PreprintPDF Available

# Computer-assisted methods for analyzing periodic orbits in vibrating gravitational billiards

Authors:

## Figures

Computer-assisted methods for analyzing periodic orbits in
vibrating gravitational billiards
Kevin E. M. Church
Department of Mathematics and Statistics, McGill University
805 Sherbrooke St W, Montreal, Quebec, H3A 0B9, Canada
kevin.church@mcgill.ca
Cl´ement Fortin
Department of Physics, McGill University
805 Sherbrooke St W, Montreal, Quebec, H3A 0B9, Canada
clement.fortin@mail.mcgill.ca
Using rigorous numerical methods, we prove the existence of 608 isolated periodic orbits in a
gravitational billiard in a vibrating unbounded parabolic domain. We then perform pseudo-
arclength continuation in the amplitude of the parabolic surface’s oscillation to compute large,
global branches of periodic orbits. These branches are themselves proven rigorously using
computer-assisted methods. Our numerical investigations strongly suggest the existence of mul-
tiple pitchfork bifurcations in the billiard model. Based on the numerics, physical intuition and
existing results for a simpliﬁed model, we conjecture that for any pair (k, p), there is a constant
ξfor which periodic orbits consisting of kimpacts per period pcan not be sustained for am-
plitudes of oscillation below ξ. We compute a veriﬁed upper bound for the conjectured critical
amplitude for (k, p) = (2,2) using our rigorous pseudo-arclength continuation.
Keywords: gravitational billiards, time-varying domain, periodic orbit, rigorous numerics, nu-
merical continuation
1. Introduction
Adynamical billiards is a Hamiltonian dynamical system, or an abstraction of the game billiards (or pool).
A particle moves according to some equations of motion until it hits the boundary of the domain, at which
point it undergoes specular reﬂection. Contingent upon the case study, this reﬂection can either be elastic
or inelastic. Though the dynamics can seem trivial, these systems are known to be chaotic [Chernov &
Markarian, 2003], depending on the geometry of the boundary. A number of studies [Baxter & Umble,
2007; Biswas, 1997; Boshernitzan et al., 1998; Troubetzkoy, 2005] have focused on studying the dynamics
of such systems in closed polygonal domains. Some other “billiard-like” systems have been considered in
the context of switched and impacting systems [Huang & Fu, 2019; Huang & Luo, 2017; Tang et al., 2019].
In some cases, a potential ﬁeld is incorporated, allowing for periodic orbits to exist in open domains, with
parabolic, wedge-shaped or hyperbolic boundaries. When the potential corresponds to the gravitational
ﬁeld, the dynamical system is called gravitational billiards. The chaotic behavior of these systems are well-
studied both analytically and numerically [Chatterjee et al., 1996; M´aty´as & Barna, 2011; Peraza-Mues
et al., 2017; Korsch & Lang, 1991] and often categorize the stability of periodic orbits. However, only a
few papers have studied the dynamics of a particle in a spatially time-varying domain [Hartl et al., 2013;
1
2K. E. M. Church &C. Fortin
Feldt & Olafsen, 2005; Peraza-Mues et al., 2017; Costa et al., 2015; Langer & Miller, 2015] and even less
so make use of computer-assisted proofs.
In this paper, we provide a computer-assisted approach for proving existence of periodic orbits and for
studying their behavior as we vary a parameter. To showcase the strength of such approach, we study a
nonconservative boundary-particle system, with a vertically-vibrating parabolic boundary in a gravitational
ﬁeld. That is, the billiards is moving and is subject to a (gravitational) potential. We prove the existence of
at least 608 periodic orbits for speciﬁc parameters values. Moreover, strong evidence suggests the existence
of pitchfork bifurcations when doing continuation in the amplitude of oscillation of the boundary. We do not
incorporate a method for doing continuation past pitchforks but one can be found in [Lessard et al., 2017].
The code used to ﬁnd and prove results in this paper is available through the second author’s GitHub1.
An INTLAB [Rump, 1999] license is needed to complete the computer-assisted proofs.
1.1. The rigorous numerics approach to periodic orbits
Computer-assisted methods for validation of approximate numerical zeroes of nonlinear functions have
found numerous applications in nonlinear dynamics. For an overview, we refer the reader to the survey
articles [G´omez-Serrano, 2019] and [Lessard et al., 2015], where the general method for problems of ﬁnite
and inﬁnite dimensions is outlined. We describe the idea here as it applies in ﬁnite dimensions.
Let fC1U, RNwhere URN. We assume that zeroes of fin some way correspond to solutions
of our nonlinear problem. We refer to this as a zero-ﬁnding problem. We are therefore initially interested
in computing zeroes of f, or reﬁning an approximate zero using an interative procedure. The basic idea is,
of course, Newton’s method.
Deﬁnition 1.1. Let fC1U, RNwhere URNis an open set. A point ˜xUis a nondegenerate zero
of fif fx) = 0 and DF (˜x) is invertible. If fC2U, RNwhere URNis an open set and A:RNRN
is injective, we can form the Newton operator Tby
T(x):=xAf(x).
Reﬁning an approximate zero xof the map fcan then be accomplished by iterating the Newton operator T
on the approximate zero. The classical Newton’s method takes A:= (Df(x))1, but this can be computa-
tionally expensive. An alternative is to use the initial guess xXfor a zero of fand set A:= (Df(x))1.
The rigorous numerics approach is as follows.
(1) Compute an approximate zero xof f, and reﬁne it with a Newton operator until the residual ||f(x)|| is
small (i.e. near machine precision).
(2) Determine a closed ball of radius r > 0 around xon which the Newton operator Twith A:= (Df (x))1
is a contraction.
Once this is accomplished, the contraction mapping principle guarantees that fhas a unique zero in the
ball Br(x). If rcan be chosen very small, then we get a tight enclosure (i.e. our approximate zero xis
“good”). If rcan be chosen larger, then we get information about how isolated the zero is. In practice, we
compute intervals of such suitable radii rusing the radii polynomial approach: see Theorem 1 and Theorem
3. These computations are all done on a computer, and as MATLAB is used for the implementation, we
use interval arithmetic with the INTLAB [Rump, 1999] package to track roundoﬀ error and accomplish a
We make a brief remark that stability of the periodic orbits will not studied here. Veriﬁcation of stability
of periodic orbits of arbitrary period (and number of bounces) is a fair bit more diﬃcult, especially if we
wanted to match the level of mathematical rigorous with which we prove the existence of the orbits. It
would not be diﬃcult to gain some preliminary insight into stability using numerical simulations, but since
this system has so many isolated periodic orbits – see Theorem 2 – many of them are likely to be unstable,
and those stable ones might have very small basins of attraction.
1https://github.com/ClementFortin/BilliardOscillatingParabola
Computer-assisted methods for periodic orbits in vibrating gravitational billiards 3
1.2. Overview
In Section 2, we describe the vibrating parabolic gravitational billiards, establish an equivalent zero-ﬁnding
problem for its periodic orbits, and prove some elementary results concerning such periodic orbits. The
rigorous numerics approach to periodic orbits is presented in Section 3 with our main theorems concerning
the abundance of periodic orbits for speciﬁc parameter values and the rigorous branch continuation. Section
4 ends with a conclusion.
2. Vibrating Parabolic Domain in Two Dimensions
Let (x, y) be the position coordinates in R2of a particle and assume without loss of generality that its
mass is 1. Denote (v, w) such that
˙x=v,
˙y=w,
˙v= 0,
˙w=g. (1)
Let a parabolic oscillating domain P(t) be deﬁned by
P(t) = {(x, y)R2:yαx2+sin(2πωt)}(2)
where α, , ω R+describe the steepness of the curve, the amplitude of its oscillation and its frequency
of oscillation, respectively. We eliminate the frequency parameter by deﬁning the rescaled time ˆ
t=ωt,
gravity ˆg=g/ω and velocities ˆv=ωv and ˆw=ωw. Completing the change of variables and dropping the
hats, the parameterization of the boundary of P(t) becomes
y=αx2+sin(2πt),(3)
which is referred to as the oscillating surface.
From an initial condition (x0, y0, v0, w0) at time t0>0, the particle evolves according to (1). At some
time tt0and position (x(t), y(t)), the particle comes into contact with the surface, thereby changing
velocity. In other words, we have that
y(t) = αx(t)2+sin(2πt).
