2.1 Full-order model formulation

To study the interaction of an elastically mounted cylinder with a fluid, we consider a variational fluid formulation based on the arbitrary Lagrangian–Eulerian (ALE) description and semi-discrete time stepping (Liu, Jaiman & Gurugubelli Reference Liu, Jaiman and Gurugubelli2014; Jaiman et al. Reference Jaiman, Sen and Gurugubelli2015). We consider a fluid domain $\unicode[STIX]{x1D6FA}^{f}(t)$ with spatial and temporal coordinates denoted by $\boldsymbol{x}^{f}$ and $t$ respectively. The Navier–Stokes (NS) equations governing an incompressible flow in the ALE reference frame are

where $\unicode[STIX]{x1D70C}^{f}$ , $\boldsymbol{u}^{f}$ , $\boldsymbol{w}$ , $\unicode[STIX]{x1D748}^{f}$ and $\boldsymbol{b}^{f}$ are the fluid density, the fluid velocity, the ALE mesh velocity, the Cauchy stress tensor and the body force per unit mass respectively. For the partial time derivative in (2.1), the ALE referential coordinate $\unicode[STIX]{x1D74C}$ is held fixed and for a Newtonian fluid $\unicode[STIX]{x1D748}^{f}$ is defined as

where $p$ , $\unicode[STIX]{x1D707}^{f}$ and $\unicode[STIX]{x1D644}$ are the pressure, the dynamic viscosity of the fluid and an identity tensor respectively. A rigid-body structure submerged in the fluid experiences unsteady fluid forces and consequently may undergo flow-induced vibrations if the body is mounted elastically. To simulate translational motion of a two-dimensional rigid body about its centre of mass, the Lagrangian motion along the Cartesian axes is given by

where $\boldsymbol{u}^{s}(t)$ represents the velocity of the immersed rigid body in the fluid domain, $\unicode[STIX]{x1D753}^{s}(t)$ denotes the position of the centre of the rigid body at time $t$ , and $\unicode[STIX]{x1D662},\unicode[STIX]{x1D658}$ , $\unicode[STIX]{x1D660}$ are mass, damping and stiffness coefficient matrices for the translational motions. Here, $\boldsymbol{F}^{s}$ is the fluid force on the rigid body.

Let $\unicode[STIX]{x1D6E4}(t)$ be the fluid–structure interface at time $t$ . The coupled system needs to satisfy the continuity of velocity and the force equilibrium at the fluid–body interface $\unicode[STIX]{x1D6E4}$ as follows:

where $\boldsymbol{n}$ is the outer normal to the fluid–body interface. In (2.7), the first term represents the force exerted by the fluid and the second term is the solid load vector applied in (2.4). The ALE mesh nodes on the fluid domain $\unicode[STIX]{x1D6FA}^{f}(\boldsymbol{x}^{f},t)$ can be updated by solving a linear steady pseudo-elastic material model,

where $\unicode[STIX]{x1D748}^{m}$ is the stress experienced by the ALE mesh due to the strain induced by the rigid-body movement, $\boldsymbol{\unicode[STIX]{x1D702}}^{f}$ represents the ALE mesh node displacement and $k_{m}$ is a mesh stiffness variable chosen as a function of the element area to limit the distortion of small elements located in the immediate vicinity of the fluid–body interface.

The weak variational form of (2.1) is discretized in space using $\mathbb{P}_{n}/\mathbb{P}_{n-1}$ iso-parametric finite elements for the fluid velocity and pressure, where $\mathbb{P}_{n}$ denotes the standard $n$ th order Lagrange finite element space on the discretized fluid domain. To satisfy the inf–sup condition, a $\mathbb{P}_{2}/\mathbb{P}_{1}$ finite element mesh is adopted and the second-order backward scheme is used for the time discretization of the NS system (Liu et al. Reference Liu, Jaiman and Gurugubelli2014). In the present study, a partitioned staggered scheme is considered for the full-order simulations of the fluid–structure interaction (Jaiman et al. Reference Jaiman, Geubelle, Loth and Jiao2011). The motion of the structure is driven by the traction forces exerted by the flowing fluid at the fluid–structure interface $\unicode[STIX]{x1D6E4}$ , whereby the structural motion predicts the new interface position and the geometry changes for the ALE fluid domain at each time step. The movement of the internal ALE fluid nodes is accommodated such that the mesh quality does not deteriorate as the motion of the solid structure becomes large. The above coupled variational formulation completes the presentation of the FOM for the direct numerical simulation of the fluid–structure interaction. We next present a state-space formulation of the model reduction using a system identification technique based on the input–output dynamics.

2.2 Basic state-space formulation and model reduction

The linear time-invariant (LTI) and multiple-input multiple-output (MIMO) model represented in a state-space form at discrete times $t=k\unicode[STIX]{x0394}t$ , $k=0,1,2,\ldots ,$ with a constant sampling time $\unicode[STIX]{x0394}t$ reads as

where $\unicode[STIX]{x1D66D}_{r}$ is an $n_{r}$ -dimensional state vector, $\unicode[STIX]{x1D66A}$ denotes a $q$ -dimensional input vector and $\unicode[STIX]{x1D66E}_{r}$ is a $p$ -dimensional output vector. The integer $k$ is a sample index for the time stepping. The system matrices are $(\unicode[STIX]{x1D63C}_{r},\unicode[STIX]{x1D63D}_{r},\unicode[STIX]{x1D63E}_{r},\unicode[STIX]{x1D63F}_{r})$ , whereby the transition matrix $\unicode[STIX]{x1D63C}_{r}$ characterizes the dynamics of the system. Here, $\unicode[STIX]{x1D63D}_{r}$ , $\unicode[STIX]{x1D63E}_{r}$ and $\unicode[STIX]{x1D63F}_{r}$ denote the input, output and feed-through matrices respectively. For the given output vector $\unicode[STIX]{x1D66E}_{r}$ , the statement of system realization is to construct the system matrices $(\unicode[STIX]{x1D63C}_{r},\unicode[STIX]{x1D63D}_{r},\unicode[STIX]{x1D63E}_{r},\unicode[STIX]{x1D63F}_{r})$ such that the vector $\unicode[STIX]{x1D66E}_{r}$ is reproduced by the state-space model. In a discrete-time setting, the state-space realization matrices $(\unicode[STIX]{x1D63C}_{r},\unicode[STIX]{x1D63D}_{r},\unicode[STIX]{x1D63E}_{r},\unicode[STIX]{x1D63F}_{r})$ of the dynamical system are constructed by the ERA, in which only the impulse response function (IRF) of the original full-order system is required for the system realization. The impulse response of the full-order linear system is first defined as $\unicode[STIX]{x1D66E}=[\unicode[STIX]{x1D66E}_{1},\unicode[STIX]{x1D66E}_{2},\unicode[STIX]{x1D66E}_{3},\ldots ,\unicode[STIX]{x1D66E}_{n_{i}}]$ , where $n_{i}$ represents the length of the impulse response and $\unicode[STIX]{x1D66E}_{i}$ denotes the IRF with the dimension $p\times q$ . Based on the impulse response, the generalized block Hankel matrices $r\times s$ can be constructed as

From the partitioned singular value decomposition of the Hankel matrix $\unicode[STIX]{x1D643}(0)$ , we have

where the diagonal matrices $\unicode[STIX]{x1D72E}$ are the Hankel singular values (HSVs) $\unicode[STIX]{x1D70E}_{i}$ , which represent the dynamical significance through sorting such that $\unicode[STIX]{x1D70E}_{1}\geqslant \cdots \geqslant \unicode[STIX]{x1D70E}_{n}\geqslant 0$ . The block matrix $\unicode[STIX]{x1D72E}_{2}$ contains the zeros or negligible elements. By truncating the dynamically less significant states, we estimate $\unicode[STIX]{x1D643}(0)\approx \unicode[STIX]{x1D650}_{1}\unicode[STIX]{x1D72E}_{1}\unicode[STIX]{x1D651}_{1}^{\ast }$ . The reduced system matrices $(\unicode[STIX]{x1D63C}_{r},\unicode[STIX]{x1D63D}_{r},\unicode[STIX]{x1D63E}_{r},\unicode[STIX]{x1D63F}_{r})$ are then defined as

Here, $\unicode[STIX]{x1D640}_{\unicode[STIX]{x1D662}}^{\ast }=[\unicode[STIX]{x1D644}_{q}\;\mathbf{0}]_{q\times N}$ and $\unicode[STIX]{x1D640}_{\unicode[STIX]{x1D669}}^{\ast }=[\unicode[STIX]{x1D644}_{p}\;\mathbf{0}]_{p\times M}$ , where $N=s\times q$ , $M=r\times p$ , and $\unicode[STIX]{x1D644}_{p}$ and $\unicode[STIX]{x1D644}_{q}$ are the identity matrices. We next present the ERA to construct the fluid ROM, which relies on the incompressible NS equations to represent the dynamics of a small-amplitude perturbation around the equilibrium base flow.

