
−3
and a second that is constant,
so that
_
ζ
k
≠ constant even for k ≲ aH.
For either slow-roll or nonattractor dynamics the evolution
of the canonical field χ at quadratic order can be found in
terms of two-mode squeezing and rotation operators. One
first needs to solve for the time-dependence of the ladder
operators using the Heisenberg equation of motion,
d
ˆ
c
k
dη
¼ −i½
ˆ
c
k
;
ˆ
H¼−i
c
s
k
ˆ
c
k
þ i
z
0
z
ˆ
c
†
−
k
; ð3Þ
which can in turn be solved by a Bogoliubov transformation
with a choice of initial condition at time η
0
,
ˆ
c
k
ðηÞ¼
u
k
ðηÞ
ˆ
c
k
ðη
0
Þþv
k
ðηÞ
ˆ
c
†
−
k
ðη
0
Þ, where ju
k
ðηÞj
2
−jv
k
ðηÞj
2
¼1.
The Bogoliubov transformation can be written as
u
k
ðηÞ¼e
iθ
k
ðηÞ
cosh r
k
ðηÞ; ð4aÞ
v
k
ðηÞ¼e
−iθ
k
ðηÞþ2iϕ
k
ðηÞ
sinh r
k
ðηÞ; ð4bÞ
where r
k
is the squeezing parameter, ϕ
k
is the squeezing
angle, and θ
k
is an angle rotating the conjugate field and
momenta (which is the same for the
k and −
k modes). The
leading order time-dependence in the exact de Sitter back-
ground approximation for slow-roll inflation, is given by [26]
r
SR
k
ðηÞ¼−ArcSinh
1
2c
s
kη
; ð5aÞ
ϕ
SR
k
ðηÞ¼−
π
4
−
1
2
ArcTan
1
2c
s
kη
; ð5bÞ
θ
SR
k
ðηÞ¼−kη − ArcTan
1
2c
s
kη
; ð5cÞ
while for the nonattractor case we find instead
r
NA
k
ðηÞ¼−2ArcSinh
3
2c
s
kη
; ð6aÞ
ϕ
NA
k
ðηÞ¼
π
4
−
1
2
ArcTan
3
2c
s
kη
; ð6bÞ
θ
NA
k
ðηÞ¼−kη − 2
ffiffiffi
2
p
ArcTan
3
ffiffiffi
2
p
2c
s
kη
: ð6cÞ
For nonattractors the solution is approximate and only valid
when c
s
kη ≪ 1.
Note that the equation of motion for χ is the same
in both, slow-roll and nonattractor models, i.e., χ
00
k
þ
ðc
2
s
k
2
−
2
η
2
Þχ
k
¼ 0. At any instant of time, however, the
position χ
k
and conjugate momentum p
−
k
can be different
between the two models, as indicated by the distinct time-
dependence of the squeezing parameters above. The
commutator, ½χ
k
ðηÞ;p
−
k
0
ðηÞ ¼ ið2πÞ
3
δ
3
ð
k −
k
0
Þ, remains
preserved at all times, as expected.
IV. DEFINING THE SYSTEM AND BATH
We use bands of comoving momenta to define the
system and bath, and assume that at some initial (con-
formal) time η
0
we can factorize the Hilbert space as
1
H ¼ H
UV
⊗ H
Obs
⊗ H
NIR
⊗ H
IR
: ð7Þ
We focus only on how the evolution of modes in H
Obs
,
which satisfy k
min
≤ k
Obs
≤ k
max
, is affected by interactions
with unobservable modes in the near infrared H
NIR
, which
satisfy k
IR
<k
NIR
<k
min
. We assume that modes in H
UV
can be properly accounted for with usual renormalization
techniques. We also assume that modes far in the infrared,
k ∈ IR, were accounted for in defining the Hamiltonian at
time η
0
. This organization is shown diagrammatically in
Fig. 1. Note that the comoving Hubble radius decreases as a
function of time.
Further, we consider a cubic interaction term of the
form, S
3
¼ M
2
p
R
d
3
xdηa
4
ð3ϵ=c
2
s
Þðc
2
s
− 1Þζ
_
ζ
2
. Expressed
in terms of the field χ in momentum space this reads
1
If we worked with physical momenta such a factorization
should remain valid at all times, but, since we are using comoving
momenta and the scale factor a is fluctuating, the factorization
would not quite hold. However, this prescription should capture the
dominant features of the scenario (and we can check the physics
using what is known about gauge issues for ζ
k
correlators).
OPEN QUANTUM COSMOLOGICAL SYSTEM PHYS. REV. D 98, 083535 (2018)
083535-3