The particle’s new velocity is then a function of its velocity before the impact, the position at which the
impact happened as well as the time at which it happened.
When the surface is not vibrating – that is = 0 – the global attractor is always a periodic orbit or the
union of periodic orbits [Korsch & Lang, 1991]. We are interested in studying periodic orbits when  > 0.
To this end, in what follows the symbol kNwill denote the number of impacts of a periodic orbit. The
symbol ϕwill typically be used to refer to a periodic orbit (in terms of a coordinate system we will later
specify), and p > 0 its period: that is the time it takes the particle to come back to its original position
(x0, y0) with velocity (v0, w0). Because of equation (3), the period pof any periodic orbit is necessarily a
positive integer. In the following section, we build a zero-ﬁnding problem whose zeroes will encode periodic
orbits, and characterize diﬀerent types of its solutions.
2.1. Zero-Finding Problem
Let us deﬁne a coeﬃcient of restitution e[0,1] which relates the pre-impact velocity of the particle to its
post-impact velocity. For e6= 1 the collision is not perfectly elastic and a portion of the particle’s kinetic
energy is lost. The velocity component aﬀected by this collision is the one parallel to the surface’s normal
force. Let B ∈ R2×2be the linear operator that changes the usual x-y coordinate system to the surface’s
perpendicular and parallel coordinate system, such that
B=1
p1 + (2αx)21 2αx
2αx 1
4K. E. M. Church &C. Fortin
The reset law R:R4R2is obtained by applying B1EB with
E=1 0
0e
to the particle’s velocity in the surface frame of reference using a Galilean transformation, and moving
back to the particle’s reference frame so that
R=R1
R2=B1EB v
w2π cos(2πt)+0
2π cos(2πt).
More explicitly,
R(t, x, v, w) = 1
1 + (2αx)21e(2αx)22αx(1 + e)
2αx(1 + e) (2αx)2ev
w+2π cos(2πt)(e+ 1)
1 + (2αx)22αx
1(4)
where (α, e, ) are ﬁxed parameters. For presentation purposes, we denote ˙
x+= (v+, w+) = R(t, x, ˙
x)>,
the particle’s velocity vector immediately after an impact occurs.
Deﬁne tnthe time at which the nth impact occurs, xn= (xn, yn) the position of the nth impact and
˙
xn= (vn, wn) the velocity immediately before the nth impact. The periodic orbits ϕare speciﬁed by the
coordinates (tn,xn,˙
xn). We will ﬁrst derive an equation that characterizes the impact times tn. Since each
impact happens on the surface, the n+ 1th impact has to satisfy
yn+1(t) = αx2
n+1 +sin(2πtn+1),(5)
for any n∈ {0,1, ..., k 1}, where the particle’s post-impact velocity vector can be found using the reset
law. From there, the particle will evolve according to (1) and Newton’s equations of motion are used to
ﬁnd (xn+1,˙
xn+1) i.e. the coordinates of the next impact:
xn+1 =xn+v+
n(tn+1 tn),
yn+1 =yn+w+
n(tn+1 tn)g(tn+1 tn)2
2.
Substituting both of these in (5) we obtain a zero-ﬁnding problem that tn+1 must satisfy;
0 = α(xn+v+
n(tn+1 tn))2+sin(2πtn+1)ynw+
n(tn+1 tn) + g(tn+1 tn)2
2.
Since every impact happens on the surface, the height of the particle must be the same as the surface’s
such that yn=αx2
n+sin(2πtn), for any n∈ {0,1, ..., k 1}. Henceforth, let T:R5Rbe deﬁned by
T(τ, t, x, R(t, x, ˙
x)) = α[(x+R1(t, x, ˙
x)(τt))2x2] + [sin(2πτ )sin(2πt)]
− R2(t, x, ˙
x)(τt) + g(τt)2
2,(6)
where ˙
x= (v, w) and where τand tcorrespond to the time values of the next and current impact
respectively. The impact coordinates must therefore satisfy
T(t1, t0, x0,R(t0, x0,˙
x0)) = 0,
T(t2, t1, x1,R(t1, x1,˙
x1)) = 0,
.
.
.
T(tk, tk1, xk1,R(tk1, xk1,˙
xk1)) = 0.
(7)
and hence for kimpacts of total period pthere is a sequence of time values
t0< t1< ... < tk1< tk=p+t0
Computer-assisted methods for periodic orbits in vibrating gravitational billiards 5
representing the time at which each impact (xn,˙
xn) occurs (for any n∈ {0,1, ..., k 1}). In some cases,
there can be multiple diﬀerent values for tn+1 that solve T(tn+1, tn, . . . ) = 0. However, only one value
satisﬁes
tn+1 = min{t>tn:T(t, tn, xn,R(tn, xn,˙
xn)) = 0}.(8)
Physically, this condition makes sure that the particle never crosses the surface in the interval [tn, tn+1].
Once a candidate for a periodic orbit has been computed, we verify that it satisﬁes (8) using the function
verifynlssall of INTLAB [Rump, 1999; Hargreaves, 2002] which allows for rigorous enclosure of all zeros
of elementary functions within a speciﬁed interval. For a validated periodic orbit, the particle is hence
restricted to the domain P(t).
To obtain the position-velocity components of the particle for the n+ 1th impact, deﬁne the map
S:R6R4with
S(τ, t, x,R(t, x, ˙
x)) =
x+R1(t, x, ˙
x)(τt)
y+R2(t, x, ˙
x)(τt)g(τt)2
2
R1(t, x, ˙
x)
R2(t, x, ˙
x)g(τt)
,(9)
such that
[xn+1,˙
xn+1] = S(tn+1 , tn,xn,R(tn, xn,˙
xn)).
Since ynis uniquely deﬁned in terms of xnand tnaccording to (5), we need not consider ynas an explicit
variable. Therefore, periodic orbits are characterized by kimpacts, each determined by the four associated
coordinates (tn, xn, vn, wn). We hence build a map with f:ϕR4kR4ksuch that
f(ϕ):=
T(t1, t0, x0,R(t0, x0,˙
x0))
x0+R1(t0, x0,˙
x0)(t1t0)
R1(t0, x0,˙
x0)
R2(t0, x0,˙
x0)g(t1t0)
x1
v1
w1
.
.
.
T(tn+1, tn, xn,R(tn, xn,˙
xn))
xn+R1(tn, xn,˙
xn)(tn+1 tn)
R1(tn, xn,˙
xn)
R2(tn, xn,˙
xn)g(tn+1 tn)
xn+1
vn+1
wn+1
.
.
.
T(t0+p, tk1, xk1,R(tk1, xk1,˙
xk1))
xk1+R1(tk1, xk1,˙
xk1)(t0+ptk1)
R1(tk1, xk1,˙
xk1)
R2(tk1, xk1,˙
xk1)g(t0+ptk1)
x0
v0
w0
(10)
where ϕ= (t0, x0, v0, w0, . . . , tk1, xk1, vk1, wk1). Zeros of fmust satisfy (4), (7) and (9) and thus deﬁne
periodic orbits ϕof kimpacts.
Remark 2.1. is treated here as a constant but will later be considered a variable. The domain of fwill
thus change from R4kto R4k+1.
For a periodic orbit ϕto exist with e6= 1, there must be a balance between the loss in the particle’s
kinetic energy due to the coeﬃcient of restitution and the gain that the surface induces upon impact.
Hence, it should be suspected that no periodic orbits can exist if , the amplitude of the oscillating surface,
is too small.
6K. E. M. Church &C. Fortin
2.2. Symmetric periodic orbits and their properties
The following deﬁnitions specify types of periodic orbits.
Deﬁnition 2.1. We call periodic solutions with k > 1symmetric and denote them ϕsym if they have
opposite horizontal coordinates between consecutive impacts and equal vertical coordinates for all impacts.
That is,
xn+1 =xn, yn+1 =yn,
vn+1 =vn, wn+1 =wn,
for all n∈ {0,1, . . . , k 1}.
Deﬁnition 2.2. Periodic orbits that have xn=vn= 0 for all n∈ {0,1, . . . , k 1}are called trivial as they
describe a one-dimensional motion.
Intuitively, it might seem like nontrivial symmetric solutions should arise for any value of α, given ﬁxed
values of k, p, e, and . However, as the following proposition illustrates, there exists only one speciﬁc α
for which this is possible.