2.3 The ERA-based model reduction for VIV

In the present work, we only consider the transverse motion of a cylinder in a flowing stream for the sake of simplicity. However, the formulation of the ERA-based ROM is general for any fluid–structure system. The cylinder is mounted on a spring system in the cross-flow direction, which allows the cylinder to vibrate through unsteady lift comprising the pressure and shear stresses of the fluid. Due to the direct solution of the NS equations, the effects of added-mass and added-damping forces are implicitly captured in the present model. To perform linear stability analysis, the fluid ROM constructed by the ERA is coupled with the linear structural model. The non-dimensional structural equation for a transversely vibrating cylinder with one degree of freedom can be written as

where $Y$ is the transverse displacement, $C_{l}$ is the lift coefficient, $m^{\ast }$ and $\unicode[STIX]{x1D701}$ are the ratio of the mass of the vibrating structure to the mass of the displaced fluid and the damping coefficient respectively and $F_{s}$ is the reduced natural frequency of the structure, defined as $F_{s}=f_{N}D/U=1/U_{r}$ , where $U_{r}$ is the reduced velocity which is an alternative parameter to describe the frequency lock-in phenomenon. The mass ratio $m^{\ast }$ is a key parameter for VIV lock-in and it is defined as the ratio of the mass of the vibrating structure to the mass of the displaced fluid. The characteristic length scale factor $a_{s}$ is related to the geometry of the body. For example, the values are $a_{s}=2/\unicode[STIX]{x03C0}$ for a circular cylinder and $a_{s}=0.5$ for a square cylinder. There exists a complex dynamical relation between the transverse amplitude $Y$ and the lift force $C_{l}$ . One of the main objectives of this work is to construct a state-space relationship between the transverse force and the amplitude directly from the NS equations subject to an impulse. We next proceed to the model reduction of the fluid–structure system.

The non-dimensional structural equation (2.13) can be cast into a state-space formulation as

where the state matrices and vectors are

It is straightforward to solve (2.14) using a standard time integrator (Yao & Jaiman Reference Yao and Jaiman2016). In the present work, an exact state-space discretization of (2.14) is considered as follows:

where the state matrices are $\unicode[STIX]{x1D63C}_{\unicode[STIX]{x1D668}\unicode[STIX]{x1D659}}=\text{e}^{\unicode[STIX]{x1D63C}_{\unicode[STIX]{x1D668}}\unicode[STIX]{x0394}t}$ , $\unicode[STIX]{x1D63D}_{\unicode[STIX]{x1D668}\unicode[STIX]{x1D659}}={\unicode[STIX]{x1D63C}_{\unicode[STIX]{x1D668}}}^{-1}(\text{e}^{\unicode[STIX]{x1D63C}_{\unicode[STIX]{x1D668}}\unicode[STIX]{x0394}t}-\unicode[STIX]{x1D644})\unicode[STIX]{x1D63D}_{\unicode[STIX]{x1D668}}$ at discrete times $t=k\unicode[STIX]{x0394}t$ , $k=0,1,2,\ldots ,$ with a constant sampling time $\unicode[STIX]{x0394}t$ , $\unicode[STIX]{x1D644}$ is an identity matrix of the same size as $\unicode[STIX]{x1D63C}_{\unicode[STIX]{x1D668}}$ , and $\unicode[STIX]{x1D63E}_{\unicode[STIX]{x1D668}\unicode[STIX]{x1D659}}=[1\;0]$ . Through the input–output dynamics, the fluid ROM is derived by the ERA method as described in (2.9). The input for the ROM is the transverse displacement $Y$ , and the output is the lift coefficient $C_{l}$ . The ERA-based ROM with the single input and single output (SISO) can reformulated as

By substituting (2.17) into (2.16), the resultant ROM can be expressed as

where $(\unicode[STIX]{x1D63C}_{r},\unicode[STIX]{x1D63D}_{r},\unicode[STIX]{x1D63E}_{r},\unicode[STIX]{x1D63F}_{r})$ are the ROM matrices defined by the ERA method as given in (2.9), $\unicode[STIX]{x1D63C}_{\unicode[STIX]{x1D65B}\unicode[STIX]{x1D668}}$ denotes the coupled fluid–structure matrix in the discrete state-space form and $\unicode[STIX]{x1D66D}_{\unicode[STIX]{x1D65B}\unicode[STIX]{x1D668}}=[\unicode[STIX]{x1D66D}_{\unicode[STIX]{x1D668}}\;\;\unicode[STIX]{x1D66D}_{r}]^{\text{T}}$ .

The present ERA-based ROM reproduces the input–output dynamics of the full-order system. The linear stability analysis of the VIV system can be expressed as an eigenvalue problem of (2.18). The eigenvalue distribution of the coupled fluid–structure matrix $\unicode[STIX]{x1D63C}_{\unicode[STIX]{x1D65B}\unicode[STIX]{x1D668}}$ characterizes the stability of the VIV system. Here, ( $\unicode[STIX]{x1D706},\widehat{\unicode[STIX]{x1D66D}}$ ) correspond to continuous-time eigenvalue/eigenvectors of $\unicode[STIX]{x1D63C}_{\unicode[STIX]{x1D65B}\unicode[STIX]{x1D668}}$ , whereby the spatial structure is characterized by the complex vector field $\widehat{\unicode[STIX]{x1D66D}}$ and the temporal behaviour by the complex scalar $\unicode[STIX]{x1D706}$ . The stability analysis can be easily accomplished by tracing the trajectory of the complex eigenvalue $\unicode[STIX]{x1D706}$ in the complex plane, whereby $\widehat{\unicode[STIX]{x1D66D}}$ provides the spatial global modes of the ROM. Based on the leading global mode or least damped eigenvalue of the ERA-based ROM, we define a growth (amplification) rate $\unicode[STIX]{x1D70E}=Re(\unicode[STIX]{x1D706})$ and frequency $f=\text{Im}(\unicode[STIX]{x1D706}/2\unicode[STIX]{x03C0})$ . The construction of the above ERA-based ROM model is computationally efficient as it only relies on the impulse response of the FOM. While the aforementioned formulation is presented for transverse-only vibration of the structure, it is general for any coupled fluid–structure system. After reviewing the mathematical formulation and ERA-based ROM technique, we next present the numerical set-up and verification of our solver.

3.1 Problem definition

Figure 1 shows a schematic diagram of the set-up used in our simulation study for an elastically mounted bluff body with various cross-sections in a flowing stream. The coordinate origin is located at the geometric centre of the bluff body. The streamwise and transverse directions are denoted $x$ and $y$ respectively. A stream of incompressible fluid enters into the domain from an inlet boundary $\unicode[STIX]{x1D6E4}_{in}$ at a horizontal velocity $(u,v)=(U,0)$ , where $u$ and $v$ denote the streamwise and transverse velocities respectively. The bluff body with mass $m$ and characteristic diameter $D$ is mounted on a linear spring in the transverse direction. The damping coefficient $\unicode[STIX]{x1D701}$ is set to zero in the present work. The computational domain and the boundary conditions are also illustrated in figure 1. A no-slip wall condition is implemented on the surfaces of the bluff body and a traction-free boundary condition is implemented along the outlet $\unicode[STIX]{x1D6E4}_{out}$ , while a slip wall condition is implemented on the top $\unicode[STIX]{x1D6E4}_{top}$ and bottom $\unicode[STIX]{x1D6E4}_{bottom}$ boundaries. The numerical domain extends from $-10D$ at the inlet to $30D$ at the outlet and from $-15D$ to $15D$ in the transverse direction. Except when stated otherwise, all positions and length scales are normalized by the characteristic dimension $D$ , velocities by the free-stream velocity $U$ and frequencies by $U/D$ . The Reynolds number $Re$ of the flow is based on the characteristic dimension $D$ , the kinematic viscosity of the fluid and the free-stream speed $U$ .