Proposition 1. For any pair (k, p)N2and (e, )R2, nontrivial symmetric periodic orbits can only
exist if αgp2= 2k2.
Proof. Since symmetric solutions must satisfy f(ϕsym) = 0, we have that v+
nvn+1 = 0 and hence
v+
n=vn+1 =vn. Therefore,
xn+v+
n(tn+1 tn)xn+1 =xn+1 v+
n+1(tn+2 tn+1) + xn+2 = 0,
vn+1(tn+1 tn) = vn+1 (tn+2 tn+1),
tn+1 tn=tn+2 tn+1,
where we have used the fact that xn=xn+2. By transitivity, we infer that the time between two impacts
is the same for any two impacts. Let us now determine this value. Remark that t2t1=t1t0
t2= 2(t1t0) + t0. This establishes the base case in our inductive argument. Suppose ti=i(t1t0) + t0
for some i > 0, then ti+1 ti=titi1=· · · =t1t0by the above argument. Therefore, ti+1 =
i(t1t0) + t0+t1t0= (i+ 1)(t1t0) + t0and by induction, tn=n(t1t0) + t0for all nN.
Recall that for kimpacts, one has that tk=t0+p, where pis the total period between the ﬁrst impact
and the last. Thus,
tk=t0+p=k(t1t0) + t0,
t1t0=p
k.
We now show that if xn6= 0, ϕsym can only exist for a speciﬁc value of α. From
xn+v+
n(tn+1 tn)xn+1 = 0,
yn+w+
n(tn+1 tn)g
2(tn+1 tn)2=yn=yn+1,
we get that the post-impact velocity components are
v+
n=2xnk
p=vn, w+
n=gp
2k.
Moreover, 0 = w+
ng(tn+1 tn)wn+1 =gp/2kgp/k wn+1 wn+1 =wn=gp/2k. Denote
Computer-assisted methods for periodic orbits in vibrating gravitational billiards 7
βn:= 2αxn=αvnp/k and consider the ﬁrst output of the Rmap for the nth impact,
v+
n=(1 2
n)vn+βn(1 + e)wn2πβncos(2πtn)(1 + e)
1 + β2
n
,
2xnk
p(1 + β2
n) = (1 2
n)2xnk
p(1 + e)βngp
2k2πβncos(2πtn)(1 + e),
0 = 2βnk
αp +β3
nk
αp (1 e)βngp
2k(1 + e)2πβncos(2πtn)(1 + e),
0 = βnβ2
nk
αp (1 e) + 2k
αp (1 + e)gp
2k2π cos(2πtn)(1 + e).
Since βn= 2αxnwith xn6= 0, then βn6= 0. Thus,
2π cos(2πtn)(1 + e) = β2
nk
αp (1 e) + 2k
αp (1 + e)gp
2k.(11)
We now consider the second output of the Rmap,
w+
n=βn(1 + e)vn+ (β2
ne)wn+ 2π cos(2πtn)(1 + e)
1 + β2
n
,
gp
2k(1 + β2
n) = (1 + e)β2
nk
αp + (eβ2
n)gp
2k+ 2π cos(2πtn)(1 + e),
0 = β2
n
αp(1 + e) + gp
2k(e1) β2
ngp
k+ 2π cos(2πtn)(1 + e)
and
2π cos(2πtn)(1 + e) = β2
ngp
kk
αp(1 + e)gp
2k(e1).(12)
Equating (11) and (12),
β2
nk
αp (1 e) + 2k
αp (1 + e)gp
2k=β2
ngp
kk
αp(1 + e)gp
2k(e1),
β2
nk
αp(1 e)gp
k+k
αp(1 + e)= (1 + e)gp
2k2k
αp gp
2k(e1),
β2
n2k
αp gp
k=2k
αp gp
k.
Since βnR,then β2
n0 and we must have that 2k2=αgp2.
Remark 2.2. Suppose αgp2= 2k2, so the conclusion of the previous proposition holds. We can explicitly
compute the candidates for horizontal impact positions xnbased on the associated impact times. From
(11),
2π cos(2πtn)(1 + e) = β2
nk
αp (1 e) + 2k
αp (1 + e)gp
2kwhere α=2k2
gp2, βn= 2αxn,
=8k3x2
n
gp3(1 e) + gp
2k(1 e).(13)
Solving for xnwe obtain
xn=±s2π cos(2πtn)(1 + e) + gp
2k(e1)gp3
8k3(1 e).(14)
Now suppose for example that k= 2. Since x0=x1for a nontrivial symmetric solution and t1=t0+p
for some positive integer p, it suﬃces to choose some t0such that the radical of the above equation (for
n= 0) is positive. The particular situation where (e, ) = (1,0) is treated in [Korsch & Lang, 1991].
8K. E. M. Church &C. Fortin
Remark 2.3. The time diﬀerence of two subsequent impacts p/k has to be a multiple of the surface’s period
of oscillation, which is given by 1. This comes from the deﬁnition of a symmetric solution as well as the
injective (bijective) behavior of the reset law when considering one side of the y-axis. Physically, the particle
has to bounce back from a given position (x, y) with a speciﬁc velocity vector induced by the surface’s
normal force. Hence, p/k N.
Referring to Remark 2.2, suppose we restrict to the case k= 2. There is a continuous interval of time
values for which the radical of (14) is nonnegative. Since these time values uniquely determine symmetric
solutions with k= 2 impacts per period, such solutions are parameterized by a continuous parameter,
namely t0. As such, these symmetric solutions are not isolated and hence do not qualify as nondegenerate
zeros of f, see Deﬁnition 1.1. This hinders the process of identifying such solutions as the injective linear
operator Agiven by the inverse of the Jacobian will suﬀer from numerical instabilities. A similar argument
applies for nontrivial symmetric solutions with a generally even number of impacts per period. For this
reason, we will not be searching for nontrivial symmetric periodic orbits.
We now characterize the existence of trivial ϕsym using an adapted result of [Luo & Han, 1996] from
the Period-1 Motion section.
Proposition 2. Trivial (symmetric) solutions either exist in pair or do not exist. Additionally, if the
periodic orbit’s impacts are not perfectly elastic, such solutions do not exist for
0,gp
4πk
1e
1 + e.(15)
Proof. By deﬁnition, a trivial solution satisﬁes xn= 0, vn= 0, wn=gp/2ksuch that
R2t0,0,0,gp
2k=gp
2k=egp
2k+ 2π cos(2πt0)(1 + e),
cos(2πt0) = gp
4πk
1e
1 + e.(16)
for all n∈ {0,1, . . . , k 1}. Since the coeﬃcient of restitution is smaller or equal to one, the last equation’s
right-hand side is always nonnegative. Therefore, whenever (k, p) are such that the latter is smaller than
one, exactly two diﬀerent values of t0will satisfy this equation:
t0=1
2πarccos gp
4πk
1e
1 + e, t0= 1 1
2πarccos gp
4πk
1e
1 + e.
We now consider the second claim. If = 0, then necessarily e= 1 and the motion is perfectly elastic. If
instead  > 0, we use (16) to get
2π(1 + e)2πcos(2πt0)(1 + e) = gp
2k(1 e),
from which we conclude that must satisfy (15).
3. Computer-assisted proof of solutions and branches
With the equivalence between the zero-ﬁnding problem (10) and periodic orbits established, we will now
show how to prove the existence of periodic orbits. We start with isolated zeroes in Section 3.1 before
moving to branches in Section 3.2.
3.1. Isolated periodic solutions
To ﬁnd a solution, we use Newton’s method in higher dimensions. We start with an initial guess ϕ(0) for a
solution of the fmap. Newton’s method deﬁnes a sequence of vectors {ϕ(0), ϕ(1) ,...ϕ(n), ϕ(n+1), . . . }such
that
ϕ(n+1) =ϕ(n)(Df (ϕ(n)))1f(ϕ(n)).
Computer-assisted methods for periodic orbits in vibrating gravitational billiards 9
If the zero-ﬁnding problem evaluated at ϕ(n+1) is within a certain tolerance of zero such that kf(ϕ(n+1))k<
tol, we stop the process and denote ¯ϕ=ϕ(n+1) a numerical solution of the system. The nonzero partial
derivatives of the Jacobian matrix Df are shown in Appendix A.
The following theorem will be used to construct an enclosure of an approximate zero ϕof the map f
of (10) – that is, a closed ball Br(ϕ) in which fhas a unique zero.