Figure 1.

Schematic diagram of a representative bluff body of an elastically mounted cylinder in a uniform horizontal flow. The computational domain and boundary conditions are shown.

3.2 Linear stability analysis

We verify the validity of our ERA-based ROM by investigating the stability of the laminar wake behind a two-dimensional circular cylinder. To study the mesh convergence for our simulation study, we consider three representative discretizations $M1$ , $M2$ and $M3$ consisting of 9517, 22 004 and 41 132 $\mathbb{P}_{2}/\mathbb{P}_{1}$ iso-parametric elements. A representative M2 mesh is shown in figure 2, and the corresponding central square represents the fine mesh region around the cylinder body. The mesh in the cylinder wake is appropriately refined to resolve the alternate vortex shedding. The quality of the mesh convergence is determined by the prediction of the growth rate $\unicode[STIX]{x1D70E}$ and the frequency $f$ of the fluid ROM for the flow past a circular cylinder at $Re=60$ . The non-dimensional time-step size is $\unicode[STIX]{x0394}t=0.05$ . Based on the procedure described in the previous section, we next briefly describe the process of extracting ROMs from the incompressible NS equations.

Figure 2.

A representative finite element mesh with $\mathbb{P}_{2}/\mathbb{P}_{1}$ discretization: (a) full domain discretization and (b) close-up view of the finite element mesh in the vicinity of the cylinder. All other meshes for different bluff-body geometries are similarly created.

Figure 3.

Base flow of a stationary circular cylinder at $Re=60$ ; the streamwise velocity contours are shown. The contour levels are from $-$ 0.1 to 1.2 in increments of 0.1 and the flow is from left to right.

The ERA-based ROM is constructed in the neighbourhood of the base flow, which is computed via fixed-point iteration without the time-dependent term in (2.1). At $Re=60$ with the $M2$ mesh, figure 3 shows the streamwise velocity contours of the base flow with a symmetric recirculation bubble. The bubble length (measured from the centre of the cylinder) is $L_{w}=4.1$ , which agrees reasonably well with literature values ( $L_{w}=4.2$ , Giannetti & Luchini (Reference Giannetti and Luchini2007); $L_{w}=4.1$ , Mao & Blackburn (Reference Mao and Blackburn2014)). The Hankel matrix shown in (2.10) is obtained from the output lift signal ( $C_{l}$ ) subject to the impulse input of the transverse displacement $Y$ . A sufficiently small amplitude ( $\unicode[STIX]{x1D6FF}=10^{-4}$ ) is considered such that the flow evolves linearly for a relatively large time. To ensure that the unstable modes start to dominate the essential dynamics of the input–output relationship, an adequate number of cycles is required to capture the linear dynamics of the system. However, an excessively long simulation time should be avoided before the nonlinearity appears via exponential growth of the unstable modes.

The linearity of the unstable system is confirmed by comparing the response subject to two impulse inputs with $\unicode[STIX]{x1D6FF}=10^{-4}$ and $\unicode[STIX]{x1D6FF}=10^{-3}$ . A set of $1000$ impulse response outputs ( $C_{l}$ ) is then stacked at each time step $\unicode[STIX]{x0394}t=0.05$ , resulting in the final simulation time $tU/D=50$ . The adequate number of simulation cycles is then determined by examining the convergence of the fluid unstable eigenvalues computed from Hankel matrices with dimensions of $500\times 20$ , $500\times 50$ , $500\times 150$ and $500\times 200$ . It is found that the $500\times 150$ Hankel matrix is sufficient, which corresponds to 32.5 convective time units. The order of the ERA-based ROM is then determined by examining the singular values (HSVs) of the Hankel $500\times 150$ matrix. As shown in figure 4(a), the output lift signal $C_{l}$ gradually diverges as a function of the convective time $tU/D$ at $Re=60$ . This asymptotic divergence behaviour indicates that the wake flow begins to lose its stability via a Hopf bifurcation, which breaks the symmetry of the flow field and gives rise to a periodic vortex street. The first $30$ singular values are shown in figure 4(b). The rapidly decaying singular values indicate that a low-order ROM can be constructed. To further confirm the accuracy, the ERA-based ROM with a number of modes of $n_{r}=25$ is compared with the FOM result in figure 4(a). A good match between the impulse responses of the ROM and FOM can be seen in the figure. It is worth noting that it is not necessary for the Hankel matrix to be a square matrix for the suitability of the ERA-based ROM (Juang & Pappa Reference Juang and Pappa1985). As pointed by Juang & Pappa (Reference Juang and Pappa1985), further consideration is required to determine the optimal $r$ and $s$ in (2.10). Therefore, the Hankel matrix can be tall ( $r>s$ ), wide ( $r<s$ ) or square ( $r=s$ ). In the present study, the dimension $r$ is fixed while $s$ is tuned to obtain a reasonably converged unstable eigenvalue for a good match between the impulse response of the FOM and ROM.

Figure 4.

The ERA-based ROM for the unstable wake behind a stationary circular cylinder at $Re=60$ : (a) lift history of the FOM and the ROM based on linearized dynamics subject to the impulse response and (b) HSV distribution corresponding to a $500\times 150$ Hankel matrix.

Table 1 summarizes the comparison of the growth rate and frequency, which shows that the difference between $M2$ and $M3$ is less than $1\,\%$ . Therefore, the mesh $M2$ is adequate for the stability analysis of VIV. This study also confirms the convergence properties of our ERA-based ROM procedure for unstable wake flow. For further verification, we next show that the ERA-based ROM is able to accurately predict the onset of the unstable wake state due to a Hopf bifurcation. The onset of the unstable wake is commonly determined by the linearized NS equations in the literature (Giannetti & Luchini Reference Giannetti and Luchini2007; Marquet et al. Reference Marquet, Sipp and Jacquin2008; Mettot et al. Reference Mettot, Renac and Sipp2014). The growth rate $\unicode[STIX]{x1D70E}$ and the frequency $f$ are plotted as a function of the Reynolds number in figure 5. Instability of the base flow occurs when the growth rate crosses the $\unicode[STIX]{x1D70E}=0$ line at the critical Reynolds number $Re_{cr}\approx 46.8$ , which is in good agreement with the numerical prediction of Giannetti & Luchini (Reference Giannetti and Luchini2007) and Marquet et al. (Reference Marquet, Sipp and Jacquin2008) and the experimental measurement of Williamson (Reference Williamson1996). The frequency predicted by the ERA-based ROM at this critical Reynolds number is $f=0.119$ , which matches quite well with the results of Giannetti & Luchini (Reference Giannetti and Luchini2007) and Marquet et al. (Reference Marquet, Sipp and Jacquin2008). However, it is worth noting that the frequency given by the linear model is only valid in the vicinity of the critical Reynolds number and fails to capture the frequency in the region far away from the critical point, where nonlinear effects start to dominate the wake dynamics.

To extract the most energetic structures via the POD method, snapshots of the flow solutions are stored during the process of the ROM construction, i.e. the flow solution is recorded at each physical time step subjected to the impulse response. For the unstable wake case at $Re_{cr}\approx 46.8$ , the first POD mode corresponding to the streamwise and cross-flow velocity is shown in figure 6. The spatial structures are antisymmetric and they travel downstream with the formation of Kelvin–Helmholtz instabilities.

Figure 5.

Prediction of the critical Reynolds number via the ERA-based ROM for the flow past a stationary circular cylinder: (a) growth rate $\unicode[STIX]{x1D70E}$ and (b) frequency $f$ . The cylinder wake becomes unstable when the growth rate crosses the $\unicode[STIX]{x1D70E}=0$ line at the critical $Re_{cr}\approx 46.8$ and vortex shedding emanates.

Figure 6.

The first POD mode at $Re_{cr}\approx 46.8$ : (a) streamwise velocity and (b) cross-stream velocity. The flow is from left to right.

Table 1.

Mesh convergence study: comparison of growth rate and frequency for meshes $M1$ , $M2$ and $M3$ for the flow past a stationary circular cylinder at $Re=60$ .

3.3 Unstable flow past a stationary cylinder

As discussed in § 3.2, the wake flow becomes unstable through a Hopf bifurcation when $Re>Re_{cr}$ and vortex shedding appears behind a stationary cylinder at the frequency $f_{vs}$ . The unstable flow finally reaches a fully nonlinear state with the alternate time-periodic vortex shedding. The flow field in the whole domain behaves like a global oscillator, which causes unsteady lift and drag forces on the immersed body. To further establish the validity of the numerical method and the desired accuracy for the VIV simulation, the dimensionless vortex shedding frequency $St=f_{vs}D/U$ and the root mean square (r.m.s.) value of the lift coefficient $C_{l}$ are compared with the works of Williamson (Reference Williamson1989) and Zhang et al. (Reference Zhang, Fey, Noack, Knig and Eckelmann1995) for the Reynolds number $Re<180$ . The results are summarized in table 2, which demonstrates good agreement with the literature. This further confirms that the numerical methodology and the mesh discretization are adequate to capture the vortex dynamics and the stability characteristics of the VIV response.

Table 2.

Comparison of the r.m.s. value of the lift coefficient ( $C_{l}$ ) and the Strouhal number $(St)$ obtained with previous studies for a range of Reynolds numbers. A constant time-step size $\unicode[STIX]{x0394}t=0.05$ is employed for the present computations.

4.1 Assessment of the ERA-based ROM

In this section, the constructed ERA-based ROM is first applied to analyse the stability properties of a transversely vibrating circular cylinder at $(Re,m^{\ast })=(60,10)$ . Consistent with the previous literature of Meliga & Chomaz (Reference Meliga and Chomaz2011) and Zhang et al. (Reference Zhang, Li, Ye and Jiang2015), we use the terminology of SM and WM to classify the distinct eigenvalue trajectories of the linear fluid–structure system governed by (2.18). When the eigenvalue of the fluid–structure system approaches that of the stationary cylinder, the resulting mode is defined as a WM. The fluid–structure mode is considered as an SM if the eigenvalue comes close to the natural frequency of the cylinder-alone system.

As discussed in Zhang et al. (Reference Zhang, Li, Ye and Jiang2015), VIV lock-in may result either from instability of the WM alone or via simultaneous existence of unstable SM and WM. In the event of the first scenario, the lock-in occurs due to the closeness of the frequencies of the WM and SM. This type of VIV branch is termed as resonance-induced lock-in. For the second scenario, the instability to sustain the VIV lock-in occurs via combined mode instability of the SM and WM, which is referred to as flutter-induced lock-in or coupled-mode flutter (De Langre Reference De Langre2006). In the present study, the WM is also considered as the leading mode of the unstable fluid system.

We consider the continuous-time eigenvalues in the context of linear stability analysis via the ERA-based ROM. Using the fluid–structure matrix (2.18), the eigenvalue is obtained from $\unicode[STIX]{x1D706}=\text{log}(\text{eig}(\unicode[STIX]{x1D63C}_{\unicode[STIX]{x1D65B}\unicode[STIX]{x1D668}}))/\unicode[STIX]{x0394}t$ , where $\unicode[STIX]{x0394}t=0.05$ is the time step. To graphically depict the dynamical behaviour, we study the eigenvalue distribution of the system in the complex plane via the root locus. The positions of the eigenvalues provide the information about the stability of the fluid–structure system. For example, the roots in the right half-plane depict the unstable modes of the system, whereas the roots on the real axis characterize the asymptotic (non-oscillatory) behaviour. Roots that are closest to the right half-plane are lightly damped oscillatory modes.

Figure 7.

Eigenspectrum of the ERA-based ROM for a circular cylinder at $(Re,m^{\ast })=(60,10)$ : (a) root loci as a function of the reduced natural frequency $F_{s}$ , where the unstable right half-plane $(\text{Re}(\unicode[STIX]{x1D706})>0)$ is shaded in grey colour and the hollow arrow indicates increasing $F_{s}$ ; (b) real and imaginary parts of the root loci, where the lock-in region is shaded in grey colour. Two branches of lock-in, namely resonance and flutter, can be seen in (b).

Figure 7(a) shows the eigenvalue trajectory of the fluid–structure system as a function of the reduced natural frequency $F_{s}$ with $0.05\leqslant F_{s}\leqslant 0.25$ and the increment is $\unicode[STIX]{x0394}F_{s}=0.025$ . In the figure, the WM branch originates in the vicinity of the eigenvalue of the stationary cylinder (uncoupled WM) and loops back as the reduced natural frequency $F_{s}$ increases. It is expected that the WM will finally recover to the eigenvalue of the stationary cylinder as $F_{s}$ approaches infinity or the cylinder becomes stationary (i.e. without elastic mounting). It is also evident that the WM remains unstable ( $\unicode[STIX]{x1D70E}>0$ ) throughout the sweeping as $Re=60>Re_{cr}$ . On the other hand, the SM branch rises from the bottom of the complex plane (low-frequency regime) to the upper complex plane (high-frequency regime). As elucidated in figure 7(b), the SM becomes unstable only when $0.147<F_{s}\leqslant 0.179$ , which is determined by the real part of the eigenvalue. As mentioned earlier, the unstable SM phenomenon can be considered as coupled-mode flutter or combined mode instability. As shown in figure 7(b), the imaginary part of the eigenvalue as a function of $F_{s}$ reveals that two distinct frequencies (WM and SM frequencies) co-exist in the combined mode instability. In addition, figure 7(b) also indicates that the frequencies of the WM and SM come closer when $0.11\leqslant F_{s}\leqslant 0.147$ , which is recognized as the resonance mode. It should be noted that the lower left boundary of the resonance mode cannot be pinpointed precisely from the ROM due to the overlapping of the SM and WM trajectories. Thus, we determine the frequency at the lower left boundary from the FOM simulation, which is found to be $F_{s}=0.11$ .

To further verify the stability results predicted by the ERA-based ROM, the VIV response is computed by direct numerical simulation using the FOM. As shown in figure 8(a), the vortex shedding frequency begins to synchronize with the natural frequency of the structure at $F_{s}=0.11$ and recovers to the vortex shedding frequency at $F_{s}=0.179$ . As the nonlinearity starts to dominate the VIV response, the two distinct frequencies of the WM and SM corresponding to the flutter mode are eventually synchronized with the natural frequency of the structure $F_{s}$ . Figure 8(b) suggests that the cylinder rises to the peak VIV amplitude at $F_{s}=0.179$ (lock-in onset $U_{r}\approx 5.59$ ), which compares accurately with the upper boundary of the flutter mode predicted by the present ERA-based ROM.

It is worth mentioning that the cylinder acquires the maximum amplitude at $F_{s}=0.179$ or $F_{s}/St=1.31$ , which is not at $F_{s}/St\approx 1$ , as expected from the classical resonance interpretation of VIV lock-in. This phenomenon suggests that the onset of VIV lock-in results from the amplification of energy input as a consequence of an unstable SM, in which the structure is able to optimally absorb energy from the surrounding fluid system. This is analogous to the pitch and plunge flutter observed in the aeroelastic airfoil configuration. The flutter mode of VIV lock-in results from the coupling of periodic vortex shedding and the structural transverse displacement. In the flutter regime ( $1.07<F_{s}/St\leqslant 1.31$ ), the unstable SM and WM jointly sustain VIV lock-in, whereas the WM dominates the resonance regime ( $0.8\leqslant F_{s}/St\leqslant 1.07$ ) until the VIV goes into the lock-out region.

Figure 8.

The VIV results as a function of the reduced natural frequency $F_{s}$ using the FOM at $(Re,m^{\ast })=(60,10)$ : variation of (a) the normalized vortex shedding frequency $f$ and (b) the r.m.s. value of the lift coefficient ( $C_{l}$ ) and the normalized maximum amplitude ( $Y_{max}$ ). The lock-in is shaded in grey colour.