Theorem 1. Let Xbe a Banach space and T:XXa Frchet diﬀerentiable mapping. Let x0X, and
suppose that r, Y, Z > 0have
kT(x0)x0kXY, (17)
kDT (z)kB(X)Z, where zBr(x0).(18)
p(r)=(Z1)r+Y.
If there exists r0such that rr0>0and p(r0)<0,then there is a unique ﬁxed point ˜xBr0(x0)of
the contraction T.
Proof. Assume that rr0>0 has that p(r0)<0. Then,
Zr0+Y < r0,
and since r06= 0 we have that
Z+Y
r0
<1.
Since Y, r0are positive we have that Z < 1. For x, y Br0(x0) we have that
kT(x)T(y)kXsup
zBr0(x0)
kDT (z)kB(X)kxykXZkxykX.
Since Z < 1, Tis a contraction on Br0(x0). To see that Tmaps the closed ball into itself, choose xBr0(x0).
Then,
kT(x)x0kX≤ kT(x)T(x0)kX+kT(x0)x0kX
sup
zBr0(x0)
kDT (z)kB(X)kxx0kX+Y
Zr0+Y
< r0.
It follows from the contraction mapping theorem that there exists a unique ﬁxed point ˜xBr0(x0) of T.
In this case, the contraction mapping T:R4kR4kis deﬁned as T(ϕ):=ϕAf(ϕ), where R4k
equipped with the sup-norm is a Banach space. Remark that fC, which allows us to set A:=
(Df ( ¯ϕ))1and Y:=kDf( ¯ϕ)1f( ¯ϕ)kwhere we compute Anumerically for each ¯ϕR4k. In particular,
the invertibility of the linear operator A yields that D T ( ¯ϕ) is well-deﬁned over all of R4k. With this in
mind, deﬁne Br( ¯ϕ)R4ka ball of radius rcentered at ¯ϕwith r>0. Since ris ﬁnite, Br( ¯ϕ)R4kis
a compact subset and DT is bounded over Br( ¯ϕ) by the extreme value theorem. We can now set
Z:= sup
zBr( ¯ϕ)
kDT (z)k, Y :=kT( ¯ϕ)¯ϕk(19)
with the radii polynomial given by
p(r)=(Z1)r+Y.
10 K. E. M. Church &C. Fortin
Applying Theorem 1 and denoting rr0>0 such that p(r0)<0, we conclude that there exists a unique
equilibrium solution ˜ϕin the ball Br( ¯ϕ) for any rI= [r0, r], where Iis called the existence interval.
Lastly, ˜ϕsatisﬁes f( ˜ϕ) = 0 with Df ( ˜ϕ) invertible. In practice, Zis computed with INTLAB by computing
DT (z) for a thick interval zcorresponding to the ball Br(ϕ).
Since a periodic orbit ϕis equivalent (by the zero-ﬁnding problem) to a set of vectors, one can rep-
resent it in diﬀerent ways that are in some sense equivalent. The following proposition characterizes this
equivalence.
Proposition 3. An asymmetric periodic orbit ϕwith kimpacts has kequivalent ways of being represented
at the dynamics level, in the sense that these orbits trace the same path in state space. If ϕis a nontrivial
periodic orbit, it is equivalent to a family of 2knontrivial periodic orbits by way of horizontal reﬂections
and cyclic permutations of coordinates in terms of the variables of the fmap.
Proof. Let ϕR4ka periodic orbit representing the sequence of vectors {φ0, φ1, . . . , φk1}where φn=
(tn, xn,˙
xn). Since there are kmany vectors, there are kcyclic permutations of the sequence, namely
{φ0, φ1, . . . , φk2, φk1},
{φ1, φ2, . . . , φk1, φ0},
.
.
.
{φk1, φ0, . . . , φk3, φk2},
where the time values are modiﬁed such that t0[0,1). The ordered sequence of impact coordinates shown
above can be found by applying the maps R,Tand Siteratively. Since the path drawn by the particle
throughout the periodic orbit is the same for each of the above sequences, the latter all have equivalent
dynamics. In the case where ϕis not a trivial periodic orbit, each sequence has a reﬂection with respect to
the y axis, see Figure 1 for a visual representation. However, a sequence and its reﬂection do not draw the
same path in parameter space. Thus they do not have equivalent dynamics and are only equivalent at the
fmap’s level.
Fig. 1. Simpliﬁed visual representation of two nontrivial periodic orbits that are the reﬂection of one another. Three permu-
tations are possible as there are k= 3 impacts. The surface’s oscillation is not shown for presentation purposes.
Remark 3.1. The Zbound depends on the size of r, thus the choice of rwill be case dependent so as to
maximize the existence interval of the unique equilibrium solution ˜ϕ. In this case, r0provides tight bounds
on the location of the real unique solution ˜ϕ, whereas rcorresponds to the latter’s domain of isolation.
With the method outlined above, we rigorously compute solutions of the fmap for diﬀerent combinations
of the parameters (α, e, ) and (k, p). Since periodic orbits of one impact only bounce vertically at x= 0,
we consider cases with k > 1. To ﬁnd solutions, we generate random guesses that live in 4kdimensions.
We will restrict the search to two impacts, as it requires less computation and facilitates presentation.
Computer-assisted methods for periodic orbits in vibrating gravitational billiards 11
However, our code is general, and using it together with Theorem 1, we are able to prove solutions for any
(k, p).
3.1.1. Results for parameters g= 9.81 and (α, e, ) = (2,0.778,1)
Let g= 9.81 and (α, e, ) = (2,0.778,1) such that nontrivial ϕsym do not exist. Since t0= 0 and t2=p+t0,
the zero-ﬁnding problem is of eight dimensions. Indeed, the variables are ϕ= (t0, x0, v0, w0, t1, x1, v1, w1)
with y0=αx2
0+sin(2πt0) and y1=αx2
1+sin(2πt1). The zero-ﬁnding problem (10) then becomes
f=
T(t0, x0,R(t0, x0,˙
x0))
x0+R1(t0, x0,˙
x0)(t1t0)
R1(t0, x0,˙
x0)
R2(t0, x0,˙
x0)g(t1t0)
x1
v1
w1
T(t1, x1,R(t1, x1,˙
x1))
x1+R1(t1, x1,˙
x1)(p+t0t1)
R1(t1, x1,˙
x1)
R2(t1, x1,˙
x1)g(p+t0t1)
x0
v0
w0
.(20)
To ﬁnd zeroes, we apply Newton’s method numerically until fattains a certain tolerance. In this case,
we set kf( ¯ϕ)k<tol such that tol = 1013. We search for solutions that have periods between p= 1 and
p= 10, as there is an abundance of orbits for these values. Furthermore, for a ﬁxed period, we focus on
p= 2 as it is the least value that allows for (trivial) symmetric solutions. For these cases, we give a lower
bound on the total number of periodic orbits that exist. The plots in Figure 2 correspond to one complete
period of solutions that have (k, p) = (2,2).
Fig. 2. Trajectory of periodic orbits (red) bouncing on the oscillating surface and its corresponding impact coordinates. Left:
nontrivial (asymmetric) orbits. Right: trivial orbits. The uniqueness intervals I= [Imin, Imax] all have Imin 1.9·1010 and
Imax 8.5×106. Their coordinates are provided in Table 3.1.1.
With Proposition 3 in mind and the computer-assisted proof complete, we are able to give a more
accurate lower bound on the number of periodic orbits there are for the parameter set in question.
Theorem 2. For g= 9.81,(α, e, ) = (2,0.778,1) and k= 2, there are at least 608 periodic orbits with
12 K. E. M. Church &C. Fortin
Table 1. Coordinates (in impact time and state-space location) of the periodic orbits from Figure 2.
Figure t0x0v0w0t1x1v1w1
Top Left 0.76171 -0.10155 -0.25742 -9.64597 0.84303 0.39236 6.07384 5.14225
Top Right 0.23446 0 0 -4.905 1.23446 0 0 -4.905
Bottom Left 0.73258 1.65511 3.31680 4.23915 2.21578 -0.05902 -1.15570 -9.63543
Bottom Right 0.09841 0 0 2.60147 1.73768 0 0 -9.00244
their ﬁrst bounce t0occurring in the interval [0,1), with
8periodic orbits for p= 1,40 periodic orbits for p= 2,
42 periodic orbits for p= 3,46 periodic orbits for p= 4,
52 periodic orbits for p= 5,70 periodic orbits for p= 6,
68 periodic orbits for p= 7,92 periodic orbits for p= 8,
92 periodic orbits for p= 9,98 periodic orbits for p= 10.