More quantitative insight into the VIV lock-in mechanism can be obtained from figure 9(a), which shows the phase angle $\unicode[STIX]{x1D719}$ estimated by the ERA-based ROM (see appendix A). The phase angle $\unicode[STIX]{x1D719}$ of the ROM is a function of ( $F_{s},\unicode[STIX]{x1D706}$ ) and its sign is determined by the real part of the eigenvalues. The instantaneous phase angle of the FOM is obtained by the Hilbert transformation of lift and displacement, as described in Tham et al. (Reference Tham, Gurugubelli, Li and Jaiman2015). In figure 9(a), we present the phase angles of the FOM and ROM as functions of $F_{s}$ at $(Re,m^{\ast })=(60,10)$ . Compared with the FOM result, the WM trajectory yields a continuous transition from $0^{\circ }$ to $180^{\circ }$ as $F_{s}$ decreases in the lock-in region. In contrast, the phase angle of the SM remains positive only within the flutter mode ( $0.147<F_{s}\leqslant 0.179$ ).

Figure 9.

The VIV results as a function of the reduced natural frequency $F_{s}$ at $(Re,m^{\ast })=(60,10)$ : (a) comparison of phase angle difference $\unicode[STIX]{x1D719}$ between the ERA-based ROM and the FOM, where the lock-in is shaded in grey colour; (b) lift $C_{l}$ history at two representative frequencies $F_{s}=(0.14,0.177)$ . In (a), (– - – - –) represents $F_{s}=0.13$ .

It is also interesting to note from figure 8(b) that the lift coefficient is significantly amplified in the vicinity of the VIV lock-in onset reduced velocity ( $U_{r}\approx 5.59$ ). A gradual decrease and eventually recovery to the value of the stationary cylinder counterpart as $U_{r}$ increases ( $F_{s}$ decreases) can be seen in the figure. To further assess the behaviour of the lift coefficient in the flutter and resonance regimes, the lift histories for two representative reduced frequencies $F_{s}=0.177$ and 0.140 are compared in figure 9(b). The minimum r.m.s. value of the lift coefficient $C_{l}$ occurs at $F_{s}\approx 0.13$ or $F_{s}/St\approx 0.95$ , which coincides with the phase angle jump, as shown in figure 9(a). Therefore, we can infer that the reduction in the r.m.s. lift coefficient has a direct relation with the phase angle. In § 4.2, we further confirm that the lift reduction is not associated with the resonance mode.

Geometric and physical parameters such as the cross-sectional shape and mass ratio play an important role with regard to the coupling strength of the fluid–structure interaction. A classification of the fluid–structure eigenmodes as WM and SM is suitable for weak fluid–structure interaction (e.g. very large mass ratio), which is elucidated in figure 7 by two distinct eigenvalue branches of the WM and SM. Owing to weaker fluid–structure coupling at large $m^{\ast }$ (i.e. in the limit of $m^{\ast }\rightarrow \infty$ ), the eigenfrequency of the WM approaches that of the stationary cylinder wake for all values of $F_{s}$ and the frequency of the SM comes close to the natural frequency of the cylinder-only system. However, the root loci of the WM and SM can coalesce and form coupled modes in certain conditions, such as in the limit of low mass ratio (Meliga & Chomaz Reference Meliga and Chomaz2011; Navrose & Mittal Reference Mittal2016) and for different geometrical shapes, as demonstrated in § 4.4. These coupled modes do not offer distinct characteristics of the WM and SM, since the two branches exchange their roles when coalescence of eigenvalue occurs. Similarly to Navrose & Mittal (Reference Mittal2016), we term these mixed or coupled modes as wake–structure mode I (WSMI) and wake–structure mode II (WSMII). For higher values of $F_{s}$ , WSMI behaves as the WM and WSMII recovers to the SM. On the other hand, for a smaller $F_{s}$ range, WSMI and WSMII represent the SM and WM respectively. A characteristic anticrossing with a frequency splitting can be also observed at low mass ratios, which is one of the traits of strongly coupled harmonic oscillators (Novotny Reference Novotny2010). To demonstrate the effect of the mass ratio, further details on the coupled modes WSMI and WSMII are discussed in appendix B.

Figure 10.

The effect of the Reynolds number on the eigenspectrum of the ERA-based ROM at $Re=(60,70,100)$ and $m^{\ast }=10$ : (a) root loci as a function of the reduced natural frequency $F_{s}$ , where the unstable right half-plane $(\text{Re}(\unicode[STIX]{x1D706})>0)$ is shaded in grey colour and the hollow arrow indicates increasing $F_{s}$ ; (b) the real and imaginary parts of the root loci. The SM data points are denoted by the filled symbols with the same shape as the WM points in (a,b). The boundary of $\text{Re}(\unicode[STIX]{x1D706})>0$ for SM at $Re=70$ is defined at $F_{s}=0.106$ (– – –) and $F_{s}=0.187$ (– - – - –) in the real parts of the root loci in (b).

In this section, the ERA-based ROM is successfully employed to perform linear stability analysis of the VIV of a transversely vibrating circular cylinder. Consistent with the previous study of Zhang et al. (Reference Zhang, Li, Ye and Jiang2015) on the mechanism of VIV, we clearly observe two distinct lock-in patterns of the flutter and the resonance from our eigenmode analysis. However, the regime classification of Zhang et al. (Reference Zhang, Li, Ye and Jiang2015) was only based on the VIV linear analysis at $Re=60$ . In the next section, we extend the existence and dependence of the two distinct lock-in modes to a larger parameter space of Reynolds number in the laminar flow regime.

4.2 Effect of Reynolds number

As shown in figure 5(a), the growth rate is amplified as $Re$ increases, which indicates that the coupling between the WM and the SM tends to become stronger for higher $Re$ . To further investigate the effect of Reynolds number, the VIV ROMs for $Re=(70,100)$ and $m^{\ast }=10$ are constructed and a stability analysis similar to that for $Re=60$ is carried out. The root loci as a function of the natural frequency $F_{s}$ are shown in figure 10(b). Compared with the root loci at $Re=60$ , figure 10(b) shows that the range of unstable SM or flutter mode increases slightly to $0.106\leqslant F_{s}\leqslant 0.187$ or $0.71\leqslant F_{s}/St\leqslant 1.25$ and covers the entire lock-in region. This is also evident from the FOM results, as shown in figure 11, where the lock-in region is $0.11\leqslant F_{s}\leqslant 0.192$ or $0.73\leqslant F_{s}/St\leqslant 0.128$ . This new finding from the present work suggests that the extents of the flutter and resonance modes are highly sensitive to the Reynolds number. This can be further elucidated by looking into the stability region where the magnitude of the velocity leading eigenmode is large (Giannetti & Luchini Reference Giannetti and Luchini2007). The complex modal velocity components, the streamwise velocity $\widehat{u}$ and the transverse velocity $\widehat{v}$ , are computed from the linearized NS equations around the base flow. To visualize the magnitude of the leading modal velocity, we first compute the amplitudes of the complex modal velocity components ( $|\widehat{u}|$ and $|\widehat{v}|$ ) and then evaluate the pointwise modal velocity magnitude as $|\widehat{U}|=\sqrt{|\widehat{u}|^{2}+|\widehat{v}|^{2}}$ . As shown in figure 12, the stability region shifts gradually to the bluff body, which indicates that the coupling effect between the unstable wake and the bluff body is enhanced as $Re$ increases.

Figure 11.

The VIV results as a function of the reduced natural frequency $F_{s}$ from the FOM at $(Re,m^{\ast })=(70,10)$ : variation of (a) the normalized vortex shedding frequency $f$ , (b) the r.m.s. value of the lift coefficient ( $C_{l}$ ) and the maximum amplitude ( $Y_{max}$ ), and (c) the phase angle ( $\unicode[STIX]{x1D719}$ ). The lock-in is shaded in grey colour.

Figure 10 shows that the unstable SM branch becomes gradually more pronounced and covers the lock-in region as the Reynolds number increases, whereas the size of the WM loop becomes smaller. The threshold Reynolds number is approximately $Re=70$ when the resonance mode ceases to exist for $m^{\ast }=10$ . This result suggests that the frequency lock-in VIV is pure flutter mode instability for $Re\geqslant 70$ , which is consistent with the theoretical analysis of De Langre (Reference De Langre2006) using the wake-oscillator model. However, due to the simplification in the wake-oscillator model and without nonlinear and dissipative terms, a general statement on VIV lock-in as a coupled flutter mode may not be valid for all flow conditions. On the other hand, the numerical study of Zhang et al. (Reference Zhang, Li, Ye and Jiang2015) is only valid for $Re_{cr}<Re<70$ . Table 3 summarizes the existence of flutter and resonance modes at different Reynolds numbers for a vibrating circular cylinder.

Figure 12.

The influence of the Reynolds number on the stability regions defined by the pointwise modal velocity magnitude $|\widehat{U}|$ of the leading eigenmodes for $Re=$ (a) 60 and (b) 70. The flow is from left to right.

Figure 13.

Full-order results for a circular cylinder at $(Re,m^{\ast })=(70,10)$ . Instantaneous vorticity contours at $F_{s}=$ (a) 0.13, (b) 0.15, (c) 0.17 and (d) 0.19. The contour levels are from $-$ 0.5 to 0.5 in increments of 0.077 and the flow is from left to right.

Table 3.

The VIV lock-in regimes of a circular cylinder for Reynolds number in the range $30\leqslant Re\leqslant 100$ and mass ratio $m^{\ast }=10$ . For $Re>Re_{cr}$ , the flutter regime comprises both unstable eigenvalues $(Re(\unicode[STIX]{x1D706})>0)$ of the WM and SM, whereas the resonance regime has only unstable WM.

It is also interesting to note from figure 11(b) that a reduction in the r.m.s. value of the lift coefficient also appears, although the resonance regime does not exist at $Re=70$ . This observation suggests that the reduction of the r.m.s. lift force does not interlink with the flutter or resonance regime, which is different from the conclusion reached by Zhang et al. (Reference Zhang, Li, Ye and Jiang2015) that the r.m.s. lift coefficient is attenuated in the resonance regime but amplified in the flutter region. As shown in figure 11(b,c), the minimum lift coefficient appears at $F_{s}\approx 0.135$ or $F_{s}/St\approx 0.92$ , where the phase angle changes from $0^{\circ }$ to $180^{\circ }$ . This further confirms that the lift reduction phenomenon is explicitly linked with the phase angle.

Furthermore, figure 13 shows the instantaneous patterns of vortex shedding investigated for a broad range of reduced natural frequencies. We adopt the classical terminology of Williamson & Roshko (Reference Williamson and Roshko1988) to identify the vortex shedding patterns. In the 2S mode, a single vortex is alternately shed from each side of the cylinder per shedding cycle, whereas the vortices start to coalesce in the far wake in the C(2S) mode. As reported by Zhang et al. (Reference Zhang, Li, Ye and Jiang2015), the 2S mode is observed in the resonance regime, whereas the C(2S) mode appears in the flutter region. However, we argue that the vorticity topology changes gradually from the C(2S) to the 2S mode as $F_{s}$ decreases from 0.19 to 0.13, which indicates that the topology variation is associated with the vibration amplitude. The C(2S) mode starts to appear at VIV lock-in onset $U_{r}\approx 5.21$ ( $F_{s}=0.192$ ) and gradually transits to the 2S mode as the amplitude decreases. To further generalize our ERA-based ROM for the VIV lock-in regime, we next investigate the influence of rounding in the VIV lock-in mechanism of a square cylinder.

4.3 Effect of rounding

In the previous section, the effects of Reynolds number and mass ratio have been considered for the transverse VIV of a circular cylinder. It is interesting to explore whether the aforementioned VIV lock-in regimes of a circular cylinder still apply to an elastically mounted square cylinder. It is known that the presence of sharp corners on a square cylinder vastly alters the flow dynamics compared with the dynamics for cylinders with circular/elliptical sections having smooth curves. Besides the angle of incidence, the sharp corners are important contributing factors in the geometry of bluff body, as they affect the flow separation points, which in turn dictate the wake dynamics. By gradual removal of the sharp corners of a square cylinder, a circular cross-section can be recovered. As reported in Jaiman et al. (Reference Jaiman, Sen and Gurugubelli2015), the VIV response of a square cylinder is dramatically different in comparison with its circular cylinder counterpart. For example, the lock-in region of a square cylinder is narrower and the amplitude is smaller for similar VIV operational parameters ( $Re,m^{\ast },\unicode[STIX]{x1D701}$ ), as extensively discussed in Jaiman et al. (Reference Jaiman, Guan and Miyanawala2016a ,Reference Jaiman, Pillalamarri and Guan b ). Recently, a rounded square was also numerically studied in terms of primary and secondary instabilities (Park & Yang Reference Park and Yang2016); it was shown that sharp corners alter the flow topology significantly, subsequently changing the stability properties of the wake dynamics. It is therefore important to consider the effect of rounding the corners of a square cylinder for the VIV mechanism. The VIV stability properties of five different cross-sections including a circle and a square are explored in the context of eigenvalue distributions.

Figure 14.

Square-section bluff body with projected width $D$ and rounding radius $r_{s}$ in uniform flow. Rounding is introduced by inscribing a quarter circle with $r_{s}$ at each edge of the square geometry. The square and circular cylinders correspond to rounding radii of $r_{s}=0$ and $r_{s}=0.5D$ respectively.

Figure 14 schematically depicts a square cylinder with a rounding radius $r_{s}$ . The edge length of the square with sharp corners is denoted by $D$ . Rounding is introduced by inscribing a quarter circle with $r_{s}$ at each edge of the square geometry, where $r_{s}=0.5D$ corresponds to a circular shape and $r_{s}=0$ recovers to the basic square shape. The characteristic dimension $D$ of all of the cross-sections is identical, where $D$ is the maximum dimension of the cylinder normal to the flow. The origin of the Cartesian coordinate system coincides with the centre of the square.

Figure 15.

The effect of the rounding $r_{s}$ on the eigenspectrum of the ERA-based ROM at $(Re,m^{\ast })=(60,10)$ : (a) root loci as a function of the reduced natural frequency $F_{s}$ , where the unstable right half-plane $(\text{Re}(\unicode[STIX]{x1D706})>0)$ is shaded in grey and the hollow arrow indicates increasing $F_{s}$ , and (b) the real and imaginary parts of the root loci. In (a), the unstable eigenvalues of the stationary square cylinder (uncoupled WM) with different rounding values are connected by a black curve with a solid arrow indicating increasing $r_{s}$ . The SM is denoted by filled symbols with the same shape as those for the WM in (a,b).

Figure 16.

The VIV results for a square cylinder with sharp corners using the FOM at $(Re,m^{\ast })=(60,10)$ : variation of (a) the normalized vortex shedding frequency $f$ , (b) the r.m.s. value of the lift coefficient ( $C_{l}$ ) and the maximum amplitude ( $Y_{max}$ ), and (c) the phase angle ( $\unicode[STIX]{x1D719}$ ). The lock-in is shaded in grey colour.

To begin with, the VIV linear analysis is conducted for a square cylinder with sharp corners at ( $Re,m^{\ast }$ ) = (60, 10). It is evident from figure 15(a) that the SM is stable, which suggests that the lock-in is entirely dominated by the resonance mode. Due to the absence of lock-in via the flutter mode, the onset reduced velocity ( $U_{r}$ ) for a square cylinder may not be clearly recognized compared with its circular cylinder counterpart. Furthermore, the region of the WM loop for the square cylinder is smaller than its circular cylinder counterpart, implying that the coupling between the fluid and the structure is reduced due to the sharp corners. In figure 15, the root loci for the fluid–structure system of a square cylinder provide an explanation that the lock-in only consists of the resonance mode and the flutter state disappears due to the presence of sharp corners for $(Re,m^{\ast })=(60,10)$ . As expected, the sharp corners suppress the continuous movement of separation points, whereby the smooth circular cylinder promotes the movement of separation points. Figure 16 shows the frequency, the VIV response and the lift coefficient from the FOM simulation for a vibrating square cylinder. The extent of the lock-in region can be observed as $0.11\leqslant F_{s}\leqslant 0.154$ or $0.87\leqslant F_{s}/St\leqslant 1.21$ . The results illustrate that the maximum amplitude is also acquired approximately at the lock-in onset ( $U_{r}\approx 6.49$ ) even if no flutter regime exists. It is notable that the maximum amplitude is smaller and the lock-in region is narrower than its circular cylinder counterpart, which is observed earlier for the square cylinder (Zhao, Cheng & Zhou Reference Zhao, Cheng and Zhou2013; Jaiman et al. Reference Jaiman, Pillalamarri and Guan2016b ).