In a given class of (k, p), these periodic orbits are isolated.
The data (e.g. coordinates, radii of isolation) associated to these periodic orbits can be found at the
second author’s GitHub2. Note that there exist periodic orbits for periods larger than p= 10 as well. For
each period p, equation (16) is used to verify the existence of a pair of trivial symmetric orbits. As per
Proposition 3, some of the asymmetric solutions of the theorem trace out the same path in the phase space.
However, in the extended phase space consisting of spatial and time coordinates (with initial impact times
normalized to t0[0,1)), they are distinct.
3.2. Branches of periodic orbits: rigorous parameter continuation
After rigorously identifying periodic orbits of the fmap, one wonders how these solutions behave as we vary
a given parameter. For that, we use a predictor-corrector algorithm, meaning that we produce an initial
guess at a point lying on the branch of solutions and correct it using Newton’s method until it converges
within a prescribed tolerance. In particular, we use a method called pseudo-arclength continuation. This
method considers the parameter as a variable and parameterizes the branch of solutions by pseudo-arclength
using the parameter ∆s > 0. This allows for continuation past saddle-node bifurcations. The method
outlined in this section allows one to rigorously prove the existence of a smooth solution curve between
two points lying on the curve. Furthermore, we show that the union of connecting smooth curves yields
a smooth curve and we provide a strategy for detecting secondary bifurcations. For proofs of Theorem 3,
Theorem 4, Theorem 5 and Corollary 3.1 please refer to [Breden et al., 2013].
We focus on doing continuation in the surface’s amplitude of oscillation and consider the other
parameters as constants. As previously stated in Remark 2.1, becomes a variable and the new unknown
variable which describes the periodic orbit is Φ = (ϕ, ). We redeﬁne the fmap such that f:R4k+1 R4k
where the problem becomes f(Φ) = 0. We build the predictor by computing a unit tangent vector ˙
Φ0
R4k+1 to the branch at the solution point Φ0using the fact that
DΦf0)·˙
Φ0= 0 R4k.(21)
In MATLAB, the null space is computed using singular value decomposition. The predictor is thus given
by
ˆ
Φ1:= Φ0+ ∆s˙
Φ0R4k+1.
To correct the predictor, we consider the hyperplane perpendicular to the tangent vector ˙
Φ0at the
predictor:
E(Φ) := (Φ ˆ
Φ1)·˙
Φ0
2https://github.com/ClementFortin/BilliardOscillatingParabola
Computer-assisted methods for periodic orbits in vibrating gravitational billiards 13
and we apply Newton’s method to the new zero-ﬁnding problem F:R4k+1 R4k+1 deﬁned by
F(Φ) :=E(Φ)
f(Φ)= 0,(22)
where the initial condition is given by the predictor ˆ
Φ1. Denote the new solution by Φ1where kF1)k<tol.
To prove the existence of this numerical solution, we apply Theorem 1 just as previously.
We now show that there exists a smooth curve between Φ0and Φ1parameterized by Φ = Φswith
s[0,1] such that Fs)0. We ﬁrst compute a unit tangent vector to the curve ˙
Φ1at the point Φ1by
making use of equation (21) and we denote
Φs= Φ0+s∆Φ,
˙
Φs=˙
Φ0+s˙
Φ,
where ∆Φ = Φ1Φ0and ˙
Φ = ˙
Φ1˙
Φ0. Moreover, let Es(Φ) := (Φ Φs)·˙
Φssuch that
Fs(Φ) :=Es(Φ)
f(Φ) .
The following theorem proves the existence of a solution curve between Φ0and Φ1.
Theorem 3. Consider fCk(Rn+1,Rn)with k∈ {2, ..., ∞}. For any s[0,1] ,consider the pre-
dictors Φsand the tangent vectors ˙
Φswith Fsdeﬁned as in (22). Let AMn+1(R)be such that
ADΦF0)1. Let r>0be such that Brs)Rn+1 is a cylinder. Consider nonnegative bounds
Y0,ˆ
Y0, Z0,ˆ
Z0and Z2: (0, r][0,)satisfying
kAF00)k ≤ Y0,(23)
kAFss)F00)k ≤ ˆ
Y0,for all s[0,1] (24)
Z0,for all s[0,1] (26)
kADΦFs(C)DΦFss)k ≤ Z2(r)r, for all CBrs)with r(0, r].(27)
p(r):=Z2(r)r2(1 Z0ˆ
Z0)r+Y0+ˆ
Y0.
If there exists r0(0, r]such that p(r0)<0, then there exists a function
˜
Φ: [0,1] [
s[0,1]
Br0s)
with ˜
ΦCk(0,1),Rn+1, and such that
f(˜
Φ(s)) = 0,for all s[0,1] .
In this case, the solution curve is parameterized by ˜
Φ and its existence interval is given by a cylinder. The
existence interval of this solution curve is given by I= [rmin, rmax](0, r] such that any rIyields
p(r)<0. The function p(r) is called a radii polynomial.
Remark 3.2. It is often hard to set the Z2bound of (27) such that r > 0 with p(r)<0. As shown in
[Calleja et al., 2019], a general constant bound Z2=Z2(r) satisfying (27) and obtained via the mean
value inequality is provided in the case of the -norm:
Z2(r) = sup
bBrs)
max
1in+1 X
1k,mn+1 X
1jn+1
Aij
2(Fs)j
ΦmΦk
(b)
.(28)
14 K. E. M. Church &C. Fortin
Note that we do not employ any Z1bounds in this paper, as these are only necessary when an approximate
inverse is used in the context of inﬁnite-dimensional zero-ﬁnding problems [Calleja et al., 2019; Lessard
et al., 2017].
Theorem 4. Suppose the hypotheses of Theorem 3 are satisﬁed with the norm k·k, yielding a branch of
solution curve S0:=s[0,1] ˜
Φ(s). If
∆Φ ·˙
Φ0+r0k˙
Φk+|∆Φ ·˙
Φ|<0,
then S0is a smooth curve, that is, for all s[0,1] , ds˜
Φ(s)6= 0.
Assume we have computed two smooth solution curves, S0and S1using the above results. We want
to prove that the curves connect smoothly and hence produce a single smooth curve.
Theorem 5. Let S0and S1be two smooth curves, such that the hypotheses of Theorem 3 and Theorem 4
are satisﬁed between Φ0and Φ1, and Φ1and Φ2, respectively. Then, S0∪ S1is a smooth curve.
Theorem 3, Theorem 4 and Theorem 5 allow one to prove the existence of a global smooth curve
S:=
j1
[
i=0
Si
of f(Φ) = 0 for jiterations, starting at the numerical solution Φ0which was found using Newton’s method
together with Theorem 1. As the shape of the smooth curve Sis often unknown, it is beneﬁcial to be able
to recognize any secondary bifurcations that the solution curve might be undergoing. Hence, the following
corollary provides a condition for verifying the path’s regularity.
Corollary 3.1. Assume that the hypotheses of Theorem 3 and Theorem 4 hold. Then, for every s
[0,1] ,dim Ker DΦf˜
Φ(s)= 1,that is ˜
Φ(s)is a regular path.
As is shown in Section 3.2.1, strong evidence suggest that branch solutions in the billiard always appear
from pitchfork bifurcations.
3.2.1. Results for parameters g= 9.81 and (α, e) = (2,0.778)
Applying Theorem 3 and verifying both Theorem 4 and Corollary 3.1 at every step, we do rigorous pseudo-
arclength continuation in the parameter for solutions of (α, e, ) = (2,0.778,1) and (k, p) = (2,2). We
continued from all of the (k, p) = (2,2) periodic orbits of Theorem 2 until the radii polynomial method
failed. The result is we were able to prove the following.
Theorem 6. The curves appearing in Figure 3 and Figure 6 correspond uniquely to curves of periodic
orbits of type (k, p) = (2,2) of the gravitational billiards with parameters g= 9.81 and (α, e) = (2,0.778)
for varying amplitude . These branches are located in the half-space 0.195, in the sense that at least
one periodic orbit (corresponding to a point on the branch) exists at = 0.195.