Similarly to its circular counterpart, the lift coefficient for a square cylinder also experiences an amplification in the vicinity of lock-in onset, $U_{r}\approx 6.49$ , and gradually recovers to its stationary counterpart as $F_{s}$ decreases ( $U_{r}$ increases). The maximum value of the r.m.s. lift coefficient is $C_{l}=0.314$ , as shown in figure 16(b), which is approximately 3.1 times larger than the stationary square cylinder value ( $C_{l}=0.1$ ). For the VIV of the circular cylinder, on the other hand, the amplification of the r.m.s. lift coefficient is approximately $5.9$ times that of the stationary circular cylinder, as shown in figure 8 b. To understand the direction of energy transfer between the fluid and the square cylinder, the phase angle is shown in figure 16(c). Similarly to the circular cylinder, there is a sudden jump from $0^{\circ }$ to $180^{\circ }$ during the lock-in region around $F_{s}\approx 0.125$ or $F_{s}/St\approx 0.98$ , where the r.m.s. lift coefficient acquires the minimum value.

Figure 17.

Full-order results for the square cylinder at $(Re,m^{\ast })=(60,10)$ . Instantaneous vorticity contours at $F_{s}=$  (a) 0.12, (b) 0.13, (c) 0.14 and (d) 0.15. The contour levels are from $-$ 0.5 to 0.5 in increments of 0.077 and the flow is from left to right.

For the range of $F_{s}$ considered, two counter-rotating vortices (2S mode) are released every oscillation cycle from the rear of each cylinder, as shown in figure 17. From the vorticity fields, separations along the front corners of lateral edges for the square with sharp corners can be observed for the considered cases. While the C(2S) mode is observed in the vicinity of lock-in onset for the circular cylinder, the classic 2S mode is dominant for the square cylinder. This is consistent with our previous hypothesis that the wake pattern is closely related to the vibration amplitude. We next elucidate the influence of rounding on the distribution of the eigenspectrum and the unstable global modes.

Figure 18.

Stability regions shown by the spatial distribution of the pointwise modal velocity amplitude $|\widehat{U}|$ at $Re=60$ for different rounding parameters $r_{s}=$ (a) 0.0, (b) $0.1D$ , (c) $0.2D$ and (d) $0.4D$ . The flow is from left to right.

As shown in figure 15(a), the SM trajectory moves gradually towards the positive real axis $(Re(\unicode[STIX]{x1D706})>0)$ as the rounding radius $r_{s}$ increases. Meanwhile, the flutter region starts to appear and achieves its maximum extent for the rounding radius $r_{s}=0.5D$ . As expected, the rounding of the corners delays the onset of separation; hence, the rounding aids in reducing the bluffness of the square cylinder. It is also worth noting that the unstable SM and WM loops are both more pronounced, indicating that the coupling effect between the fluid and the structure becomes stronger. However, the growth rate of the uncoupled WM does not increase monotonically from the square ( $r_{s}=0$ ) to the circular ( $r_{s}=0.5D$ ) configuration, as shown by a black curve with a solid arrow in figure 15(a). The growth rate first decreases, then increases and eventually recovers to the growth rate of the circular cylinder.

By examining the real parts of the root loci in figure 15(b), it can be seen that the onset of lock-in starts to move towards the low reduced velocity (high $F_{s}$ ) as the rounding radius $r_{s}$ increases. In figure 18, from the comparison of stability regions through the contours of $|\widehat{U}|$ , we observe that the rounding has significant effects on the wake topology, subsequently altering the stability properties. The separation point can move widely as the rounding radius increases, indicating that the flutter mode is more pronounced. It is also shown that the stability region gradually moves closer to the rear of the bluff body as $r_{s}$ increases, which is similar to the effect of $Re$ , as discussed in § 4.2. Consequently, the level of fluid and structure interaction is enhanced. This trend is monotonic compared with the growth rate, suggesting that the flutter mode becomes gradually more pronounced. This explains why the SM trajectory moves towards the right half-plane monotonically. The WM branch, on the other hand, moves towards the imaginary axis positive or higher-frequency direction, monotonically. For similar VIV operating parameters, the circular cylinder is much easier to perturb compared with its square counterpart at the same Reynolds number.

Figure 19.

Schematics of bluff-body geometries with relevant dimensions. The representative elliptical cylinder (a) has aspect ratio $AR=0.5$ ; the forward triangle (b) is equilateral with angle $60^{\circ }$ ; the diamond (c) represents a square cylinder with sharp corners at $45^{\circ }$ flow incidence.

While the rounding generally stabilizes the wake flow for a stationary square cylinder, it promotes the movement of separation points along the smooth rounded surface of the vibrating cylinder. Analogously to the aeroelastic flutter with plunge-torsion mode coupling for an airfoil configuration, as described by De Langre (Reference De Langre2006), the transverse periodic displacement and the movement of separation points can form a similar dynamics for a transversely vibrating circular cylinder. The square cylinder with sharp corners restricts the free motion of separation points and is relatively stable in the sense that the lock-in onset reduced velocity $U_{r}$ is greater than its circular counterpart. Moreover, compared with the circular cylinder at $(Re,m^{\ast })=(60,10)$ , the lock-in range of the square cylinder is narrower and only a resonance regime exists. To further generalize our findings, we next examine the lock-in regimes from the eigenvalue distributions for additional bluff bodies for the smooth curve geometry of an elliptic cylinder and two sharp corner shapes of forward triangle and diamond cylinders.

4.4 Effect of geometry

In this section, a set of three representative two-dimensional geometries is assessed to elucidate the frequency lock-in regimes and to demonstrate the ability of the developed ERA-based ROM. The three additional geometries, namely an ellipse as a smooth curve and a forward triangle and a diamond with sharp corners, are shown in figure 19. The major axis with length $D$ of the elliptic cylinder is placed normal to the flow direction and the aspect ratio $AR=0.5$ is defined by the ratio of the minor, $D/2$ , to the major axis length. The forward triangle is equilateral and has an edge with length $D$ normal to the flow; the peak corner is on the leeward side. Similarly to Zhao et al. (Reference Zhao, Cheng and Zhou2013), the diamond geometry is considered as a square cylinder with $45^{\circ }$ flow incidence and the Reynolds number is defined by the edge length $D$ as the characteristic length scale. Similarly to the circular and square cylinders, the new geometries of the ellipse, forward triangle and diamond cylinders undergo unsteady wake transition via a Hopf bifurcation at a critical Reynolds number $Re_{cr}$ , which is responsible for the onset of the time-periodic vortex shedding phenomenon. Table 4 shows the critical Reynolds numbers $Re_{cr}$ for the different geometries computed by our ERA-based ROM, which match reasonably well with previous studies. It is worth noting that the forward triangle has the lowest critical Reynolds number for the initial wake transition from steady to unsteady flow. The unsteady transitions of the ellipse with aspect ratio $AR=0.5$ and the diamond with sharp corners occur lower than their circular and square counterparts. Due to unsteady lift and drag forces, the three geometries can undergo flow-induced vibration if mounted on elastic supports.

Table 4.

Comparison of critical Reynolds numbers $Re_{cr}$ between the available literature and the predicted values by the ERA-based ROM for different topologies of bluff bodies. The onset reduced velocity $U_{r}$ of VIV lock-in from the present study is also outlined in the final row.

To further elucidate the lock-in mechanism, we plot the root loci and the real and imaginary parts of the eigenvalues for the additional three bluff bodies in figure 20. The figure clearly shows that the geometry of the bluff body has a significant impact on the eigenvalue trajectory. The elliptical configuration has the lowest lock-in onset $U_{r}$ followed by the diamond and forward triangle configurations, as shown in table 4. Compared with the circular cylinder at identical conditions of $(Re,m^{\ast })=(60,10)$ , the root loci of the SM and WM coalesce for the diamond, ellipse and forward triangle configurations. Similarly to the low mass ratio effect during the VIV of the circular cylinder, the two branches exchange their roles and no distinction can be made between the WM and the SM for the three geometries. Therefore, we consider the coupled modes WSMI and WSMII to classify the stability characteristics for these geometries. When the intersection of the growth rate curves corresponding to WSMI and WSMII occurs in figure 20(b) (top), the stability roles of WSMI and WSMII switch at a specific value of $F_{s}$ for the three geometries. As shown in figure 20(b) (bottom), the two curves of $\text{Im}(\unicode[STIX]{x1D706})$ for the three geometries no longer cross, in comparison to the circular cylinder counterpart in identical conditions. Due to stronger coupling, there is a characteristic anticrossing with a frequency splitting between WSMI and WSMII for the three geometries. In addition, the forward triangle branch departs further away from the line $f=2\unicode[STIX]{x03C0}F_{s}$ compared with the diamond and elliptical cylinders.

In contrast to the square cylinder ( $0^{\circ }$ degree flow incidence) at $(Re,m^{\ast })=(60,10)$ , the diamond configuration has a flutter-dominated VIV lock-in. This difference in the lock-in behaviour can be attributed to the boundary layer movement over the front edges of the diamond cylinder, whereby the square cylinder has flow separations at the upstream corners and inhibits the co-existence of flutter and resonance regimes.

Figure 20.

Effect of geometry on the eigenspectrum of the ERA-based ROM at $(Re,m^{\ast })=(60,10)$ : (a) root loci as a function of the reduced natural frequency $F_{s}$ , where the unstable right half-plane $(\text{Re}(\unicode[STIX]{x1D706})>0)$ is shaded in grey colour and the hollow arrow indicates increasing $F_{s}$ ; (b) real and imaginary parts of the root loci. The WSMI data are denoted by filled symbols with the same shapes as those for WSMII in (a,b). The onset $U_{r}$ is computed on $F_{s}=0.184$ (forward triangle – - – - –, blue), $F_{s}=0.216$ (diamond – - – - –, red) and $F_{s}=0.239$ (ellipse – - – - –, green) in the real parts of the root loci in (b).

Figure 21.

The VIV results for the forward triangle configuration using the FOM at $(Re,m^{\ast })=(60,10)$ : variation of (a) the normalized vortex shedding frequency $f$ and (b) the r.m.s. value of the lift coefficient ( $C_{l}$ ) and the maximum amplitude ( $Y_{max}$ ). The lock-in is shaded in grey colour.

For the forward triangle configuration, it is notable that WSMI and WSMII remain unstable for $F_{s}\leqslant 0.184$ or $U_{r}\geqslant 5.43$ , as predicted by the ERA-based ROM, which indicates that flutter-dominated VIV persists. Therefore, the forward triangle is of particular interest for the present study. These linear stability results have been confirmed by the FOM simulations, as shown in figure 21. The amplitude grows continually for $F_{s}\leqslant 0.184$ or $U_{r}\geqslant 5.43$ and the lift coefficient reaches its maximum value at the lock-in onset $U_{r}$ , which is similar to the circular and square cylinders. The vortex shedding frequency starts to synchronize with the natural frequency of the structure and there exists 1:1 frequency synchronization, whereby the body is synchronized with the vortex shedding frequency. The transverse amplitude grows continually as $F_{s}$ decreases ( $U_{r}$ increases), which is referred to as galloping-dominated flow-induced vibration. This galloping regime is characterized by a low-frequency and high-amplitude vibration, whereas a circular cross-section is not susceptible to galloping. For $F_{s}<0.085$ ( $U_{r}>11.76$ ), the amplitude experiences a small but distinct increase in amplitude, and 1:3 synchronization gradually appears, as shown in figure 22. In this regime, there is net energy transfer from the base flow with the frequency component at three times the body oscillation frequency. During this energy transfer, the fluid force performs work on the body, which stores the energy in the form of kinetic energy as well as the potential energy in the spring. To further elucidate the high harmonic response for the forward triangle, figure 22 depicts motion and lift force traces with their corresponding spectra. In the figure, there is a clear third-harmonic frequency in the lift force where the body oscillates with a dominant frequency. Figure 23 shows the instantaneous vorticity contours at different values of the reduced natural frequency $F_{s}$ . The vortex shedding mode is 2S for $F_{s}=0.17$ , and remains 2S for $F_{s}=0.15$ with somewhat increased spacing between the vortices shed alternately from each side of the cylinder. By further decreasing $F_{s}$ to 0.1, the strong SM and WM interactions result in a larger vibration amplitude and the flow diverges to a wide vortex street.

Figure 22.

Full-order VIV results for a forward triangle with 1:3 synchronization: temporal variation of (a) the transverse amplitude and (c) the lift coefficient; normalized power spectrum $P$ versus $f^{\ast }$ of (b) the transverse amplitude and (d) the lift coefficient at $F_{s}=0.05$ , where $f^{\ast }=f/F_{s}$ is the frequency of lift and transverse displacement normalized by the reduced natural frequency $F_{s}$ . A third-harmonic frequency is evident in $C_{l}$ .

Figure 23.

Full-order results for the forward triangle cylinder at $(Re,m^{\ast })=(60,10)$ . Instantaneous vorticity contours at $F_{s}=$  (a) 0.05, (b) 0.1, (c) 0.15 and (d) 0.17. The contour levels are from $-$ 0.5 to 0.5 in increments of 0.077 and the flow is from left to right.

The geometry of the bluff body alters the flow structures significantly in the vicinity of the base flow, resulting in different root loci and subsequently changing the stability properties. For example, as shown in figure 15, the sharp corner of the square stabilizes the flow and the resonance mode dominates the entire lock-in for $(Re,m^{\ast })=(60,10)$ . This can be further elucidated by looking into the magnitude of the leading modal velocity $|\widehat{U}|$ for the different geometrical configurations. As illustrated in figure 24, the stability regions of the ellipse, diamond and triangle geometries shift upstream in comparison with the circular geometry, indicating that the coupling effect is enhanced. The stationary configurations of the ellipse, diamond and forward triangle have quite similar stability regions. The forward triangle is easier to perturb from the flow unsteadiness, or by a Hopf bifurcation for alternate vortex shedding. However, the sharp corners generally stabilize the fluid–structure system, as the separation point is constrained and there is less freedom for the interaction of flow and structural modes. Due to this attribute in the forward triangle configuration, there is a delay in lock-in onset $U_{r}$ with identical operating parameters $(Re,m^{\ast })$ compared with the elliptical cylinder.

Figure 24.

Stability regions shown by the spatial distribution of pointwise modal velocity amplitude $|\widehat{U}|$ at $Re=60$ : contours of modal velocity for the (a) circle, (b) ellipse, (c) diamond and (d) forward triangle configurations. The flow is from left to right.

Figure 25.

Stability phase diagram of VIV lock-in for transversely vibrating two-dimensional bluff bodies with smooth curves and sharp corners for $30\leqslant Re\leqslant 100$ , $m^{\ast }=10$ and $0.05\leqslant F_{s}\leqslant 0.25$ . Here, the solid curve (——) represents the critical Reynolds number ( $Re_{cr}$ ) of the fixed bluff body, and flutter- and resonance-induced regimes are demarcated, where ▫ represents the co-existence of flutter and resonance regimes, ○ denotes the resonance regime and ▵ represents the flutter regime. For $Re>Re_{cr}$ , the flutter regime comprises both unstable eigenvalues $(\text{Re}(\unicode[STIX]{x1D706})>0)$ of the WM and SM, whereas the resonance regime has only unstable WM.

Figure 25 summarizes the stability regimes of transversely vibrating bluff bodies for the Reynolds number range $30\leqslant Re\leqslant 100$ . In this figure, the solid curve (——) depicts the trend for the critical Reynolds number. The following observations can be made from the stability phase diagram. The circle, ellipse and diamond geometries exhibit the flutter and mixed resonance–flutter modes. In contrast to its square counterpart, the diamond geometry has a movement of asymmetric boundary layers on the front lateral edges which allows the co-existence of flutter and resonance modes. For the elliptic cylinder, the mixed flutter–resonance regime occurs at a lower Reynolds number in comparison with the circular cylinder. Notably, the forward triangle configuration only shows the flutter-induced lock-in regime for this range of Reynolds number and the edges are on the leeward side with separated flow. Finally, the square-section body shows a predominant resonance regime for approximately $30\leqslant Re\leqslant 80$ , which turns into the flutter state for $Re>80$ .

The present ERA-based ROM study has been concerned with two-dimensional bluff-body configurations for which only two directions in space are resolved. All of the notions of the ROM, such as base flows and eigenvalue realization, can be easily extended to full three-dimensional settings. Thus, the present method does not pose any theoretical limitation except that there may be numerical ones with respect to memory requirements and CPU time to solve the generalized eigenvalue problem.