The branches seem to meet at several points in Figure 3. However, rigorously determining existence
of a smooth curve passing through either of these points proves diﬃcult as the Z2bound in (27) blows
up, even when using (28). Moreover, the kernel of the numerical approximation of DΦfat these points is
two-dimensional, and hence the curve does not deﬁne a regular path, see Corollary 3.1. By plotting the
radius of the domain of uniqueness obtained from Theorem 3 vs the amplitude of oscillation, we observe
sharp drops at particular amplitudes; see Figure 5. This is readily explained from the fact that branches
converge towards these values and thus get closer together. All of this strongly indicates the existence of
pitchfork bifurcations at the points represented by black dots in Figure 3. See Figure 4 for a more global
view of the (likely) bifurcation diagram.
In the left pane of Figure 3, two sets of horizontally opposite periodic orbit come into existence at
4.45 and at 1.545 before connecting in saddle-nodes at 0.511. In the right pane, pitchfork
Computer-assisted methods for periodic orbits in vibrating gravitational billiards 15
Fig. 3. Plots of connected branch solutions at (α, e) = (2,0.778) and (k, p) = (2,2). Left: 5 solution branches in the (t0, v0)
projection. Right: 8 diﬀerent branches in the (t1, x0) projection. Branches meet at black dots at 1.545 and 4.45 on the
top left and at 1.667 and twice at 0.195 on the top right.
0 1 2 3 4 5 6
6
7
8
9
10
11
12
13
14
15
Fig. 4. Bifurcation diagram in the the 2-norm of periodic orbit representatives ϕ(i.e. ﬁxed points of the map (10)). The pale
blue curve corresponds to the branches appearing in the left pane of Figure 3, and the black curve to the right pane. We have
windowed to solutions to 6 ≤ ||ϕ||215, since this region contains all likely pitchfork bifurcations and most of the branch
structure. Intersections of the black and blue curves are not bifurcation points, since these two curves are spatially isolated
from each other.
bifurcations are hypothesized to exist at amplitudes of 0.195 and 1.667. For (k, p) = (2,2), we
make the following distinctions between branch solutions.
(i) Child branches: branches that appear out of pitchfork bifurcations;
(ii) Parent branches: branches that beget child branches;
(iii) Wild branches: branches that are not connected to pitchfork bifurcations.
Since branches of the ﬁrst family are born through pitchforks, they always come in pairs from branches of
the second family. The yellow and red curves in the left pane of Figure 3 and the (left) dark blue, yellow,
orange and purple curves in the right pane are examples of child branches. The green (purple) and blue
curves in the left pane and the dark red, (right) dark blue, light blue and green curves in the right pane
are examples of parent branches. Using this classiﬁcation, it is possible for a branch to be both a child and
parent branch.
16 K. E. M. Church &C. Fortin
1.5 2 2.5 3 3.5
0
0.5
1
1.5
2
2.5
310-4
1234567
0.5
1
1.5
2
2.5 10-5
Fig. 5. Plots of the radius of the domain of uniqueness as a function of the amplitude of oscillation. The left plot corresponds
to the green curve in the top left pane of Figure 3, and the right plot corresponds to the red curve lying on the x0= 0 line in
the top right pane. Sharp drops in rmax are observed at 1.545 and 1.667, respectively.
Near pitchforks, the step size ∆shas to be carefully chosen so that it is not too large that it crosses
the bifurcation, in which case the Z2bound blows up, but not too small that it takes too long to advance
near the pitchfork. Therefore, it might happen that wild branches do not actually exist but are rather
a consequence of a poorly chosen step size and bifurcate oﬀ of a parent branch. In any case, potential
candidates of this branch family are shown in Figure 6.
0 2 4 6 8 10 12
12
13
14
15
16
17
18
19
20
21
Fig. 6. Wild branch candidates for (α, e) = (2,0.778) and (k, p) = (2,2). Left: 6 branches in the (v0, w0) projection. Right:
Two-norm of the solution branches of the left plot.
Remark 3.3. The code provided through Github3uses an adaptive step size, as it is dependent upon i,
the number of times the Newton operator has been applied to reach convergence in the previous proof of
existence of solution branch segment. The step size of the (n+ 1)th iteration is given by
s(n+1) = 24i
3s(n),
with an initial (and maximum) step size of ∆s(0) = 0.1. Nonetheless, one might notice that convergence is
sometimes hard to achieve even with a small step size. Indeed, there are unstable (in the sense of iteration
of the Newton operator) eight-dimensional zones near branch solutions. In some cases, a small step size is
enough to yield a diverging sequence. One can usually solve this problem by taking a larger step size.
3https://github.com/ClementFortin/BilliardOscillatingParabola
REFERENCES 17
In Figure 3, branches always stop before the surface’s amplitude becomes zero. Indeed, child branches
connect back to other child branches through saddle-nodes and parent branches stop at speciﬁc points. For
collisions that are not perfectly elastic i.e. e < 1, existence of periodic orbits is simply not possible past a
certain value of . In the particle-surface system, there is a loss in energy that the surface is no longer able
to balance out, preventing the particle from describing a periodic trajectory. The following conjecture is a
generalization of Proposition 2.
Conjecture 1. For ﬁxed numbers of impacts k, period pand parameters (g, α, e), there is a critical amplitude
ξ > 0 such that no periodic orbit exist for [0, ξ).
A consequence of Theorem 6 is that if the above conjecture is true, the critical amplitude for (g, α, e) =
(9.81,2,0.778) and (k, p) = (2,2) satisﬁes ξ < 0.195. While physically this conjecture seems perfectly
reasonable, we make no attempt to prove it here.
4. Concluding Remarks
The motion of a particle is studied in a domain bounded by a vertically-vibrating parabola. The particle is
subject to a constant gravitational ﬁeld. We build a nonlinear map in (10) for which the zeros correspond
to periodic orbits of the particle-boundary system. We then classify periodic orbits as symmetric when
the path drawn by the particle is symmetric with respect to the vertical axis, and prove some elementary
results concerning when such orbits can exist. Using a computer-assisted approach, we rigorously prove
the existence of at least 608 asymmetric or trivial (i.e. vertical motion only) periodic orbits for speciﬁc
test parameter values. We then do rigorous pseudo-arclength continuation in the boundary’s amplitude of
oscillation. Again, using computer-assisted proofs, we provide strong evidence for the existence of pitchfork
bifurcations while proving the existence of some global smooth branches of periodic orbits. Lastly, we
conjecture that no periodic orbit exist for a small enough value of the boundary’s amplitude of oscillation,
and rigorously compute an upper bound for this critical amplitude for the test parameters. It remains
to rigorously study the likely pitchfork bifurcations, characterize wild orbits, and study the behavior of
(nontrivial) symmetric periodic orbits as the vibration amplitude varies. We have proposed that periodic
orbits can not persist below a critical amplitude of surface oscillation, and have computed an upper bound
for this critical amplitude in a particular case. It would be interesting to explore this conjecture in more
detail even for the particular test parameters and (k, p) = (2,2) periodic orbits, as we have done here in a
preliminary way.
Acknowledgments
Thank you to the anonymous reviewers, whose comments led to several improvements to the paper. We
thank Olivier H´enot for providing the proof of Theorem 1. Kevin E. M. Church acknowledges the support of
NSERC (Natural Sciences and Engineering Research Council of Canada) through the NSERC Postdoctoral
Fellowships Program.
References
Baxter, A. M. & Umble, R. [2007] “Periodic orbits for billiards on an equilateral triangle,” The American
Mathematical Monthly 115, 479–491, doi:10.1080/00029890.2008.11920555.
Biswas, D. [1997] “Periodic orbits in polygonal billiards,” Pramana 48, 487–501, doi:10.1007/bf02845658.
Boshernitzan, M., Galperin, G., Krger, T. & Troubetzkoy, S. [1998] “Periodic billiard orbits are dense in
rational polygons,” Transactions of the American Mathematical Society 350, 3523–3535, doi:10.1090/
s0002-9947-98-02089-3.
Breden, M., Lessard, J.-P. & Vanicat, M. [2013] “Global bifurcation diagrams of steady states of systems of
pdes via rigorous numerics: a 3-component reaction-diﬀusion system,” Acta Applicandae Mathematicae
128, 113–152, doi:10.1007/s10440-013-9823-6, URL https://doi.org/10.1007/s10440-013-9823-6.
Calleja, R., Garca-Azpeitia, C., Lessard, J.-P. & James, J. D. M. [2019] “Torus knot choreographies in the
n-body problem,” arXiv:1901.03738 .
18 REFERENCES
Chatterjee, R., Jackson, A. D. & Balazs, N. L. [1996] “Rigid-body motion, interacting billiards, and billiards
on curved manifolds,” Physical Review E 53, 5670–5679, doi:10.1103/physreve.53.5670.
Chernov, N. & Markarian, R. [2003] Introduction to the ergodic theory of chaotic billiards (Impa).
Costa, D. R. D., Dettmann, C. P. & Leonel, E. D. [2015] “Circular, elliptic and oval billiards in a grav-
itational ﬁeld,” Communications in Nonlinear Science and Numerical Simulation 22, 731–746, doi:
10.1016/j.cnsns.2014.08.030.
Feldt, S. & Olafsen, J. S. [2005] “Inelastic gravitational billiards,” Physical Review Letters 94, 224102,
doi:10.1103/physrevlett.94.224102.
omez-Serrano, J. [2019] “Computer-assisted proofs in pde: a survey,” SeMA Journal 76, 459–484.
Hargreaves, G. [2002] “Interval analysis in matlab,” Manchester Centre for Computational Mathematics,
Numerical Analysis Reports 416.
Hartl, A. E., Miller, B. N. & Mazzoleni, A. P. [2013] “Dynamics of a dissipative, inelastic gravitational
billiard,” Phys. Rev. E 87, 032901, doi:10.1103/PhysRevE.87.032901.
Huang, J. & Fu, X. [2019] “Stability and chaos for an adjustable excited oscillator with system switch,”
Communications in Nonlinear Science and Numerical Simulation 77.
Huang, J. & Luo, A. C. J. [2017] “Complex dynamics of bouncing motions on boundaries and corners in
a discontinuous dynamical system,” Journal of Computational and Nonlinear Dynamics 12.
Korsch, H. J. & Lang, J. [1991] “A new integrable gravitational billiard,” Journal of Physics A: Mathe-
matical and General 24, 45–52, doi:10.1088/0305-4470/24/1/015.
Langer, C. K. & Miller, B. N. [2015] “A three dimensional gravitational billiard in a cone,”
arXiv:1507.06693.
Lessard, J.-P., Mireles James, J. & Hungria, A. [2016] “Rigorous numerics for analytic solutions of diﬀer-
ential equations: The radii polynomial approach,” Mathematics of Computation 85, 1427–1459.
Lessard, J.-P., Sander, E. & Wanner, T. [2017] “Rigorous continuation of bifurcation points in the diblock
copolymer equation,” Journal of Computational Dynamics 4, 71, doi:10.3934/jcd.2017003.
Luo, A. C. J. & Han, R. P. S. [1996] “The dynamics of a bouncing ball with a sinusoidally vibrating table
revisited,” Nonlinear Dynamics 10, 1–18, doi:10.1007/bf00114795.
aty´as, L. & Barna, I. [2011] “Geometrical origin of chaoticity in the bouncing ball billiard,” Chaos,
Solitons and Fractals 44, 1111–1116, doi:10.1016/j.chaos.2011.10.002.
Peraza-Mues, G. G., Carvente, O. & Moukarzel, C. F. [2017] “Rotation in a gravitational billiard,” Inter-
national Journal of Modern Physics C 28, 1750021, doi:10.1142/s0129183117500218.
Rump, S. [1999] “INTLAB - INTerval LABoratory,” Developments in Reliable Computing, ed. Csendes, T.
(Kluwer Academic Publishers, Dordrecht), pp. 77–104.
Tang, X., Fu, X. & Sun, X. [2019] “Periodic motion for an oblique impact system with single degree of
freedom,” Journal of Vibration Testing and System Dynamics 3, 71–89.
Troubetzkoy, S. [2005] “Periodic billiard orbits in right triangles,” Annales de linstitut Fourier 55, 29–46,
doi:10.5802/aif.2088.
Appendix A Partial Derivatives
For (tn, xn, vn, wn, tn+1, xn+1 , vn+1, wn+1 ) and f= [f1, f2, . . . , f4k]|, we present the entries of the Jacobian
Df of (10). The nonzero derivatives with respect to the function’s n+ 1 row i.e. Tn, are given by
∂f4n+1
∂tn
= 2αxn+v+
n(tn+1 tn) ∂v+
n
∂tn
(tn+1 tn)v+
n!2π cos(2πtn) + w+
n
∂w+
n
∂tn
+g!(tn+1 tn),
∂f4n+1
∂xn
= 2α(tn+1 tn) v+
n+xn
∂v+
n
∂xn!+ 2αv+
n
∂v+
n
∂xn
(tn+1 tn)2
∂w+
n
∂xn
(tn+1 tn),
∂f4n+1
∂vn
= 2α∂v+
n
∂vn
(tn+1 tn)[xn+v+
n(tn+1 tn)]
∂w+
n
∂vn
(tn+1 tn),
∂f4n+1
∂wn
= 2α∂v+
n
∂wn
(tn+1 tn)[xn+v+
n(tn+1 tn)]
∂w+
n
∂wn
(tn+1 tn),
∂f4n+1
∂tn+1
= 2αv+
nxn+v+
n(tn+1 tn)+ 2π cos(2πtn+1 )w+
n+g(tn+1 tn).
REFERENCES 19
The derivatives of the n+ 2, n + 3 and n+ 4 rows are
∂f4n+2
∂tn
=∂v+
n
∂tn
(tn+1 tn)v+
n,∂f4n+2
∂xn
= 1 + v+
n
∂xn
(tn+1 tn),∂f4n+2
∂vn
=∂v+
n
∂vn
(tn+1 tn),∂f4n+2
∂wn
=∂v+
n
∂wn
(tn+1 tn),
∂f4n+2
∂tn+1
=v+
n,∂f4n+2
∂xn+1
=1,∂f4n+2
∂vn+1
= 0,∂f4n+2
∂wn+1
= 0.
∂f4n+3
∂tn
=∂v+
n
∂tn
,∂f4n+3
∂xn
=∂v+
n
∂xn
,∂f4n+3
∂vn
=∂v+
n
∂vn
,∂f4n+3
∂wn
=∂v+
n
∂wn
,
∂f4n+3
∂tn+1
= 0,∂f4n+3
∂xn+1
= 0,∂f4n+3
∂vn+1
=1,∂F4n+3
∂wn+1
= 0,
∂f4n+4
∂tn
=∂w+
n
∂tn
+g, ∂f4n+4
∂xn
=∂w+
n
∂xn
,∂F4n+4
∂vn
=∂w+
n
∂vn
,∂F4n+4
∂wn
=∂w+
n
∂wn
,
∂f4n+4
∂tn+1
=g, ∂f4n+4
∂xn+1
= 0,∂f4n+4
∂vn+1
= 0,∂f4n+4
∂wn+1
=1,
where the partial derivatives of v+
nand w+
nare taken directly from the reset law,
∂v+
n
∂tn
=8απ2xnsin(2πtn)(e+ 1)
1 + (2αxn)2,∂w+
n
∂tn
=
4π2sin(2πtn)(e+ 1)
1 + (2αxn)2,
∂v+
n
∂xn
=
2α(e+ 1) 4α2wnx2
n
wn+ 2π cos(2πtn)+4αvnxn
8α2x2
nπcos(2πtn)
(1 + (2αxn)2)2,
∂w+
n
∂xn
=2α(e+ 1) vn
4α2vnx2
n+ 4αwnxn
8αxnπcos(2πtn)
(1 + (2αxn)2)2,
∂v+
n
∂vn
=142x2
n
1 + (2αxn)2,∂w+
n
∂vn
=2αxn(e+ 1)
1 + (2αxn)2,∂v+
n
∂wn
=2αxn(e+ 1)
1 + (2αxn)2,∂w+
n
∂wn
=4α2x2
n
e
1 + (2αxn)2.
The partial derivatives with respect to the parameter are given by
∂f4n+1
∂ = 2α v+
n
∂ (tn+1 tn)xn +v+
n(tn+1 tn)+ sin (2πtn)
∂w+
n
∂ (tn+1 tn),
∂f4n+2
∂ = v+
n
∂ (tn+1 tn), f4n+3
∂ = v+
n
∂ , f4n+4
∂ = w+
n
∂
and the derivatives with respect to the parameter yield
∂v+
n
∂ =2πcos (2πtn) (e+ 1) (2αxn)
1 + (2αxn)2,∂w+
n
∂ =2πcos (2πtn) (e+ 1)
1 + (2αxn)2
where n∈ {0,1, ..., k 1}.
ResearchGate has not been able to resolve any citations for this publication.
Article
Full-text available
We develop a systematic approach for proving the existence of spatial choreogra-phies in the gravitational n body problem. After changing to rotating coordinates and exploiting symmetries, the equation of a choreographic configuration is reduced to a delay differential equation (DDE) describing the position and velocity of a single body. We study periodic solutions of this DDE in a Banach space of rapidly decaying Fourier coefficients. Imposing appropriate constraint equations lets us isolate choreographies having prescribed symmetries and topological properties. Our argument is constructive and makes extensive use of the digital computer. We provide all the necessary analytic estimates as well as a working implementation for any number of bodies. We illustrate the utility of the approach by proving the existence of some spatial torus knot choreographies for n = 4, 5, 7, and 9 bodies.
Article
Full-text available
In this paper, from the local theory of flow at the corner in discontinuous dynamical systems, obtained are analytical conditions for switching impact-alike chatter at corners. The objective of this investigation is to find the dynamics mechanism of border-collision bifurcation in discontinuous dynamical systems. Multi-valued linear vector fields are employed in the discontinuous dynamical system, and generic mappings are defined among the boundaries and corners. From mapping structures, periodic motions switching on the boundaries and corners are determined, and the corresponding stability and bifurcations of periodic motions are investigated by eigenvalue analysis. However, the grazing and sliding bifurcations are determined by the local singularity theory in discontinuous dynamical systems. From such analytical conditions, the corresponding parameter map are developed for periodic motions in such multi-valued dynamical systems in the single domain with corners. Numerical simulations of periodic motions are presented for illustrations of motions complexity and catastrophe in the discontinuous dynamical system.
Article
Full-text available
We consider the problem of rigorously computing continuous branches of bifurcation points of equilibria in the one-dimensional diblock copolymer model. We apply the method both to fold points and to pitchfork bifurcations which are forced through symmetries in the equation.
Article
Full-text available
Judicious use of interval arithmetic, combined with careful pen and paper estimates, leads to effective strategies for computer assisted analysis of nonlinear operator equations. The method of radii polynomials is an efficient tool for bounding the smallest and largest neighborhoods on which a Newton-like operator associated with a nonlinear equation is a contraction mapping. The method has been used to study solutions of ordinary, partial, and delay differential equations such as equilibria, periodic orbits, solutions of initial value problems, heteroclinic and homoclinic connecting orbits in the Ck category of functions. In the present work we adapt the method of radii polynomials to the analytic category. For ease of exposition we focus on studying periodic solutions in Cartesian products of infinite sequence spaces. We derive the radii polynomials for some specific application problems, and give a number of computer assisted proofs in the analytic framework.
A linear spring-damper dynamic oscillator with excitations is studied in the paper, and a set of subsystems are defined by adjusting the constant and magnitude of the periodic external forces. A triangular domain is defined in the phase plane coordinate system, and such a dynamic system will switch to the corresponding subsystem when the flow arrives at the boundary or corner for such a domain. Through employing the theory of discontinuous, the vector fields for the subsystems have been determined which is the necessary conditions of motion ‘bouncing’ within such a triangular domain. To describe the periodic motions of such an oscillator, the generic mappings are constructed. The periodicity and stability for the motion in the steady-state have been discussed. Analytical and numerical predictions have been carried out through phase plane and switching sections to illustrate the effectiveness of the design of the subsystems under the proposed switching scheme. Periodic and chaotic motions have been simulated to institutively illustrate the system switch and stability for such a spring-damper oscillator with adjustable excitations.
Article
In this survey we present some recent results concerning computer-assisted proofs in partial differential equations, focusing in those coming from problems in incompressible fluids. Particular emphasis is put on the techniques, as opposed to the results themselves.
Article
Gravitational billiards composed of a viscoelastic frictional disk bouncing on a vibrating wedge have been studied previously, but only from the point of view of their translational behavior. In this work, the average rotational velocity of the disk is studied under various circumstances. First, an experimental realization is briefly presented, which shows sustained rotation when the wedge is tilted. Next, this phenomenon is scrutinized in close detail using a precise numerical implementation of frictional forces. We show that the bouncing disk acquires a spontaneous rotational velocity whenever the wedge angle is not bisected by the direction of gravity. Our molecular dynamics (MD) results are well reproduced by event-driven (ED) simulations. When the wedge aperture angle (Formula presented.), the average tangential velocity (Formula presented.) of the disk scales with the typical wedge vibration velocity (Formula presented.), and is in general a nonmonotonic function of the overall tilt angle (Formula presented.) of the wedge. The present work focuses on wedges with (Formula presented.), which are relevant for the problem of spontaneous rotation in vibrated disk packings. This study makes part of the PhD Thesis of G. G. Peraza-Mues.
Article
Billiard systems offer a simple setting to study regular and chaotic dynamics. Gravitational billiards are generalizations of these classical billiards which are amenable to both analytical and experimental investigations. Most previous work on gravitational billiards has been concerned with two dimensional boundaries. In particular the case of linear boundaries, also known as the wedge billiard, has been widely studied. In this work, we introduce a three dimensional version of the wedge; that is, we study the nonlinear dynamics of a billiard in a constant gravitational field colliding elastically with a linear cone of half angle $\theta$. We derive a two-dimensional Poincar\'{e} map with two parameters, the half angle of the cone and $\ell$, the $z$-component of the billiard's angular momentum. Although this map is sufficient to determine the future motion of the billiard, the three-dimensional nature of the physical trajectory means that a periodic orbit of the mapping does not always correspond to a periodic trajectory in coordinate space. We demonstrate several integrable cases of the parameter values, and analytically compute the system's fixed point, analyzing the stability of this orbit as a function of the parameters as well as its relation to the physical trajectory of the billiard. Next, we explore the phase space of the system numerically. We find that for small values of $\ell$ the conic billiard exhibits behavior characteristic of two-degree-of-freedom Hamiltonian systems with a discontinuity, and the dynamics is qualitatively similar to that of the wedge billiard, although the correspondence is not exact. As we increase $\ell$ the dynamics becomes on the whole less chaotic, and the correspondence with the wedge billiard is lost.
Article
persing billiards, Ergodic Theory & Dynam. Systems 19 (1999), 201--226. [Sz1] D. Szasz, Boltzmann&apos;s ergodic hypothesis, a conjecture for centuries ?, Studia Sci. Math. Hung. 31 (1996), 299--322. [Sz2] D. Szasz, Hard ball systems and the Lorentz gas, Edited by D. Szasz. Springer, Berlin (2000). [Ta1] S. Tabachnikov, Billiards. Panor. Synth. No. 1, SMF, Paris (1995). [Ta2] S. Tabachnikov, Exact transverse line fields and projective billiards in a ball, Geom. Funct. Anal. 7 (1997), 594--608. [Ta3] S. Tabachnikov, Introducing projective billiards, Ergodic Theory & Dynam. Systems 17 (1997), 957--976. [Va] L . N. Vaserstein, On Systems of particles with finite - range and/or repulsive interactions, Commun. Math. Phys. 69 (1979), 31--56. [Vi] M. Viana, Stochastic dynamics of deterministic systems
Article
The seminal physical model for investigating formulations of nonlinear dynamics is the billiard. This article expands on our previously published work concerning a real-world billiard. Here we provide a detailed mathematical model for describing the motion of a realistic billiard for arbitrary boundaries, where we include rotational effects and additional forms of energy dissipation. Simulations of the model are applied to parabolic, wedge, and hyperbolic billiards that are driven sinusoidally. The simulations demonstrate that the parabola has stable, periodic motion, while the wedge and hyperbola (at high driving frequencies) appear chaotic. The hyperbola, at low driving frequencies, behaves similarly to the parabola, i.e., has regular motion. Direct comparisons are made between the model's predictions and previously published experimental data. The representation of the coefficient of restitution employed in the model resulted in approximate agreement with the experimental data for all boundary shapes investigated. We show how the coefficient of restitution varies under different model assumptions. It is shown that the data can be successfully modeled with a simple set of parameters.